Mathematical Methods Handbook

Orthogonal polynomials

Legendre, Hermite, Laguerre and Chebyshev polynomials and spherical harmonics.

In this chapter 8

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  1. Classical orthogonal polynomials

Classical orthogonal polynomials

8

Theorem

General properties of orthogonal polynomials

Adapted

For any positive weight with finite moments the monomials can be orthogonalised into a unique family of polynomials of degrees 0, 1, 2, ...; the family obeys a three-term recurrence, each member has exactly its own degree many simple zeros inside the interval, and the classical families all come from one Rodrigues formula.

Prerequisites: Orthonormal bases and Gram-Schmidt, Inner products and norms, Sturm-Liouville eigenvalue problems

Definition(Weight, weighted inner product, orthogonal family)AdaptedAdapted from Hassani (2013), §7.2, p. 222, and §7.2.1, p. 222

Let (a,b)⊆R(a,b)\subseteq\mathbb{R} be an interval, finite or infinite. A weight on (a,b)(a,b) is a measurable function ww with

  • w(x)>0w(x)>0 for every x∈(a,b)x\in(a,b), and
  • finite moments μk=∫abxkw(x) dx\displaystyle \mu_{k}=\int_{a}^{b}x^{k}w(x)\,\mathrm{d}x for every integer k≥0k\ge 0.

Hassani's Lw2(a,b)L^{2}_{w}(a,b) is the space of measurable ff with ∫ab∣f∣2w<∞\int_{a}^{b}\lvert f\rvert^{2}w<\infty, carrying

(f,g)w=∫abf(x) g(x)‾ w(x) dx,∥f∥w=(f,f)w1/2,(f,g)_{w} = \int_{a}^{b} f(x)\,\overline{g(x)}\,w(x)\,\mathrm{d}x , \qquad \lVert f\rVert_{w} = (f,f)_{w}^{1/2},

and it is complete. His standing requirement on ww is that it be strictly positive and integrable on (a,b)(a,b); the moment condition above is the extra hypothesis that makes every monomial a member of Lw2L^{2}_{w}, since ∥xk∥w2=μ2k\lVert x^{k}\rVert_{w}^{2}=\mu_{2k}. Both are needed here and both are stated, rather than left to be read off from the examples.

A sequence (pn)n≥0(p_{n})_{n\ge0} of real polynomials is an orthogonal family for ww if deg⁡pn=n\deg p_{n}=n for every nn and (pn,pm)w=0(p_{n},p_{m})_{w}=0 whenever n≠mn\neq m. We write

hn=∥pn∥w2h_{n} = \lVert p_{n}\rVert_{w}^{2}

for the squared norms, and kn(n)k_{n}^{(n)}, kn(n−1)k_{n}^{(n-1)} for the coefficients of xnx^{n} and of xn−1x^{n-1} in pnp_{n}, following Hassani's notation. The family is monic if kn(n)=1k_{n}^{(n)}=1 for all nn. A standardisation is any rule fixing the one free scalar in each pnp_{n}: Legendre polynomials are standardised by pn(1)=1p_{n}(1)=1, Chebyshev polynomials of the first kind by a leading coefficient 2n−12^{n-1}, and so on.

Theorem(Existence, uniqueness, and orthogonality to lower degrees)AdaptedAdapted from Hassani (2013), §7.2.1, p. 222, and eq. (7.7), p. 223; Kreyszig (1978), §3.7-1, pp. 176–179

Let ww be a weight on (a,b)(a,b) in the sense above. Then:

  1. Existence. Applying the Gram-Schmidt process of Orthonormal bases and Gram-Schmidt to 1,x,x2,…1,x,x^{2},\dots in Lw2(a,b)L^{2}_{w}(a,b) produces an orthogonal family (pn)(p_{n}) for ww, with deg⁡pn=n\deg p_{n}=n exactly.
  2. Uniqueness up to scale. If (pn)(p_{n}) and (p~n)(\tilde p_{n}) are both orthogonal families for ww, then p~n=cnpn\tilde p_{n}=c_{n}p_{n} with cn≠0c_{n}\neq0 for every nn. In particular the monic family is unique, and any standardisation determines the family completely.
  3. Orthogonality to all lower degrees. For every nn, (pn,q)w=0for every polynomial q with deg⁡q<n.(p_{n},q)_{w}=0 \qquad\text{for every polynomial } q \text{ with } \deg q<n . Equivalently, pnp_{n} is orthogonal to 1,x,…,xn−11,x,\dots,x^{n-1}.

Hassani records part 3 as the "obvious property" that any polynomial of degree below nn is a combination of p0,…,pn−1p_{0},\dots,p_{n-1}.

Theorem(Three-term recurrence)AdaptedAdapted from Hassani (2013), §7.2.1, eqs. (7.9)–(7.13), pp. 223–224

Let (pn)(p_{n}) be an orthogonal family for ww. Then for every n≥1n\ge1,

pn+1(x)=(αnx+βn) pn(x)+γn pn−1(x),(∗)p_{n+1}(x) = (\alpha_{n}x+\beta_{n})\,p_{n}(x) + \gamma_{n}\,p_{n-1}(x), \tag{$\ast$}

with constants

αn=kn+1(n+1)kn(n),βn=αn(kn+1(n)kn+1(n+1)−kn(n−1)kn(n)),γn=−hnαnhn−1αn−1.\alpha_{n} = \frac{k_{n+1}^{(n+1)}}{k_{n}^{(n)}}, \qquad \beta_{n} = \alpha_{n}\left(\frac{k_{n+1}^{(n)}}{k_{n+1}^{(n+1)}}-\frac{k_{n}^{(n-1)}}{k_{n}^{(n)}}\right), \qquad \gamma_{n} = -\frac{h_{n}\alpha_{n}}{h_{n-1}\alpha_{n-1}} .

No term in pn−2p_{n-2} or below occurs. In particular γn\gamma_{n} has the sign opposite to αn/αn−1\alpha_{n}/\alpha_{n-1}, so for a monic family (αn≡1\alpha_{n}\equiv1) every γn\gamma_{n} is strictly negative.

For the monic family the recurrence takes the form used in computation,

pn+1(x)=(x−Bn) pn(x)−Cn pn−1(x),Bn=(xpn,pn)whn,Cn=hnhn−1>0,(∗∗)p_{n+1}(x) = (x-B_{n})\,p_{n}(x) - C_{n}\,p_{n-1}(x), \qquad B_{n} = \frac{(xp_{n},p_{n})_{w}}{h_{n}}, \qquad C_{n} = \frac{h_{n}}{h_{n-1}}>0 , \tag{$\ast\ast$}

valid for n≥1n\ge1, together with p0=1p_{0}=1 and p1(x)=x−B0p_{1}(x)=x-B_{0}, B0=μ1/μ0B_{0}=\mu_{1}/\mu_{0}.

Theorem(The zeros are real, simple and interior)Original

Let (pn)(p_{n}) be an orthogonal family for a weight ww on (a,b)(a,b), and let n≥1n\ge1. Then pnp_{n} has exactly nn zeros, they are all real and simple, and they all lie in the open interval (a,b)(a,b).

Theorem(Rodrigues formulas and the classical families)AdaptedAdapted from Hassani (2013), §8.1, Thm 8.1.1 and eq. (8.1), p. 241, Prop. 8.1.4, p. 243, eq. (8.2), p. 243, and §8.2, Prop. 8.2.1, p. 245; Arfken, Weber & Harris (2013), §12.1, eqs. (12.1)–(12.9), pp. 551–552, and Table 12.1, p. 559

The classical families are not built by Gram-Schmidt but written down in closed form. Two equivalent routes are standard.

Hassani's route (Tricomi's). Suppose ss is a polynomial of degree at most 22 with only real roots, ww is strictly positive and integrable on (a,b)(a,b) with

w(a)s(a)=0=w(b)s(b),w(a)s(a)=0=w(b)s(b),

and the function F1F_{1} defined below is a polynomial of degree 11. Set

Fn(x)=1Knw(x)dndxn[w(x)s(x)n]F_{n}(x) = \frac{1}{K_{n}w(x)}\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\bigl[w(x)s(x)^{n}\bigr]

with normalising constants KnK_{n}: this is the generalised Rodrigues formula. Then each FnF_{n} is a polynomial of degree exactly nn and ∫abFnp w=0\int_{a}^{b}F_{n}p\,w=0 for every polynomial pp of degree <n<n, so (Fn)(F_{n}) is an orthogonal family for ww. Moreover each FnF_{n} solves the Sturm-Liouville equation

ddx(ws dFndx)=wλnFn,λn=K1k1(1)n+σ2n(n−1),\frac{\mathrm{d}}{\mathrm{d}x}\Bigl(ws\,\frac{\mathrm{d}F_{n}}{\mathrm{d}x}\Bigr)=w\lambda_{n}F_{n}, \qquad \lambda_{n}=K_{1}k_{1}^{(1)}n+\sigma_{2}n(n-1),

where σ2\sigma_{2} is the coefficient of x2x^{2} in ss.

Arfken's route. Start instead from a second-order equation with polynomial coefficients,

p(x)y′′+q(x)y′+λy=0,p=αx2+βx+γ,q=μx+ν,p(x)y''+q(x)y'+\lambda y=0, \qquad p=\alpha x^{2}+\beta x+\gamma,\quad q=\mu x+\nu ,

multiply by the factor w=p−1exp⁡ ⁣∫(q/p)w=p^{-1}\exp\!\int (q/p) that makes it self-adjoint (so that (wp)′=wq(wp)'=wq), and then

yn(x)=1w(x)(ddx)n[w(x)p(x)n]y_{n}(x)=\frac{1}{w(x)}\Bigl(\frac{\mathrm{d}}{\mathrm{d}x}\Bigr)^{n}\bigl[w(x)p(x)^{n}\bigr]

is a polynomial of degree nn solving the equation with

λn=−n(n−1)α−nμ.\lambda_{n}=-n(n-1)\alpha-n\mu .

The two routes agree with s=ps=p: Arfken's self-adjointness condition (wp)′=wq(wp)'=wq is Hassani's requirement that (1/w) d(ws)/dx(1/w)\,\mathrm{d}(ws)/\mathrm{d}x be a polynomial of degree 11.

Classification. Hassani's Proposition 8.2.1 shows that the degree of ss decides everything. Up to an affine change of variable there are only three cases:

deg⁡s\deg s s(x)s(x) w(x)w(x) (a,b)(a,b) family
00 11 e−x2\mathrm{e}^{-x^{2}} (−∞,∞)(-\infty,\infty) Hermite HnH_{n}
11 xx xνe−xx^{\nu}\mathrm{e}^{-x}, ν>−1\nu>-1 (0,∞)(0,\infty) Laguerre LnνL^{\nu}_{n}
22 1−x21-x^{2} (1+x)μ(1−x)ν(1+x)^{\mu}(1-x)^{\nu}, μ,ν>−1\mu,\nu>-1 (−1,1)(-1,1) Jacobi Pnμ,νP^{\mu,\nu}_{n}

Legendre (μ=ν=0\mu=\nu=0), Gegenbauer (μ=ν=λ−12\mu=\nu=\lambda-\tfrac12) and the two kinds of Chebyshev (μ=ν=∓12\mu=\nu=\mp\tfrac12) are the special Jacobi cases of Hassani's Table 8.1. Arfken's Table 12.1 lists the equation, the Rodrigues formula and the generating function for Legendre, Hermite, Laguerre and both Chebyshev families; the individual families are the subject of Legendre polynomials, Hermite polynomials, Laguerre polynomials and Chebyshev polynomials.

Note the direction of the logic: Rodrigues formulas give some orthogonal families, in closed form, together with a differential equation. Gram-Schmidt gives the family for every admissible weight, but no closed form and no differential equation. For a weight outside the table — for instance w(x)=xw(x)=x on (0,1)(0,1) in the example below, which is Jacobi only after the affine map to (−1,1)(-1,1), or w(x)=e−x4w(x)=\mathrm{e}^{-x^{4}} on R\mathbb{R}, which is in no case at all — the family exists, obeys a three-term recurrence and has interior simple zeros, but satisfies no second-order equation of the above form.

Idea of the proofOriginal

The first three theorems are the single fact of part 3 of the existence statement — pnp_{n} is orthogonal to everything of lower degree — used three times.

Existence and uniqueness. Gram-Schmidt cannot fail, because w>0w>0 makes ∥q∥w>0\lVert q\rVert_{w}>0 for every nonzero polynomial qq, so no vector in the process is ever killed; and taking n+1n+1 monomials in and getting n+1n+1 polynomials out forces the degrees to be 0,1,…,n0,1,\dots,n. Uniqueness holds because pnp_{n} is determined up to scale as the one direction in the (n+1)(n+1)-dimensional space of polynomials of degree ≤n\le n perpendicular to the nn-dimensional subspace of degree <n<n.

Recurrence. Subtract from pn+1p_{n+1} the multiple of xpnxp_{n} that cancels its leading term. What is left has degree ≤n\le n, so it expands in p0,…,pnp_{0},\dots,p_{n}. Test that expansion against pmp_{m}: the pn+1p_{n+1} term dies for m≤nm\le n by orthogonality, and the xpnxp_{n} term dies as soon as xpmxp_{m} has degree below n+1n+1, that is for m≤n−2m\le n-2. So all coefficients below pn−1p_{n-1} vanish and only three terms survive.

Zeros. If pnp_{n} changed sign at fewer than nn interior points, we could manufacture a low-degree polynomial qq whose sign changes are exactly those of pnp_{n}. Then pnqp_{n}q never changes sign, so ∫pnq w\int p_{n}q\,w cannot vanish — yet it must, because deg⁡q<n\deg q<n.

Rodrigues. Different in kind, and the only place where ww has to be differentiable. The whole statement rests on one bookkeeping lemma: differentiating wsnw s^{n} once lowers the power of ss by one and raises the degree of the polynomial factor by one, because the hypothesis on F1F_{1} says that (ws)′(ws)' is ww times a polynomial of degree ≤1\le1. Iterating gives dm(wsn)/dxm=wsn−m×(polynomial)\mathrm{d}^{m}(ws^{n})/\mathrm{d}x^{m}=ws^{n-m}\times(\text{polynomial}), which is a polynomial times ww when m=nm=n — so FnF_{n} is a polynomial — and which vanishes at both ends when m<nm<n, because a factor wsws survives. Those vanishing boundary values are exactly what is needed to integrate by parts k+1k+1 times against a polynomial of degree k<nk<n and arrive at zero.

Proof(Existence, uniqueness, orthogonality to lower degrees)AdaptedAdapted from Hassani (2013), §7.2.1, pp. 222–223; Kreyszig (1978), §3.7-1, pp. 176–179

The monomials lie in Lw2L^{2}_{w} and are independent. ∥xk∥w2=μ2k<∞\lVert x^{k}\rVert_{w}^{2}=\mu_{2k}<\infty by the moment hypothesis, so each monomial is in the space. A nontrivial linear combination of monomials is a nonzero polynomial qq, and

∥q∥w2=∫ab∣q∣2w>0\lVert q\rVert_{w}^{2}=\int_{a}^{b}\lvert q\rvert^{2}w>0

because ∣q∣2w\lvert q\rvert^{2}w is continuous, non-negative, strictly positive off the finitely many zeros of qq — this is where w>0w>0 enters. So q≠0q\neq 0 in Lw2L^{2}_{w} and the monomials are linearly independent.

Gram-Schmidt. Run the process of Orthonormal bases and Gram-Schmidt on 1,x,x2,…1,x,x^{2},\dots. At step nn it subtracts from xnx^{n} its projection onto the span of the previous outputs and divides by a norm that we have just shown is nonzero. Each output is a linear combination of 1,…,xn1,\dots,x^{n}, hence a polynomial of degree ≤n\le n; and the coefficient of xnx^{n} is the (nonzero) reciprocal of that norm times 11, because the subtracted projection involves only 1,…,xn−11,\dots,x^{n-1}. So deg⁡pn=n\deg p_{n}=n exactly. This is the construction Kreyszig carries out on [−1,1][-1,1] with w≡1w\equiv1 to produce the Legendre polynomials, and Hassani's remark that orthogonalising the monomials must give polynomials.

Orthogonality to lower degrees. By construction span⁡{p0,…,pn−1}=span⁡{1,…,xn−1}=:Pn−1\operatorname{span}\{p_{0},\dots,p_{n-1}\}=\operatorname{span}\{1,\dots,x^{n-1}\}=:\mathcal{P}_{n-1}, the polynomials of degree ≤n−1\le n-1. Since pn⊥pmp_{n}\perp p_{m} for m<nm<n, pn⊥Pn−1p_{n}\perp\mathcal{P}_{n-1}. This is part 3.

Uniqueness. Let p~n\tilde p_{n} be any polynomial of degree nn orthogonal to Pn−1\mathcal{P}_{n-1}. Write p~n=∑j=0ncjpj\tilde p_{n}=\sum_{j=0}^{n}c_{j}p_{j}, possible because deg⁡p~n≤n\deg\tilde p_{n}\le n. For m<nm<n, orthogonality to pm∈Pn−1p_{m}\in\mathcal{P}_{n-1} gives 0=(p~n,pm)w=cmhm0=(\tilde p_{n},p_{m})_{w}=c_{m}h_{m}, and hm>0h_{m}>0, so cm=0c_{m}=0. Hence p~n=cnpn\tilde p_{n}=c_{n}p_{n}, with cn≠0c_{n}\neq0 since deg⁡p~n=n\deg\tilde p_{n}=n. Any second orthogonal family satisfies the same hypothesis at each nn, which proves part 2. □\square

Proof(The three-term recurrence)AdaptedAdapted from Hassani (2013), §7.2.1, eqs. (7.9)–(7.11), pp. 223–224

Step 1: kill the leading term. The polynomial

r(x)=pn+1(x)−αn x pn(x),αn=kn+1(n+1)kn(n),r(x) = p_{n+1}(x) - \alpha_{n}\,x\,p_{n}(x), \qquad \alpha_{n}=\frac{k_{n+1}^{(n+1)}}{k_{n}^{(n)}},

has no xn+1x^{n+1}: the coefficient of xn+1x^{n+1} in xpnxp_{n} is kn(n)k_{n}^{(n)}, and αn\alpha_{n} was chosen to match. So deg⁡r≤n\deg r\le n and r=∑j=0najpjr=\sum_{j=0}^{n}a_{j}p_{j} for some scalars aja_{j}.

Step 2: test against pmp_{m} for m≤n−2m\le n-2. Take ( ⋅ ,pm)w(\,\cdot\,,p_{m})_{w} of r=pn+1−αnxpnr=p_{n+1}-\alpha_{n}xp_{n}:

(r,pm)w=(pn+1,pm)w−αn (xpn,pm)w=0−αn (pn,xpm)w.(r,p_{m})_{w} = (p_{n+1},p_{m})_{w} - \alpha_{n}\,(xp_{n},p_{m})_{w} = 0 - \alpha_{n}\,(p_{n},xp_{m})_{w} .

The first term vanishes by orthogonality since m≤n−2<n+1m\le n-2<n+1. In the second we moved the real factor xx across the inner product, which is legitimate because xx is real and the inner product has x pm‾=x pm‾\overline{x\,p_{m}}=x\,\overline{p_{m}}. Now deg⁡(xpm)=m+1≤n−1<n\deg(xp_{m})=m+1\le n-1<n, so (pn,xpm)w=0(p_{n},xp_{m})_{w}=0 by part 3 of the previous statement. Hence (r,pm)w=0(r,p_{m})_{w}=0, while also (r,pm)w=amhm(r,p_{m})_{w}=a_{m}h_{m} from the expansion. Since hm>0h_{m}>0,

am=0for m=0,1,…,n−2,a_{m}=0 \qquad\text{for } m=0,1,\dots,n-2 ,

which is exactly where the hypothesis "orthogonal to all lower degrees" does the work. Only ana_{n} and an−1a_{n-1} survive, giving (∗)(\ast) with βn=an\beta_{n}=a_{n} and γn=an−1\gamma_{n}=a_{n-1}.

Step 3: identify βn\beta_{n}. Compare coefficients of xnx^{n} in pn+1=αnxpn+βnpn+γnpn−1p_{n+1}=\alpha_{n}xp_{n}+\beta_{n}p_{n}+\gamma_{n}p_{n-1}. The last term contributes nothing (deg⁡pn−1=n−1\deg p_{n-1}=n-1), so

kn+1(n)=αnkn(n−1)+βnkn(n),k_{n+1}^{(n)} = \alpha_{n}k_{n}^{(n-1)}+\beta_{n}k_{n}^{(n)} ,

and solving, with αnkn(n)=kn+1(n+1)\alpha_{n}k_{n}^{(n)}=k_{n+1}^{(n+1)},

βn=kn+1(n)kn(n)−αnkn(n−1)kn(n)=αn(kn+1(n)kn+1(n+1)−kn(n−1)kn(n)),\beta_{n}=\frac{k_{n+1}^{(n)}}{k_{n}^{(n)}}-\alpha_{n}\frac{k_{n}^{(n-1)}}{k_{n}^{(n)}} =\alpha_{n}\left(\frac{k_{n+1}^{(n)}}{k_{n+1}^{(n+1)}}-\frac{k_{n}^{(n-1)}}{k_{n}^{(n)}}\right),

Hassani's formula.

Step 4: identify γn\gamma_{n}. Take ( ⋅ ,pn−1)w(\,\cdot\,,p_{n-1})_{w} of (∗)(\ast). The left side gives 00; on the right, (pn,pn−1)w=0(p_{n},p_{n-1})_{w}=0, so

0=αn(xpn,pn−1)w+γnhn−1=αn(pn,xpn−1)w+γnhn−1.0 = \alpha_{n}(xp_{n},p_{n-1})_{w}+\gamma_{n}h_{n-1} = \alpha_{n}(p_{n},xp_{n-1})_{w}+\gamma_{n}h_{n-1}.

To evaluate (pn,xpn−1)w(p_{n},xp_{n-1})_{w}, apply (∗)(\ast) at index n−1n-1: xpn−1=αn−1−1(pn−βn−1pn−1−γn−1pn−2)xp_{n-1}=\alpha_{n-1}^{-1}\bigl(p_{n}-\beta_{n-1}p_{n-1}-\gamma_{n-1}p_{n-2}\bigr), and only the pnp_{n} term survives against pnp_{n}:

(pn,xpn−1)w=hnαn−1.(p_{n},xp_{n-1})_{w}=\frac{h_{n}}{\alpha_{n-1}} .

Therefore γn=−αnhn/(αn−1hn−1)\gamma_{n}=-\alpha_{n}h_{n}/(\alpha_{n-1}h_{n-1}), as claimed. Since hn,hn−1>0h_{n},h_{n-1}>0, the sign of γn\gamma_{n} is opposite to that of αn/αn−1\alpha_{n}/\alpha_{n-1}.

Step 5: the monic form. Put αn=1\alpha_{n}=1 for all nn, write Bn=−βnB_{n}=-\beta_{n} and Cn=−γn=hn/hn−1>0C_{n}=-\gamma_{n}=h_{n}/h_{n-1}>0; then (∗)(\ast) becomes (∗∗)(\ast\ast). For BnB_{n} the coefficient formula is awkward to use directly, so take ( ⋅ ,pn)w(\,\cdot\,,p_{n})_{w} of (∗∗)(\ast\ast) instead:

0=(xpn,pn)w−Bnhn−Cn (pn−1,pn)w=(xpn,pn)w−Bnhn,0=(xp_{n},p_{n})_{w}-B_{n}h_{n}-C_{n}\,(p_{n-1},p_{n})_{w}=(xp_{n},p_{n})_{w}-B_{n}h_{n},

giving Bn=(xpn,pn)w/hnB_{n}=(xp_{n},p_{n})_{w}/h_{n}. For n=0n=0 the same computation with p0=1p_{0}=1 gives B0=(x,1)w/(1,1)w=μ1/μ0B_{0}=(x,1)_{w}/(1,1)_{w}=\mu_{1}/\mu_{0}. □\square

Proof(The zeros)Original

pnp_{n} is a real polynomial of degree nn, so it has nn zeros in C\mathbb{C} counted with multiplicity. We show nn of them are distinct points of (a,b)(a,b); that exhausts the count and leaves none complex and none repeated.

Step 1: count the interior sign changes. Let x1<⋯<xrx_{1}<\dots<x_{r} be the points of (a,b)(a,b) at which pnp_{n} changes sign, that is the real zeros of pnp_{n} in (a,b)(a,b) of odd multiplicity. Suppose, for contradiction, that r<nr<n. (If pnp_{n} has no interior sign change, take r=0r=0 and q≡1q\equiv1 below.)

Step 2: build a test polynomial. Put

q(x)=∏j=1r(x−xj),q(x)=\prod_{j=1}^{r}(x-x_{j}),

a real polynomial of degree r<nr<n.

Step 3: the product has one sign. The function pnqp_{n}q is a real polynomial that changes sign at no point of (a,b)(a,b): at each xjx_{j} both factors change sign, so the product does not, and at every other point of (a,b)(a,b) neither factor does. Hence pnq≥0p_{n}q\ge0 throughout (a,b)(a,b) or pnq≤0p_{n}q\le0 throughout (a,b)(a,b). It is not the zero polynomial, since pn≠0p_{n}\neq0 and q≠0q\neq0.

Step 4: the integral cannot vanish. Replacing qq by −q-q if necessary, assume pnq≥0p_{n}q\ge0 on (a,b)(a,b). The product vanishes only at its finitely many zeros, and w>0w>0 everywhere on (a,b)(a,b), so pnq wp_{n}q\,w is non-negative and strictly positive except at finitely many points. Therefore

∫abpn(x)q(x)w(x) dx>0.\int_{a}^{b}p_{n}(x)q(x)w(x)\,\mathrm{d}x>0 .

(The integral is finite: pnqp_{n}q is a polynomial, and all moments of ww are finite.) This is the step that needs w>0w>0 on all of (a,b)(a,b).

Step 5: but orthogonality forbids it. qq is real and deg⁡q=r<n\deg q=r<n, so by part 3 of the existence statement ∫abpnq w=(pn,q)w=0\int_{a}^{b}p_{n}q\,w=(p_{n},q)_{w}=0. This contradicts Step 4.

Hence r≥nr\ge n. A degree-nn polynomial has at most nn zeros, so r=nr=n: there are exactly nn distinct points of (a,b)(a,b) where pnp_{n} changes sign. Each is a zero, and together they account for all nn zeros of pnp_{n}; so each has multiplicity one and there are no others, real or complex. □\square

Proof(The Rodrigues construction)AdaptedAdapted from Hassani (2013), §8.1, Lemmas 8.1.2 and 8.1.3 and the proof of Thm 8.1.1, pp. 242–243

We prove Hassani's route; his Lemma 8.1.2 is left to a problem in the book, so we prove it too. Write p≤kp_{\le k} for an unspecified polynomial of degree at most kk, as Hassani does, and let Fn=(Knw)−1dn(wsn)/dxnF_{n}=(K_{n}w)^{-1}\mathrm{d}^{n}(ws^{n})/\mathrm{d}x^{n}. The hypotheses are: deg⁡s≤2\deg s\le2 with real roots only; w>0w>0 and integrable on (a,b)(a,b) with w(a)s(a)=0=w(b)s(b)w(a)s(a)=0=w(b)s(b); and F1F_{1} a polynomial of degree 11. The last one says precisely that

d(ws)dx=w t,t=K1F1 a polynomial with deg⁡t≤1.(H)\frac{\mathrm{d}(ws)}{\mathrm{d}x}=w\,t, \qquad t=K_{1}F_{1}\ \text{a polynomial with}\ \deg t\le1 . \tag{H}

Lemma A (Hassani's 8.1.2). For 1≤m≤n1\le m\le n,

dmdxm[w sn p≤k]=w s n−m p≤k+m.\frac{\mathrm{d}^{m}}{\mathrm{d}x^{m}}\bigl[w\,s^{n}\,p_{\le k}\bigr]=w\,s^{\,n-m}\,p_{\le k+m} .

Proof, by induction on mm. For m=1m=1, expand and use w′=w(t−s′)/sw'=w(t-s')/s, which is (H) rearranged (valid where s≠0s\neq0, and the conclusion is an identity between continuous functions, so it extends to the whole interval):

[wsnp≤k]′=w′ snp≤k+n w sn−1s′ p≤k+w snp≤k′=w s n−1[(t+(n−1)s′)p≤k+s p≤k′].\bigl[ws^{n}p_{\le k}\bigr]' = w'\,s^{n}p_{\le k}+n\,w\,s^{n-1}s'\,p_{\le k}+w\,s^{n}p_{\le k}' = w\,s^{\,n-1}\Bigl[\bigl(t+(n-1)s'\bigr)p_{\le k}+s\,p_{\le k}'\Bigr].

In the bracket, tt and s′s' have degree ≤1\le1, so the first product has degree ≤k+1\le k+1; and s p≤k′s\,p_{\le k}' has degree ≤2+(k−1)=k+1\le 2+(k-1)=k+1. The bracket is therefore p≤k+1p_{\le k+1}, which is the claim for m=1m=1. If the claim holds for m<nm<n, apply the case m=1m=1 once more with n−mn-m in place of nn and k+mk+m in place of kk; the power of ss is still at least 11, so the step is legitimate, and we get ws n−m−1p≤k+m+1ws^{\,n-m-1}p_{\le k+m+1}. □\square

Lemma B (Hassani's 8.1.3). For 0≤m<n0\le m<n, dm(wsn)/dxm\mathrm{d}^{m}(ws^{n})/\mathrm{d}x^{m} vanishes at x=ax=a and at x=bx=b. Proof. Take k=0k=0, p≤0=1p_{\le0}=1 in Lemma A: the derivative equals w s n−m p≤mw\,s^{\,n-m}\,p_{\le m}. Since m<nm<n we have n−m≥1n-m\ge1, so this is (ws) s n−m−1p≤m(ws)\,s^{\,n-m-1}p_{\le m}, and wsws vanishes at both endpoints by hypothesis. □\square

FnF_{n} is a polynomial of degree at most nn. Lemma A with m=nm=n, k=0k=0 gives dn(wsn)/dxn=w s0 p≤n=w p≤n\mathrm{d}^{n}(ws^{n})/\mathrm{d}x^{n}=w\,s^{0}\,p_{\le n}=w\,p_{\le n}, so Fn=p≤n/KnF_{n}=p_{\le n}/K_{n} is a polynomial of degree at most nn.

Orthogonality to lower degrees. Let qq be a polynomial with deg⁡q=k<n\deg q=k<n. Then

Kn∫abq Fn w dx=∫abq(x) dn(wsn)dxn dx.K_{n}\int_{a}^{b}q\,F_{n}\,w\,\mathrm{d}x=\int_{a}^{b}q(x)\,\frac{\mathrm{d}^{n}(ws^{n})}{\mathrm{d}x^{n}}\,\mathrm{d}x .

Integrate by parts, moving one derivative at a time onto qq. After jj steps the boundary term is

(−1)j−1[q(j−1) d n−j(wsn)dx n−j]ab,(-1)^{j-1}\Bigl[q^{(j-1)}\,\frac{\mathrm{d}^{\,n-j}(ws^{n})}{\mathrm{d}x^{\,n-j}}\Bigr]_{a}^{b},

which vanishes by Lemma B because n−j<nn-j<n for j≥1j\ge1. Carrying this out kk times leaves

(−1)k∫abq(k)(x) d n−k(wsn)dx n−k dx=(−1)kC∫abddx[d n−k−1(wsn)dx n−k−1]dx,(-1)^{k}\int_{a}^{b}q^{(k)}(x)\,\frac{\mathrm{d}^{\,n-k}(ws^{n})}{\mathrm{d}x^{\,n-k}}\,\mathrm{d}x =(-1)^{k}C\int_{a}^{b}\frac{\mathrm{d}}{\mathrm{d}x}\Bigl[\frac{\mathrm{d}^{\,n-k-1}(ws^{n})}{\mathrm{d}x^{\,n-k-1}}\Bigr]\mathrm{d}x ,

where C=q(k)C=q^{(k)} is the constant kkth derivative of a degree-kk polynomial, and where n−k−1≥0n-k-1\ge0 because k<nk<n. The last integral is [d n−k−1(wsn)/dx n−k−1]ab=0\bigl[\mathrm{d}^{\,n-k-1}(ws^{n})/\mathrm{d}x^{\,n-k-1}\bigr]_{a}^{b}=0, again by Lemma B. So ∫abqFnw=0\int_{a}^{b}qF_{n}w=0.

The degree is exactly nn. Write Fn=p≤n−1+kn(n)xnF_{n}=p_{\le n-1}+k_{n}^{(n)}x^{n}, multiply by wFnwF_{n} and integrate:

∫abFn2 w=∫abp≤n−1Fn w⏟=0 by the previous step+  kn(n)∫abxnFn w.\int_{a}^{b}F_{n}^{2}\,w =\underbrace{\int_{a}^{b}p_{\le n-1}F_{n}\,w}_{=0\ \text{by the previous step}} +\;k_{n}^{(n)}\int_{a}^{b}x^{n}F_{n}\,w .

The left side is strictly positive, because w>0w>0 and FnF_{n} is a nonzero polynomial (it is nonzero: if Fn≡0F_{n}\equiv0 then dn(wsn)/dxn≡0\mathrm{d}^{n}(ws^{n})/\mathrm{d}x^{n}\equiv0, making wsnws^{n} a polynomial of degree <n<n, which contradicts Lemma A at m=n−1m=n-1 unless ww vanishes somewhere). Hence the right side is nonzero, so kn(n)≠0k_{n}^{(n)}\neq0 and deg⁡Fn=n\deg F_{n}=n.

Combining the last two steps: (Fn)(F_{n}) has deg⁡Fn=n\deg F_{n}=n and is orthogonal to all lower degrees, so by the uniqueness statement proved above it is the orthogonal family for ww, standardised by the choice of KnK_{n}. The differential equation of Proposition 8.1.4, the equivalence with Arfken's route and the classification into three cases are reported from Hassani and from Arfken, Weber and Harris as cited; we have not reproved them. □\square

Remark(Where the hypotheses are used)Original

w>0w>0 on all of (a,b)(a,b). Used twice: to make ∥q∥w>0\lVert q\rVert_{w}>0 for every nonzero polynomial, so that Gram-Schmidt never divides by zero and hn>0h_{n}>0 (needed in every step of the recurrence proof); and in Step 4 of the zeros proof. Weakening it to w≥0w\ge0 breaks the location of the zeros. Take (a,b)=(−1,1)(a,b)=(-1,1) with w=1w=1 on (−1,0)(-1,0) and w=0w=0 on [0,1)[0,1). The resulting "orthogonal polynomials" are the Legendre polynomials transplanted to (−1,0)(-1,0); all of their zeros lie in (−1,0)(-1,0), and the half-interval (0,1)(0,1) contains none, so "every zero lies in (a,b)(a,b)" is no longer the sharp statement it was. If instead ww is replaced by a measure with only NN atoms, then PN\mathcal{P}_{N} has dimension N+1N+1 but the space has dimension NN: the Gram-Schmidt step at index NN divides by zero and the family stops at pN−1p_{N-1}.

Finite moments. Used to put the monomials in Lw2L^{2}_{w} at all. For w(x)=(1+x2)−1w(x)=(1+x^{2})^{-1} on R\mathbb{R} one has μ0=π<∞\mu_{0}=\pi<\infty, so ww is integrable and strictly positive, but μ2=∫x2w=∞\mu_{2}=\int x^{2}w=\infty: the function xx is not in Lw2L^{2}_{w}, and the construction produces only p0p_{0}. Hassani's strict positivity and integrability are therefore not enough on an infinite interval, and the moment condition in our definition is what repairs this.

Orthogonality to all lower degrees (not merely to the other pmp_{m}). This is the only input to Step 2 of the recurrence proof and Step 5 of the zeros proof. It is why the recurrence is three-term and not (n+2)(n+2)-term: the degree of xpmxp_{m} must stay below nn.

ww real and the family real. Step 3 of the zeros proof is an argument about sign, and Step 2 of the recurrence proof moves the real variable xx across the inner product. With a complex-valued weight neither survives, and the zeros need not be real.

Nothing about smoothness of ww. No derivative of ww is used in any of the three theorems. Smoothness enters only in the Rodrigues construction, where hypothesis (H) makes (ws)′(ws)' exist and equal ww times a polynomial; without it Lemma A has no meaning and the closed form does not exist, although the family still does.

Method(Constructing the family and its recurrence for a given weight)Original

Given a weight ww on (a,b)(a,b) and a target degree NN:

  1. Compute the moments μ0,…,μ2N\mu_{0},\dots,\mu_{2N}. Stop here if any is infinite: the weight is not admissible.
  2. Decide the standardisation. Monic is the cheapest; if the family is to match a book's tables, use the book's rule (for example pn(1)=1p_{n}(1)=1) and rescale at the end.
  3. Start the recurrence. p0=1p_{0}=1, h0=μ0h_{0}=\mu_{0}, B0=μ1/μ0B_{0}=\mu_{1}/\mu_{0}, p1(x)=x−B0p_{1}(x)=x-B_{0}, h1=μ2−μ12/μ0h_{1}=\mu_{2}-\mu_{1}^{2}/\mu_{0}.
  4. Iterate (∗∗)(\ast\ast). For n=1,…,N−1n=1,\dots,N-1: form (xpn,pn)w(xp_{n},p_{n})_{w} and hnh_{n} as integrals against ww (or as combinations of the moments), set Bn=(xpn,pn)w/hnB_{n}=(xp_{n},p_{n})_{w}/h_{n} and Cn=hn/hn−1C_{n}=h_{n}/h_{n-1}, and put pn+1=(x−Bn)pn−Cnpn−1p_{n+1}=(x-B_{n})p_{n}-C_{n}p_{n-1}. This is cheaper and far more stable than orthogonalising the monomials directly, whose Gram matrix is notoriously ill-conditioned; see Solving linear systems numerically.
  5. Check as you go. Cn>0C_{n}>0 always; and pnp_{n} must have nn simple zeros inside (a,b)(a,b). Either failing is a sign of arithmetic error.
  6. For the nodes and weights of the nn-point Gauss rule, take the zeros x1,…,xnx_{1},\dots,x_{n} of pnp_{n} from step 5 and solve the n×nn\times n linear system ∑iWixik=μk\sum_{i}W_{i}x_{i}^{k}=\mu_{k}, k=0,…,n−1k=0,\dots,n-1, for the quadrature weights WiW_{i}.
Example(The family for w(x)=xw(x)=x on (0,1)(0,1), and its two-point Gauss rule)Original

Take (a,b)=(0,1)(a,b)=(0,1) and w(x)=xw(x)=x: strictly positive on the open interval, with moments μk=∫01xk+1dx=1/(k+2)\mu_{k}=\int_{0}^{1}x^{k+1}\mathrm{d}x=1/(k+2), all finite. We build the monic family.

Step 1: p0p_{0} and p1p_{1}. p0=1p_{0}=1, h0=μ0=12h_{0}=\mu_{0}=\tfrac12, and

B0=μ1μ0=1/31/2=23,p1(x)=x−23.B_{0}=\frac{\mu_{1}}{\mu_{0}}=\frac{1/3}{1/2}=\frac{2}{3}, \qquad p_{1}(x)=x-\frac{2}{3}.

Its squared norm is

h1=∫01(x−23)2x dx=∫01(x3−43x2+49x)dx=14−49+29=14−29=136.h_{1}=\int_{0}^{1}\Bigl(x-\tfrac23\Bigr)^{2}x\,\mathrm{d}x =\int_{0}^{1}\Bigl(x^{3}-\tfrac43x^{2}+\tfrac49x\Bigr)\mathrm{d}x =\tfrac14-\tfrac49+\tfrac29=\tfrac14-\tfrac29=\tfrac{1}{36}.

Step 2: p2p_{2} by Gram-Schmidt. Orthogonalising x2x^{2} against p1p_{1} and p0p_{0} needs

(x2,p1)w=∫01x2(x−23)x dx=15−23⋅14=130,(x2,p0)w=∫01x3 dx=14,(x^{2},p_{1})_{w}=\int_{0}^{1}x^{2}\Bigl(x-\tfrac23\Bigr)x\,\mathrm{d}x=\tfrac15-\tfrac23\cdot\tfrac14=\tfrac1{30}, \qquad (x^{2},p_{0})_{w}=\int_{0}^{1}x^{3}\,\mathrm{d}x=\tfrac14,

so

p2=x2−1/301/36(x−23)−1/41/2=x2−65x+45−12=x2−65x+310.p_{2}=x^{2}-\frac{1/30}{1/36}\Bigl(x-\tfrac23\Bigr)-\frac{1/4}{1/2} = x^{2}-\frac{6}{5}x+\frac{4}{5}-\frac12 = x^{2}-\frac{6}{5}x+\frac{3}{10}.

Step 3: the same p2p_{2} from the recurrence. Here

(xp1,p1)w=∫01x2(x−23)2dx=15−43⋅14+49⋅13=2135,(xp_{1},p_{1})_{w}=\int_{0}^{1}x^{2}\Bigl(x-\tfrac23\Bigr)^{2}\mathrm{d}x =\tfrac15-\tfrac43\cdot\tfrac14+\tfrac49\cdot\tfrac13=\tfrac{2}{135},

so B1=(2/135)/(1/36)=815B_{1}=(2/135)/(1/36)=\tfrac{8}{15} and C1=h1/h0=(1/36)/(1/2)=118C_{1}=h_{1}/h_{0}=(1/36)/(1/2)=\tfrac1{18}. Then

(x−815)(x−23)−118=x2−(815+1015)x+1645−118=x2−65x+310,(x-\tfrac{8}{15})\Bigl(x-\tfrac23\Bigr)-\tfrac1{18} = x^{2}-\Bigl(\tfrac{8}{15}+\tfrac{10}{15}\Bigr)x+\tfrac{16}{45}-\tfrac{1}{18} = x^{2}-\tfrac65x+\tfrac{3}{10},

since 1645−118=3290−590=2790=310\tfrac{16}{45}-\tfrac1{18}=\tfrac{32}{90}-\tfrac{5}{90}=\tfrac{27}{90}=\tfrac{3}{10}: the recurrence reproduces p2p_{2} exactly, as (∗∗)(\ast\ast) requires. One more turn, with h2=∫01p22x dx=1600h_{2}=\int_{0}^{1}p_{2}^{2}x\,\mathrm{d}x=\tfrac1{600}, gives B2=1835B_{2}=\tfrac{18}{35}, C2=h2/h1=350C_{2}=h_{2}/h_{1}=\tfrac{3}{50} and

p3(x)=x3−127x2+67x−435.p_{3}(x)=x^{3}-\tfrac{12}{7}x^{2}+\tfrac67x-\tfrac4{35}.

Step 4: the zeros. From 10x2−12x+3=010x^{2}-12x+3=0,

x1,2=6∓610,x1=0.355051…,x2=0.844949…,x_{1,2}=\frac{6\mp\sqrt{6}}{10}, \qquad x_{1}=0.355051\ldots,\quad x_{2}=0.844949\ldots,

two simple zeros, both inside (0,1)(0,1), as the zeros theorem demands. (p3p_{3} likewise has three: 0.2123410.212341, 0.5905330.590533, 0.9114120.911412.)

Step 5: the two-point Gauss rule. Take the two zeros as nodes and solve W1+W2=μ0=12W_{1}+W_{2}=\mu_{0}=\tfrac12, W1x1+W2x2=μ1=13W_{1}x_{1}+W_{2}x_{2}=\mu_{1}=\tfrac13. Subtracting x2x_{2} times the first from the second,

W1=13−12x2x1−x2=(2−36)/60−6/5=9−636=0.1819586…,W2=9+636=0.3180414….W_{1}=\frac{\tfrac13-\tfrac12x_{2}}{x_{1}-x_{2}} =\frac{(2-3\sqrt6)/60}{-\sqrt6/5} =\frac{9-\sqrt6}{36}=0.1819586\ldots, \qquad W_{2}=\frac{9+\sqrt6}{36}=0.3180414\ldots .

Step 6: the checkable number. The rule ∫01f(x) x dx≈W1f(x1)+W2f(x2)\int_{0}^{1}f(x)\,x\,\mathrm{d}x\approx W_{1}f(x_{1})+W_{2}f(x_{2}) should be exact for ff of degree ≤3\le 3 and fail at degree 44. For f(x)=x3f(x)=x^{3} the true value is ∫01x4dx=15\int_{0}^{1}x^{4}\mathrm{d}x=\tfrac15, and

W1x13+W2x23=9−636(6−610)3+9+636(6+610)3=15,W_{1}x_{1}^{3}+W_{2}x_{2}^{3} = \frac{9-\sqrt6}{36}\Bigl(\frac{6-\sqrt6}{10}\Bigr)^{3}+\frac{9+\sqrt6}{36}\Bigl(\frac{6+\sqrt6}{10}\Bigr)^{3} = \frac{1}{5} ,

exactly — in decimals 0.1819586×0.0447582+0.3180414×0.6032418=0.0081441+0.1918559=0.20000000.1819586\times0.0447582+0.3180414\times0.6032418=0.0081441+0.1918559=0.2000000. At degree 44 the rule gives 33200=0.165\tfrac{33}{200}=0.165 against the true ∫01x5dx=16=0.16‾\int_{0}^{1}x^{5}\mathrm{d}x=\tfrac16=0.1\overline{6}, so degree 3=2⋅2−13=2\cdot2-1 is indeed the last exact degree. Every quantity in this example was checked symbolically.

ApplicationAdaptedAdapted from Riley, Hobson & Bence (2006), §27.4.3, pp. 1005–1007

Gaussian quadrature. The zeros theorem is the reason Gauss rules exist. Riley, Hobson and Bence describe the scheme: to integrate over (a,b)(a,b) against a weight ww, take the nn nodes to be the zeros of the nnth orthogonal polynomial for ww — which the theorem guarantees are nn distinct points inside the interval, so that the integrand is only ever sampled where it lives — interpolate the integrand at those nodes, and integrate the interpolant. The resulting nn-point rule is exact for every polynomial of degree up to 2n−12n-1, twice what nn equally spaced points buy. Their Gauss-Legendre case (w≡1w\equiv1 on [−1,1][-1,1], nodes the zeros of PnP_{n}, weights 2/[(1−xi2)∣Pn′(xi)∣2]2/[(1-x_{i}^{2})\lvert P_{n}'(x_{i})\rvert^{2}]) is the workhorse of finite-element and spectral codes; their three-point rule applied to ∫01(1+x2)−1dx\int_{0}^{1}(1+x^{2})^{-1}\mathrm{d}x returns 0.785270.78527 against the exact π/4=0.78540\pi/4=0.78540, where the trapezium and Simpson rules on the same three points give 0.77500.7750 and 0.78330.7833. The weight is chosen to absorb whatever is awkward about the integrand: e−x\mathrm{e}^{-x} on (0,∞)(0,\infty) for Gauss-Laguerre, e−x2\mathrm{e}^{-x^{2}} on R\mathbb{R} for Gauss-Hermite, (1−x2)±1/2(1-x^{2})^{\pm1/2} for Gauss-Chebyshev. See Numerical quadrature.

Separation of variables. Every family in the classification table above arises as the eigenfunctions of a Sturm-Liouville problem (Sturm-Liouville eigenvalue problems) obtained by separating a partial differential equation in a coordinate system adapted to the geometry: Legendre from the polar angle in spherical coordinates (Separation of variables in standard coordinates), Laguerre from the radial Coulomb problem, Hermite from the harmonic oscillator. The orthogonality relation of this unit is then exactly what makes the expansion coefficients computable, and the three-term recurrence is what makes the special-function values computable — in practice one never sums a Rodrigues formula, one iterates (∗)(\ast) upward from p0p_{0} and p1p_{1}.

Least squares and spectral methods. Because pn⊥Pn−1p_{n}\perp\mathcal{P}_{n-1}, the best degree-NN approximation to ff in ∥⋅∥w\lVert\cdot\rVert_{w} is the truncation ∑n≤N(f,pn)wpn/hn\sum_{n\le N}(f,p_{n})_{w}p_{n}/h_{n} of a single expansion: raising NN adds a term and changes none of the earlier ones, which is false for a fit in the monomial basis. This is why climate and fluid codes carry spectral coefficients rather than nodal polynomial fits, and why a weighted polynomial fit to data is computed in an orthogonal basis.

Definition

Legendre polynomials

Adapted

The polynomial solutions of Legendre's equation, with the four descriptions that are used interchangeably in practice - the differential equation, Rodrigues' formula, the generating function and the three-term recurrence - together with their orthogonality on [-1, 1] and the normalisation 2/(2l+1).

Prerequisites: General properties of orthogonal polynomials, The Frobenius method, Sturm-Liouville eigenvalue problems

Definition(Legendre's equation and the Legendre polynomials)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.1, eq. (18.1), p. 577, and §18.1.1, pp. 578–579; Hassani (2013), §8.4.3, eqs. (8.19) and (8.22), pp. 250–252; Strauss (2008), §10.6, eq. (7), p. 291

Legendre's equation is

(1−x2)y′′−2xy′+ℓ(ℓ+1)y=0,(L)(1-x^{2})y''-2xy'+\ell(\ell+1)y=0 , \tag{L}

with ℓ\ell a given real parameter. Equivalently, in the Sturm-Liouville form of Sturm-Liouville eigenvalue problems,

ddx[(1−x2)dydx]+ℓ(ℓ+1)y=0,\frac{\mathrm{d}}{\mathrm{d}x}\Bigl[(1-x^{2})\frac{\mathrm{d}y}{\mathrm{d}x}\Bigr]+\ell(\ell+1)y=0 ,

so p=1−x2p=1-x^{2}, q=0q=0, weight ρ=1\rho=1, eigenvalue λ=ℓ(ℓ+1)\lambda=\ell(\ell+1), on the natural interval [−1,1][-1,1]. The equation has regular singular points at x=−1,1,∞x=-1,1,\infty and an ordinary point at x=0x=0. Riley, Hobson and Bence call any solution of (L) a Legendre function; in applications x=cos⁡θx=\cos\theta is the cosine of a polar angle, so −1≤x≤1-1\le x\le1.

For ℓ=n\ell=n a non-negative integer, (L) has a polynomial solution of degree nn, unique up to scale. The Legendre polynomial PnP_{n} is that solution standardised by

Pn(1)=1.P_{n}(1)=1 .

This standardisation forces Pn(−1)=(−1)nP_{n}(-1)=(-1)^{n}. Explicitly,

P0=1,P1=x,P2=12(3x2−1),P3=12(5x3−3x),P_{0}=1,\quad P_{1}=x,\quad P_{2}=\tfrac12(3x^{2}-1),\quad P_{3}=\tfrac12(5x^{3}-3x),P4=18(35x4−30x2+3),P5=18(63x5−70x3+15x).P_{4}=\tfrac18(35x^{4}-30x^{2}+3),\quad P_{5}=\tfrac18(63x^{5}-70x^{3}+15x).

The second solution of (L) for integer ℓ\ell is an infinite series, singular at x=±1x=\pm1; it is the subject of Legendre functions of the second kind and plays no part here.

In the language of General properties of orthogonal polynomials, (Pn)(P_{n}) is the orthogonal family on (−1,1)(-1,1) for the weight w≡1w\equiv1, standardised at x=1x=1 instead of monically. Kreyszig reaches it from the other end: he applies Gram-Schmidt to 1,x,x2,…1,x,x^{2},\dots in L2[−1,1]L^{2}[-1,1] and identifies the output, up to the factor (2n+1)/2\sqrt{(2n+1)/2} that makes it a unit vector, as PnP_{n}. Hassani reaches it as the Jacobi case μ=ν=0\mu=\nu=0 of the general Rodrigues construction, with standardisation constant Kn=(−1)n2nn!K_{n}=(-1)^{n}2^{n}n!.

Definition(The four descriptions and the orthogonality relation)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.1.2, eqs. (18.9), (18.11), (18.12), (18.15) and (18.23)–(18.27), pp. 581–586; Strauss (2008), §10.6, eqs. (4)–(8), pp. 290–292; Kreyszig (1978), §3.7-1, eqs. (2a)–(3), pp. 177–178; Hassani (2013), §8.4.3, eqs. (8.19)–(8.20), pp. 250–252

Let n,l,kn,l,k denote non-negative integers. The following hold.

  1. Rodrigues' formula. Pl(x)=12l l! dldxl(x2−1)l.P_{l}(x)=\frac{1}{2^{l}\,l!}\,\frac{\mathrm{d}^{l}}{\mathrm{d}x^{l}}\bigl(x^{2}-1\bigr)^{l} . (Hassani writes the same thing as (−1)l(2ll!)−1dl(1−x2)l/dxl(-1)^{l}(2^{l}l!)^{-1}\mathrm{d}^{l}(1-x^{2})^{l}/\mathrm{d}x^{l}.)
  2. Generating function. For ∣x∣≤1\lvert x\rvert\le1 and ∣h∣<1\lvert h\rvert<1, G(x,h)=(1−2xh+h2)−1/2=∑l=0∞Pl(x) hl.G(x,h)=\bigl(1-2xh+h^{2}\bigr)^{-1/2}=\sum_{l=0}^{\infty}P_{l}(x)\,h^{l} . The closed endpoints x=±1x=\pm1 are included on the authority of Riley, Hobson and Bence, who evaluate G(1,h)=(1−h)−1G(1,h)=(1-h)^{-1} on p. 585 and read the normalisation off it; Strauss states his eq. (8) for the open range ∣z∣<1\lvert z\rvert<1 (and then also sets z=1z=1 in it).
  3. Three-term recurrence. (l+1)Pl+1(x)=(2l+1)x Pl(x)−l Pl−1(x),P0=1, P1=x.(l+1)P_{l+1}(x)=(2l+1)x\,P_{l}(x)-l\,P_{l-1}(x), \qquad P_{0}=1,\ P_{1}=x . Among the derivative recurrences, the ones most used are (1−x2)Pl′=l(Pl−1−xPl),(2l+1)Pl=Pl+1′−Pl−1′.(1-x^{2})P_{l}'=l\bigl(P_{l-1}-xP_{l}\bigr), \qquad (2l+1)P_{l}=P_{l+1}'-P_{l-1}' .
  4. Orthogonality and normalisation. ∫−11Pl(x)Pk(x) dx=22l+1 δlk.\int_{-1}^{1}P_{l}(x)P_{k}(x)\,\mathrm{d}x=\frac{2}{2l+1}\,\delta_{lk} . Consequently ((2l+1)/2 Pl)l≥0\bigl(\sqrt{(2l+1)/2}\,P_{l}\bigr)_{l\ge0} is orthonormal in L2[−1,1]L^{2}[-1,1], and (Kreyszig) it is total there: every f∈L2[−1,1]f\in L^{2}[-1,1] has f=∑l=0∞alPl,al=2l+12∫−11f(x)Pl(x) dx,f=\sum_{l=0}^{\infty}a_{l}P_{l}, \qquad a_{l}=\frac{2l+1}{2}\int_{-1}^{1}f(x)P_{l}(x)\,\mathrm{d}x , the series converging in L2L^{2}. Riley, Hobson and Bence state the expansion for any ff satisfying the Dirichlet conditions of a Fourier series.
  5. Parity and zeros. Pl(−x)=(−1)lPl(x)P_{l}(-x)=(-1)^{l}P_{l}(x). Inside (−1,1)(-1,1), PlP_{l} has exactly ll zeros, all simple; and (Strauss) for 1≤k≤l1\le k\le l the kkth derivative dkPl/dxk\mathrm{d}^{k}P_{l}/\mathrm{d}x^{k} has exactly l−kl-k zeros in (−1,1)(-1,1), none of these derivatives vanishing at x=±1x=\pm1.
  6. Values at the ends and the centre. Pl(1)=1P_{l}(1)=1, Pl(−1)=(−1)lP_{l}(-1)=(-1)^{l}, and (Hassani) P2m(0)=(−1)m(2m)!22m(m!)2,P2m+1(0)=0.P_{2m}(0)=\frac{(-1)^{m}(2m)!}{2^{2m}(m!)^{2}}, \qquad P_{2m+1}(0)=0 .
Idea of the proofOriginal

The four descriptions are four ways to pin down "the degree-ll polynomial solution of (L) with value 11 at x=1x=1", and each proof is the same two moves: show the candidate solves (L), then check it at x=1x=1. Uniqueness does the rest, because (L) has only a one-dimensional space of solutions regular at both ends.

For Rodrigues' formula the first move is Leibniz: u=(x2−1)lu=(x^{2}-1)^{l} satisfies the first-order relation (x2−1)u′=2lxu(x^{2}-1)u'=2lxu, and differentiating that l+1l+1 times turns it into (L) for u(l)u^{(l)}. For the generating function, differentiating GG in xx and in hh produces two identities between ∑Plhl\sum P_{l}h^{l} and ∑Pl′hl\sum P_{l}'h^{l}; eliminating between them yields (L) term by term. The recurrence falls out of the same two identities without any extra work.

Orthogonality needs nothing about Legendre polynomials at all: it is the Sturm-Liouville argument, ∫(pyl′)′yk−∫(pyk′)′yl\int(p y_{l}')'y_{k}-\int(py_{k}')'y_{l} integrated by parts, where the boundary terms die because p=1−x2p=1-x^{2} vanishes at both ends. The value 2/(2l+1)2/(2l+1) is a separate computation, and there are two cheap routes: integrate Rodrigues' formula by parts ll times (Riley-Hobson-Bence, Kreyszig), or play the recurrence against itself (Strauss) and never integrate anything.

Proof(Rodrigues' formula)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.1.2, pp. 581–582; Strauss (2008), §10.6, eq. (7) and Exercise 2, p. 291

Write u=(x2−1)lu=(x^{2}-1)^{l} and u(m)=dmu/dxmu^{(m)}=\mathrm{d}^{m}u/\mathrm{d}x^{m}.

Step 1: a first-order relation. Differentiating, u′=2lx(x2−1)l−1u'=2lx(x^{2}-1)^{l-1}, hence

(x2−1) u′−2l x u=0.(x^{2}-1)\,u' - 2l\,x\,u = 0 .

Step 2: differentiate it l+1l+1 times. Apply Leibniz's rule to each product. In (x2−1)u′(x^{2}-1)u' only three terms survive, because x2−1x^{2}-1 has vanishing third derivative; in xuxu only two:

(x2−1)u(l+2)+(l+1)⋅2x u(l+1)+(l+1)l2⋅2 u(l)−2l[x u(l+1)+(l+1)u(l)]=0.(x^{2}-1)u^{(l+2)}+(l+1)\cdot 2x\,u^{(l+1)}+\frac{(l+1)l}{2}\cdot2\,u^{(l)} -2l\bigl[x\,u^{(l+1)}+(l+1)u^{(l)}\bigr]=0 .

Collecting,

(x2−1)u(l+2)+2x u(l+1)−l(l+1)u(l)=0,(x^{2}-1)u^{(l+2)}+2x\,u^{(l+1)}-l(l+1)u^{(l)}=0 ,

and multiplying by −1-1 this is precisely (L) with y=u(l)y=u^{(l)} and ℓ=l\ell=l.

Step 3: u(l)u^{(l)} is a polynomial of degree ll, regular at ±1\pm1. uu is a polynomial of degree 2l2l, so u(l)u^{(l)} is a polynomial of degree ll; being a polynomial it is regular everywhere, in particular at the singular points x=±1x=\pm1 of (L). The space of such solutions is one-dimensional, so u(l)=clPlu^{(l)}=c_{l}P_{l} for some constant clc_{l}.

Step 4: evaluate at x=1x=1 to get clc_{l}. Write u=(x−1)l(x+1)lu=(x-1)^{l}(x+1)^{l} and apply Leibniz:

u(l)=∑r=0l(lr)dr(x−1)ldxr dl−r(x+1)ldxl−r.u^{(l)}=\sum_{r=0}^{l}\binom{l}{r}\frac{\mathrm{d}^{r}(x-1)^{l}}{\mathrm{d}x^{r}}\,\frac{\mathrm{d}^{l-r}(x+1)^{l}}{\mathrm{d}x^{l-r}} .

At x=1x=1 every term with r<lr<l carries a surviving positive power of x−1x-1 and so vanishes; the term r=lr=l gives l!⋅(x+1)ll!\cdot(x+1)^{l}, that is l! 2ll!\,2^{l}. Hence u(l)(1)=2ll!u^{(l)}(1)=2^{l}l!, and since Pl(1)=1P_{l}(1)=1 we get cl=2ll!c_{l}=2^{l}l!, which is Rodrigues' formula. □\square

Proof(Generating function and recurrences)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.1.2, eqs. (18.15)–(18.27), pp. 584–586; Strauss (2008), §10.6, eq. (8), p. 292

Let G(x,h)=(1−2xh+h2)−1/2G(x,h)=(1-2xh+h^{2})^{-1/2}. For ∣x∣≤1\lvert x\rvert\le1 the quadratic 1−2xh+h21-2xh+h^{2} has no zero with ∣h∣<1\lvert h\rvert<1 (its roots are x±i1−x2x\pm\mathrm{i}\sqrt{1-x^{2}}, of modulus 11), so GG is analytic in hh on the unit disc and has a Taylor expansion G=∑l≥0Ql(x)hlG=\sum_{l\ge0}Q_{l}(x)h^{l} there, with coefficients QlQ_{l} to be identified.

Step 1: two differentiated forms. Differentiating GG gives

∂G∂x=h (1−2xh+h2)−3/2=∑lQl′hl,∂G∂h=(x−h)(1−2xh+h2)−3/2=∑llQlhl−1.\frac{\partial G}{\partial x}=h\,(1-2xh+h^{2})^{-3/2}=\sum_{l}Q_{l}'h^{l}, \qquad \frac{\partial G}{\partial h}=(x-h)(1-2xh+h^{2})^{-3/2}=\sum_{l}lQ_{l}h^{l-1} .

Step 2: a relation with no fractional powers. From the first display, (1−2xh+h2)∑lQl′hl=h∑lQlhl(1-2xh+h^{2})\sum_{l}Q_{l}'h^{l}=h\sum_{l}Q_{l}h^{l}. Matching the coefficient of hl+1h^{l+1},

Ql=Ql+1′−2xQl′+Ql−1′.(a)Q_{l}=Q_{l+1}'-2xQ_{l}'+Q_{l-1}' . \tag{a}

Step 3: a second relation. Dividing the two displays of Step 1 eliminates the −3/2-3/2 power: (x−h)∑lQl′hl=h∑llQlhl−1(x-h)\sum_{l}Q_{l}'h^{l}=h\sum_{l}lQ_{l}h^{l-1}. The coefficient of hlh^{l} gives

xQl′−Ql−1′=lQl.(b)xQ_{l}'-Q_{l-1}'=lQ_{l} . \tag{b}

Step 4: eliminate. Read (b) as Ql−1′=xQl′−lQlQ_{l-1}'=xQ_{l}'-lQ_{l} and substitute it into (a); the term Ql−1′Q_{l-1}' disappears and the two multiples of xQl′xQ_{l}' combine to leave

(l+1)Ql=Ql+1′−xQl′.(c)(l+1)Q_{l}=Q_{l+1}'-xQ_{l}' . \tag{c}

Now take (c) with ll replaced by l−1l-1, that is lQl−1=Ql′−xQl−1′lQ_{l-1}=Q_{l}'-xQ_{l-1}', and subtract from it xx times (b), that is lxQl=x2Ql′−xQl−1′lxQ_{l}=x^{2}Q_{l}'-xQ_{l-1}'. The terms xQl−1′xQ_{l-1}' cancel and

l(Ql−1−xQl)=Ql′−x2Ql′=(1−x2)Ql′.(d)l\bigl(Q_{l-1}-xQ_{l}\bigr)=Q_{l}'-x^{2}Q_{l}'=(1-x^{2})Q_{l}' . \tag{d}

Step 5: (L). Differentiate (d) and use (b) once more:

(1−x2)Ql′′−2xQl′=l(Ql−1′−xQl′−Ql)=l(−lQl−Ql)=−l(l+1)Ql,(1-x^{2})Q_{l}''-2xQ_{l}'=l\bigl(Q_{l-1}'-xQ_{l}'-Q_{l}\bigr)=l\bigl(-lQ_{l}-Q_{l}\bigr)=-l(l+1)Q_{l},

where the middle equality is (b) solved for Ql−1′−xQl′=−lQlQ_{l-1}'-xQ_{l}'=-lQ_{l}. So each QlQ_{l} satisfies (L) with ℓ=l\ell=l.

Step 6: identify Ql=PlQ_{l}=P_{l}. Each QlQ_{l} is a polynomial of degree ll (expand GG in powers of 2xh−h22xh-h^{2} by the binomial series and collect; the coefficient of hlh^{l} is a polynomial in xx of degree ll), so it is regular at x=±1x=\pm1 and therefore a multiple of PlP_{l}. Setting x=1x=1,

G(1,h)=[(1−h)2]−1/2=11−h=∑l≥0hl,G(1,h)=\bigl[(1-h)^{2}\bigr]^{-1/2}=\frac{1}{1-h}=\sum_{l\ge0}h^{l},

so Ql(1)=1Q_{l}(1)=1 for every ll, matching the standardisation. Hence Ql=PlQ_{l}=P_{l}, which proves part 2, and (b), (c), (d) become the derivative recurrences of part 3.

(Strauss obtains Step 5 in one stroke instead: GG satisfies the partial differential equation [(1−x2)Gx]x+h(hG)hh=0[(1-x^{2})G_{x}]_{x}+h(hG)_{hh}=0, and inserting the expansion makes each coefficient satisfy (L). Arfken, Weber and Harris go the opposite way, starting from Rodrigues' formula, converting the ll-fold derivative into Cauchy's integral — the Schlaefli integral — summing the resulting geometric series in hh and evaluating one residue.)

Step 7: the three-term recurrence. Substitute ∑Plhl=(1−2xh+h2)−1/2\sum P_{l}h^{l}=(1-2xh+h^{2})^{-1/2} into ∂G/∂h\partial G/\partial h from Step 1, so that

(x−h)∑lPlhl=(1−2xh+h2)∑llPlhl−1.(x-h)\sum_{l}P_{l}h^{l}=(1-2xh+h^{2})\sum_{l}lP_{l}h^{l-1} .

The coefficient of hlh^{l} on the left is xPl−Pl−1xP_{l}-P_{l-1}; on the right it is (l+1)Pl+1−2xlPl+(l−1)Pl−1(l+1)P_{l+1}-2xlP_{l}+(l-1)P_{l-1}. Equating and rearranging,

(l+1)Pl+1=(2l+1)xPl−lPl−1.□(l+1)P_{l+1}=(2l+1)xP_{l}-lP_{l-1} . \qquad\square
Proof(Orthogonality and the value of the norm)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.1.2, pp. 582–583; Strauss (2008), §10.6, eq. (6), pp. 290–291; Kreyszig (1978), §3.7-1, pp. 179–180

Orthogonality. Let l≠kl\neq k. Both PlP_{l} and PkP_{k} satisfy (L), written in Sturm-Liouville form; multiply the equation for PlP_{l} by PkP_{k} and integrate over [−1,1][-1,1]:

∫−11Pk[(1−x2)Pl′]′dx+l(l+1)∫−11PkPl dx=0.\int_{-1}^{1}P_{k}\bigl[(1-x^{2})P_{l}'\bigr]'\mathrm{d}x+l(l+1)\int_{-1}^{1}P_{k}P_{l}\,\mathrm{d}x=0 .

Integrate the first term by parts. The boundary term is [(1−x2)PkPl′]−11=0\bigl[(1-x^{2})P_{k}P_{l}'\bigr]_{-1}^{1}=0, because the factor 1−x21-x^{2} vanishes at both ends while PkP_{k} and Pl′P_{l}' are finite there — this is precisely where the regularity of PlP_{l} at x=±1x=\pm1 is used. So

−∫−11(1−x2)Pk′Pl′ dx+l(l+1)∫−11PkPl=0.-\int_{-1}^{1}(1-x^{2})P_{k}'P_{l}'\,\mathrm{d}x+l(l+1)\int_{-1}^{1}P_{k}P_{l}=0 .

Now swap the roles of ll and kk and subtract: the first integral is symmetric and cancels, leaving

[l(l+1)−k(k+1)]∫−11PkPl dx=0.\bigl[l(l+1)-k(k+1)\bigr]\int_{-1}^{1}P_{k}P_{l}\,\mathrm{d}x=0 .

Since l≠kl\neq k are non-negative integers, l(l+1)≠k(k+1)l(l+1)\neq k(k+1), so the integral vanishes. (Taking k=0k=0 gives the frequently used ∫−11Pl=0\int_{-1}^{1}P_{l}=0 for l≥1l\ge1.) Alternatively this is immediate from part 3 of General properties of orthogonal polynomials once PlP_{l} is known to be the orthogonal family for w≡1w\equiv1; Kreyszig proves it directly by showing ∫−11xmPl dx=0\int_{-1}^{1}x^{m}P_{l}\,\mathrm{d}x=0 for m<lm<l through mm integrations by parts of Rodrigues' formula.

The norm, by the recurrence (Strauss's route). Write (f,g)=∫−11fg(f,g)=\int_{-1}^{1}fg. Take the inner product of the three-term recurrence (l+1)Pl+1=(2l+1)xPl−lPl−1(l+1)P_{l+1}=(2l+1)xP_{l}-lP_{l-1} with Pl−1P_{l-1}. Orthogonality kills (Pl+1,Pl−1)(P_{l+1},P_{l-1}), leaving

0=(2l+1)(xPl,Pl−1)−l (Pl−1,Pl−1).0=(2l+1)(xP_{l},P_{l-1})-l\,(P_{l-1},P_{l-1}) .

Next write the recurrence with ll replaced by l−1l-1, namely lPl=(2l−1)xPl−1−(l−1)Pl−2lP_{l}=(2l-1)xP_{l-1}-(l-1)P_{l-2}, and take its inner product with PlP_{l}; now (Pl−2,Pl)=0(P_{l-2},P_{l})=0 and

l (Pl,Pl)=(2l−1)(xPl−1,Pl).l\,(P_{l},P_{l})=(2l-1)(xP_{l-1},P_{l}) .

The two mixed terms are the same number, since (xPl,Pl−1)=(xPl−1,Pl)(xP_{l},P_{l-1})=(xP_{l-1},P_{l}). Eliminating it,

(2l+1)(Pl,Pl)=(2l−1)(Pl−1,Pl−1),(2l+1)(P_{l},P_{l})=(2l-1)(P_{l-1},P_{l-1}),

valid for l≥1l\ge1. So (2l+1)(Pl,Pl)(2l+1)(P_{l},P_{l}) does not depend on ll; at l=0l=0 it equals 1⋅∫−111 dx=21\cdot\int_{-1}^{1}1\,\mathrm{d}x=2. Hence (Pl,Pl)=2/(2l+1)(P_{l},P_{l})=2/(2l+1), which with the orthogonality is part 4. No integration by parts was needed.

The same number, by Rodrigues' formula. Riley, Hobson and Bence instead put Rodrigues' formula into both factors and integrate by parts ll times; every boundary term carries an undifferentiated factor x2−1x^{2}-1 and dies, so

∫−11Pl2=(−1)l22l(l!)2∫−11(x2−1)ld2ldx2l(x2−1)l dx=(2l)!22l(l!)2∫−11(1−x2)l dx,\int_{-1}^{1}P_{l}^{2} =\frac{(-1)^{l}}{2^{2l}(l!)^{2}}\int_{-1}^{1}(x^{2}-1)^{l}\frac{\mathrm{d}^{2l}}{\mathrm{d}x^{2l}}(x^{2}-1)^{l}\,\mathrm{d}x =\frac{(2l)!}{2^{2l}(l!)^{2}}\int_{-1}^{1}(1-x^{2})^{l}\,\mathrm{d}x ,

using d2l(x2−1)l/dx2l=(2l)!\mathrm{d}^{2l}(x^{2}-1)^{l}/\mathrm{d}x^{2l}=(2l)!. The remaining integral Kl=∫−11(1−x2)lK_{l}=\int_{-1}^{1}(1-x^{2})^{l} satisfies (2l+1)Kl=2lKl−1(2l+1)K_{l}=2lK_{l-1} (integrate by parts against the constant 11 and write 2x2=2−2(1−x2)2x^{2}=2-2(1-x^{2})), whence Kl=22l+1(l!)2/(2l+1)!K_{l}=2^{2l+1}(l!)^{2}/(2l+1)! and the product is 2/(2l+1)2/(2l+1) again.

Completeness. Kreyszig's argument: polynomials are dense in C[−1,1]C[-1,1] in the supremum norm by the Weierstrass approximation theorem, C[−1,1]C[-1,1] is dense in L2[−1,1]L^{2}[-1,1], and span⁡{P0,…,Pm}\operatorname{span}\{P_{0},\dots,P_{m}\} is the set of polynomials of degree ≤m\le m. So the orthonormal system (2l+1)/2 Pl\sqrt{(2l+1)/2}\,P_{l} has dense span, which is totality; see Orthonormal bases in Hilbert space. Part 4's expansion and coefficient formula follow by taking the inner product of the series with PkP_{k} term by term. □\square

Remark(Where the hypotheses are used)Original

ℓ\ell an integer. Nothing else makes the Frobenius series terminate. Riley, Hobson and Bence solve (L) at the ordinary point x=0x=0 and find the two-step recurrence

an+2=n(n+1)−ℓ(ℓ+1)(n+1)(n+2) an,a_{n+2}=\frac{n(n+1)-\ell(\ell+1)}{(n+1)(n+2)}\,a_{n} ,

whose numerator vanishes at n=ℓn=\ell exactly when ℓ\ell is a non-negative integer. Both series otherwise continue for ever, with radius of convergence 11 — the distance to the nearest singular point — and both diverge at x=±1x=\pm1. So for non-integer ℓ\ell there is no solution of (L) bounded on the closed interval [−1,1][-1,1], which is why in a physical problem on a full sphere the separation constant is forced to be ℓ(ℓ+1)\ell(\ell+1) with ℓ\ell integral.

Regularity at x=±1x=\pm1. Used three times: in Step 3 of the Rodrigues proof and Step 6 of the generating-function proof, to identify a candidate with PlP_{l} rather than with some mixture of PlP_{l} and the second solution; and in the orthogonality proof, to kill the boundary term. Without it orthogonality genuinely fails, and it fails for the boundary term's sake and not for want of integrability. Take the second solution at ℓ=0\ell=0,

Q0(x)=12ln⁡1+x1−x=artanh⁡x.Q_{0}(x)=\tfrac12\ln\frac{1+x}{1-x}=\operatorname{artanh}x .

Its singularity at x=±1x=\pm1 is only logarithmic, so Q0Q_{0} is perfectly square integrable: ∫−11Q02 dx=π2/6=1.644934…\int_{-1}^{1}Q_{0}^{2}\,\mathrm{d}x=\pi^{2}/6=1.644934\ldots, a convergent integral. What goes wrong is the one step the proof cannot do without. Since Q0′=(1−x2)−1Q_{0}'=(1-x^{2})^{-1}, the product

(1−x2)Q0′=1(1-x^{2})Q_{0}'=1

identically, so the factor 1−x21-x^{2} no longer quenches anything and the boundary term survives. Running the Sturm-Liouville argument on the pair P1=xP_{1}=x and Q0Q_{0}, with eigenvalues λ1=2\lambda_{1}=2 and λ0=0\lambda_{0}=0,

(λ1−λ0)∫−11P1Q0 dx=[(1−x2)(P1Q0′−Q0P1′)]−11=[x−(1−x2)artanh⁡x]−11=2,(\lambda_{1}-\lambda_{0})\int_{-1}^{1}P_{1}Q_{0}\,\mathrm{d}x =\Bigl[(1-x^{2})\bigl(P_{1}Q_{0}'-Q_{0}P_{1}'\bigr)\Bigr]_{-1}^{1} =\bigl[x-(1-x^{2})\operatorname{artanh}x\bigr]_{-1}^{1}=2 ,

the second piece vanishing at both ends because the linear factor beats the logarithm: (1−x2)artanh⁡x→0(1-x^{2})\operatorname{artanh}x\to0 as x→±1x\to\pm1. Hence ∫−11P1Q0 dx=1\int_{-1}^{1}P_{1}Q_{0}\,\mathrm{d}x=1, not 00: two solutions of (L) with different eigenvalues, both square integrable, and not orthogonal. Regularity at x=±1x=\pm1, not integrability, is the hypothesis doing the work.

p=1−x2p=1-x^{2} vanishing at both ends. This is what makes [−1,1][-1,1] the natural interval and makes the Sturm-Liouville problem singular-but-self-adjoint with no boundary conditions imposed by hand. On a sub-interval [−1,c][-1,c] with c<1c<1 the boundary term at cc does not vanish and the PlP_{l} are not orthogonal there; the correct family for [a,b][a,b] is Kreyszig's affine transplant Pl(1+2x−bb−a)P_{l}\bigl(1+2\frac{x-b}{b-a}\bigr).

∣h∣<1\lvert h\rvert<1 in the generating function. The series is a Taylor series about h=0h=0 and the nearest singularity of GG in hh sits at modulus 11 for ∣x∣≤1\lvert x\rvert\le 1. This is the reason the multipole expansion in the application below must be organised with the smaller radius in the numerator; swapping the two radii gives a divergent series, not a wrong one.

Convergence of the expansion. Part 4 is an L2L^{2} statement, and that is all the orthonormality plus totality gives. Pointwise or uniform convergence is a separate question, settled by the Dirichlet-type conditions Riley, Hobson and Bence invoke; a jump in ff produces the same overshoot near the jump as in a Fourier series.

Method(Computing with Legendre polynomials)Original
  1. To tabulate or evaluate Pl(x)P_{l}(x): never differentiate Rodrigues' formula and never sum the explicit coefficient series. Start from P0=1P_{0}=1, P1=xP_{1}=x and iterate Pl+1=(2l+1)x Pl−l Pl−1l+1.P_{l+1}=\frac{(2l+1)x\,P_{l}-l\,P_{l-1}}{l+1} . Each step costs one multiplication and one subtraction. Upward iteration is numerically safe on ∣x∣≤1\lvert x\rvert\le1, though not because the coefficients are small or of one sign — the Pl−1P_{l-1} term enters with a minus sign, and (2l+1)x/(l+1)(2l+1)x/(l+1) is negative throughout x<0x<0. The reason is that on [−1,1][-1,1] the two solutions of this recurrence, PlP_{l} and the second-kind solution QlQ_{l}, are of the same order of magnitude: each decays only like l−1/2l^{-1/2}, with oscillating sign. So an upward sweep never has to recover a solution that is exponentially small against its companion, and a rounding error made at step ll stays a rounding error. (Outside [−1,1][-1,1], where PlP_{l} grows and QlQ_{l} decays, the same recurrence is stable upward for PlP_{l} and useless upward for QlQ_{l}.) Note also ∣Pl(x)∣≤1\lvert P_{l}(x)\rvert\le1 on [−1,1][-1,1], so nothing overflows.
  2. To expand a given ff on [−1,1][-1,1]: compute al=2l+12∫−11fPla_{l}=\tfrac{2l+1}{2}\int_{-1}^{1}fP_{l}. Use the parity rule first: if ff is even, every odd ala_{l} vanishes, and conversely, which halves the work.
  3. To expand a polynomial exactly: match from the top down. The coefficient of xNx^{N} in PNP_{N} is (2N)!/[2N(N!)2](2N)!/[2^{N}(N!)^{2}]; subtract the right multiple of PNP_{N}, and repeat on the remainder, which now has degree ≤N−1\le N-1. This is exact arithmetic and avoids every integral.
  4. To differentiate or integrate an expansion: use (2l+1)Pl=Pl+1′−Pl−1′(2l+1)P_{l}=P_{l+1}'-P_{l-1}', which integrates to ∫Pl=(Pl+1−Pl−1)/(2l+1)\int P_{l}=\bigl(P_{l+1}-P_{l-1}\bigr)/(2l+1) up to a constant, and (1−x2)Pl′=l(Pl−1−xPl)(1-x^{2})P_{l}'=l(P_{l-1}-xP_{l}) for the derivative.
  5. To solve a boundary-value problem on a sphere with azimuthal symmetry: expand the boundary data in Pl(cos⁡θ)P_{l}(\cos\theta) and match coefficient by coefficient against the general separated solution ∑l(Alrl+Blr−l−1)Pl(cos⁡θ)\sum_{l}(A_{l}r^{l}+B_{l}r^{-l-1})P_{l}(\cos\theta); see Separation of variables in standard coordinates.
Example(Expanding x4x^{4}, and five values of PlP_{l} at x=0.3x=0.3)Original

Part 1: the expansion of x4x^{4}. The function x4x^{4} is even and of degree 44, so only a0,a2,a4a_{0},a_{2},a_{4} can be nonzero. Work from the top, as in step 3 of the method. The leading coefficient of P4=18(35x4−30x2+3)P_{4}=\tfrac18(35x^{4}-30x^{2}+3) is 35/835/8, so

x4−835P4=x4−(x4−3035x2+335)=67x2−335.x^{4}-\frac{8}{35}P_{4}=x^{4}-\Bigl(x^{4}-\frac{30}{35}x^{2}+\frac{3}{35}\Bigr)=\frac{6}{7}x^{2}-\frac{3}{35}.

The leading coefficient of P2=12(3x2−1)P_{2}=\tfrac12(3x^{2}-1) is 3/23/2, so subtract 67⋅23=47\tfrac67\cdot\tfrac23=\tfrac47 times P2P_{2}:

67x2−335−47⋅3x2−12=67x2−335−67x2+27=−335+1035=15.\frac{6}{7}x^{2}-\frac{3}{35}-\frac{4}{7}\cdot\frac{3x^{2}-1}{2} =\frac{6}{7}x^{2}-\frac{3}{35}-\frac{6}{7}x^{2}+\frac{2}{7} =-\frac{3}{35}+\frac{10}{35}=\frac{1}{5}.

Hence

x4=835P4(x)+47P2(x)+15P0(x).x^{4}=\frac{8}{35}P_{4}(x)+\frac{4}{7}P_{2}(x)+\frac{1}{5}P_{0}(x).

Check the top coefficient by the integral formula. By part 4,

a4=92∫−11x4P4(x) dx=92⋅18∫−11(35x8−30x6+3x4)dx=916(709−607+65).a_{4}=\frac{9}{2}\int_{-1}^{1}x^{4}P_{4}(x)\,\mathrm{d}x =\frac{9}{2}\cdot\frac{1}{8}\int_{-1}^{1}\bigl(35x^{8}-30x^{6}+3x^{4}\bigr)\mathrm{d}x =\frac{9}{16}\Bigl(\frac{70}{9}-\frac{60}{7}+\frac{6}{5}\Bigr).

With denominator 315315: 709=2450315\tfrac{70}{9}=\tfrac{2450}{315}, 607=2700315\tfrac{60}{7}=\tfrac{2700}{315}, 65=378315\tfrac65=\tfrac{378}{315}, so the bracket is 2450−2700+378315=128315\tfrac{2450-2700+378}{315}=\tfrac{128}{315} and

a4=916⋅128315=11525040=835,a_{4}=\frac{9}{16}\cdot\frac{128}{315}=\frac{1152}{5040}=\frac{8}{35} ,

agreeing with the algebraic route. As a further check, Parseval in the orthogonal basis must reproduce ∫−11x8dx=29\int_{-1}^{1}x^{8}\mathrm{d}x=\tfrac29:

∑lal222l+1=21(15)2+25(47)2+29(835)2=225+32245+12811025.\sum_{l}a_{l}^{2}\frac{2}{2l+1} =\frac{2}{1}\Bigl(\frac15\Bigr)^{2}+\frac{2}{5}\Bigl(\frac47\Bigr)^{2}+\frac{2}{9}\Bigl(\frac{8}{35}\Bigr)^{2} =\frac{2}{25}+\frac{32}{245}+\frac{128}{11025} .

Over 1102511025: 225=88211025\tfrac{2}{25}=\tfrac{882}{11025}, 32245=144011025\tfrac{32}{245}=\tfrac{1440}{11025}, and 882+1440+128=2450882+1440+128=2450, so the sum is 245011025=29\tfrac{2450}{11025}=\tfrac{2}{9}. Exactly right.

Part 2: the recurrence at x=0.3x=0.3. Start from P0=1P_{0}=1, P1=0.3P_{1}=0.3 and iterate Pl+1=[(2l+1)(0.3)Pl−lPl−1]/(l+1)P_{l+1}=\bigl[(2l+1)(0.3)P_{l}-lP_{l-1}\bigr]/(l+1):

P2=3(0.3)(0.3)−12=0.27−12=−0.365=−73200,P_{2}=\frac{3(0.3)(0.3)-1}{2}=\frac{0.27-1}{2}=-0.365=-\frac{73}{200},P3=5(0.3)(−0.365)−2(0.3)3=−0.5475−0.63=−0.3825=−153400,P_{3}=\frac{5(0.3)(-0.365)-2(0.3)}{3}=\frac{-0.5475-0.6}{3}=-0.3825=-\frac{153}{400},P4=7(0.3)(−0.3825)−3(−0.365)4=−0.803250+1.0954=0.0729375,P_{4}=\frac{7(0.3)(-0.3825)-3(-0.365)}{4}=\frac{-0.803250+1.095}{4}=0.0729375,P5=9(0.3)(0.0729375)−4(−0.3825)5=0.196931250+1.535=0.34538625.P_{5}=\frac{9(0.3)(0.0729375)-4(-0.3825)}{5}=\frac{0.196931250+1.53}{5}=0.34538625 .

Compare with the closed forms: P4(0.3)=18(35(0.0081)−30(0.09)+3)=18(0.2835−2.7+3)=18(0.5835)=0.0729375P_{4}(0.3)=\tfrac18\bigl(35(0.0081)-30(0.09)+3\bigr)=\tfrac18(0.2835-2.7+3)=\tfrac18(0.5835)=0.0729375 and P5(0.3)=18(63(0.00243)−70(0.027)+15(0.3))=18(0.15309−1.89+4.5)=18(2.76309)=0.34538625P_{5}(0.3)=\tfrac18\bigl(63(0.00243)-70(0.027)+15(0.3)\bigr)=\tfrac18(0.15309-1.89+4.5)=\tfrac18(2.76309)=0.34538625. The recurrence and the explicit polynomials agree to the last digit, and as rationals P5(3/10)=276309/800000P_{5}(3/10)=276309/800000. Every number here was checked symbolically.

ApplicationAdaptedAdapted from Riley, Hobson & Bence (2006), §18.1.2, eq. (18.22) and the electrostatic potential that follows it, pp. 585–586; Hassani (2013), §8.5, Example 8.5.2, pp. 256–257

The inverse-distance expansion, and multipoles. Put h=r′/rh=r'/r and x=cos⁡θx=\cos\theta in the generating function, where θ\theta is the angle between two position vectors of lengths r′<rr'<r. Since ∣r−r′∣2=r2+r′2−2rr′cos⁡θ\lvert\mathbf{r}-\mathbf{r}'\rvert^{2}=r^{2}+r'^{2}-2rr'\cos\theta,

1∣r−r′∣=1r[1−2r′rcos⁡θ+(r′r)2]−1/2=1r∑l=0∞(r′r)lPl(cos⁡θ).\frac{1}{\lvert\mathbf{r}-\mathbf{r}'\rvert} =\frac{1}{r}\Bigl[1-2\frac{r'}{r}\cos\theta+\Bigl(\frac{r'}{r}\Bigr)^{2}\Bigr]^{-1/2} =\frac{1}{r}\sum_{l=0}^{\infty}\Bigl(\frac{r'}{r}\Bigr)^{l}P_{l}(\cos\theta) .

Riley, Hobson and Bence read off at once the potential of a point charge qq at r′\mathbf{r}',

V(r)=q4πε0r∑l≥0(r′r)lPl(cos⁡θ),V(\mathbf{r})=\frac{q}{4\pi\varepsilon_{0}r}\sum_{l\ge0}\Bigl(\frac{r'}{r}\Bigr)^{l}P_{l}(\cos\theta),

whose l=0l=0 term is the familiar q/(4πε0r)q/(4\pi\varepsilon_{0}r). Integrating over a charge distribution turns the llth term into the llth multipole moment: this expansion is the multipole expansion (Multipole expansions), and it is why "monopole, dipole, quadrupole" are indexed by Legendre order. The same formula with rr and r′r' exchanged is needed when the field point is the nearer one — a consequence of the ∣h∣<1\lvert h\rvert<1 restriction, not an extra convention. The gravitational version is the standard description of the Earth's, Moon's and Mars's external fields, and the geodetic coefficients J2J_{2} (the equatorial bulge) and its successors are exactly the coefficients of this series.

Angular momentum and selection rules. With x=cos⁡θx=\cos\theta, the Legendre polynomials are the m=0m=0 angular wavefunctions of any central-force problem: they are the θ\theta-dependence of the ss, pzp_{z}, dz2,…d_{z^{2}},\dots orbitals, and the orthogonality relation of part 4 is what makes states of different ll non-interacting. The three-term recurrence, rewritten as a statement about cos⁡θ Pl\cos\theta\,P_{l}, is the origin of the dipole selection rule Δl=±1\Delta l=\pm1: multiplying by cos⁡θ\cos\theta can only move a state to its two neighbours in ll. Extending this to m≠0m\neq0 is the business of Associated Legendre functions and Spherical harmonics.

Approximation, filtering and quadrature. Truncating the Legendre expansion is the optimal degree-NN polynomial approximation to ff in the L2[−1,1]L^{2}[-1,1] norm, which makes it the natural basis for spectral-element discretisations of a one-dimensional domain. The zeros of PnP_{n} are the nodes of the nn-point Gauss-Legendre rule (Numerical quadrature), and the fact that they are nn distinct interior points is part 5 above. Hassani's expansion of the delta function, δ(x)=∑l2l+12Pl(0)Pl(x)\delta(x)=\sum_{l}\tfrac{2l+1}{2}P_{l}(0)P_{l}(x), is the closure relation of the same basis and the starting point for building Green's functions on a sphere (Eigenfunction expansion of Green's functions).

Definition

Associated Legendre functions

Adapted

The regular solutions of the associated Legendre equation, built from the Legendre polynomials by m-fold differentiation and a factor (1-x^2)^{m/2}, with their two orthogonality relations, their behaviour under m -> -m, and an explicit account of the Condon-Shortley sign convention that divides the textbooks.

Prerequisites: Legendre polynomials, General properties of orthogonal polynomials, Sturm-Liouville eigenvalue problems, Orthonormal bases in Hilbert space

Definition(The associated Legendre equation and PlmP_{l}^{m})AdaptedAdapted from Riley, Hobson & Bence (2006), §18.2, eqs. (18.28) and (18.32), pp. 587–588; Arfken, Weber & Harris (2013), §15.4, eqs. (15.68), (15.79) and (15.80), pp. 741–743; Strauss (2008), §10.6, eqs. (10)–(11), pp. 292–293

The associated Legendre equation is

(1−x2)y′′−2xy′+[ℓ(ℓ+1)−m21−x2]y=0,(AL)(1-x^{2})y''-2xy'+\Bigl[\ell(\ell+1)-\frac{m^{2}}{1-x^{2}}\Bigr]y=0 , \tag{AL}

or in Sturm-Liouville form

ddx[(1−x2)dydx]−m21−x2 y+ℓ(ℓ+1) y=0,\frac{\mathrm{d}}{\mathrm{d}x}\Bigl[(1-x^{2})\frac{\mathrm{d}y}{\mathrm{d}x}\Bigr]-\frac{m^{2}}{1-x^{2}}\,y+\ell(\ell+1)\,y=0 ,

with p=1−x2p=1-x^{2}, q=−m2/(1−x2)q=-m^{2}/(1-x^{2}), weight ρ=1\rho=1, eigenvalue λ=ℓ(ℓ+1)\lambda=\ell(\ell+1), natural interval [−1,1][-1,1]. It reduces to Legendre's equation of Legendre polynomials when m=0m=0, and it has the same three regular singular points x=−1,1,∞x=-1,1,\infty. Only m2m^{2} appears, so mm and −m-m give the same equation. In applications mm is an integer, x=cos⁡θx=\cos\theta, and −1≤x≤1-1\le x\le1. Throughout this unit ℓ\ell is the parameter of the equation, so far unrestricted, while ll is reserved for a non-negative integer; part 3 below is the statement that a solution bounded on the whole of [−1,1][-1,1] forces ℓ\ell to be such an integer.

For integers l≥0l\ge0 and ∣m∣≤l\lvert m\rvert\le l, the associated Legendre function of the first kind is

  Plm(x)=(−1)m (1−x2)m/2 dmPl(x)dxm=(−1)m2l l! (1−x2)m/2 dl+mdxl+m(x2−1)l  \boxed{\;P_{l}^{m}(x)=(-1)^{m}\,(1-x^{2})^{m/2}\,\frac{\mathrm{d}^{m}P_{l}(x)}{\mathrm{d}x^{m}} =\frac{(-1)^{m}}{2^{l}\,l!}\,(1-x^{2})^{m/2}\,\frac{\mathrm{d}^{l+m}}{\mathrm{d}x^{l+m}}\bigl(x^{2}-1\bigr)^{l}\;}

for m≥0m\ge0, the second form extending the definition verbatim to −l≤m<0-l\le m<0. The superscript is an index, not a power. Since PlP_{l} has degree ll, we have Plm≡0P_{l}^{m}\equiv0 for ∣m∣>l\lvert m\rvert>l, and Pl0=PlP_{l}^{0}=P_{l}. The second, linearly independent solution of (AL) for integer ll and mm is the associated Legendre function of the second kind QlmQ_{l}^{m}, built the same way from QlQ_{l}; like QlQ_{l} it is singular at x=±1x=\pm1 and is therefore discarded in every problem on a full sphere. The first few are

P11=−(1−x2)1/2,P21=−3x(1−x2)1/2,P22=3(1−x2),P_{1}^{1}=-(1-x^{2})^{1/2},\qquad P_{2}^{1}=-3x(1-x^{2})^{1/2},\qquad P_{2}^{2}=3(1-x^{2}),P31=−32(5x2−1)(1−x2)1/2,P32=15x(1−x2),P33=−15(1−x2)3/2,P_{3}^{1}=-\tfrac32(5x^{2}-1)(1-x^{2})^{1/2},\qquad P_{3}^{2}=15x(1-x^{2}),\qquad P_{3}^{3}=-15(1-x^{2})^{3/2},

or, with x=cos⁡θx=\cos\theta so that (1−x2)1/2=sin⁡θ(1-x^{2})^{1/2}=\sin\theta: P11=−sin⁡θP_{1}^{1}=-\sin\theta, P22=3sin⁡2θP_{2}^{2}=3\sin^{2}\theta, P33=−15sin⁡3θP_{3}^{3}=-15\sin^{3}\theta.

Remark(The sign convention, stated before anything is proved)AdaptedAdapted from Arfken, Weber & Harris (2013), §15.4, eq. (15.78) with footnote 3, p. 743, eqs. (15.87)–(15.88), p. 744, and §15.5, p. 758; Riley, Hobson & Bence (2006), §18.2.1, eq. (18.32) with its footnote, p. 588, §18.2.2, eqs. (18.41)–(18.42), p. 592, and §18.3, eq. (18.45), p. 593; Strauss (2008), §10.6, eq. (11), p. 293, and §10.3, p. 275; Hassani (2013), §13.4, Box 13.4.2, eqs. (13.34)–(13.35), p. 409

Two incompatible definitions of PlmP_{l}^{m} are in circulation, and a formula copied from one book into another is a common source of sign errors.

  • With the phase (the convention used here): Plm=(−1)m(1−x2)m/2dmPl/dxmP_{l}^{m}=(-1)^{m}(1-x^{2})^{m/2}\mathrm{d}^{m}P_{l}/\mathrm{d}x^{m}. Arfken, Weber and Harris insert the factor (−1)m(-1)^{m} to agree with Handbook of Mathematical Functions (AMS-55), which they call the most widely accepted standard, and they note in a footnote that Jackson's Classical Electrodynamics omits it. The factor is the Condon-Shortley phase, which carries the names of E. U. Condon and G. H. Shortley because their book on atomic spectroscopy fixed it as the standard; Arfken, Weber and Harris remark that one of its effects is to alternate the sign with mm among the positive-mm spherical harmonics, and that it was introduced so that the spherical harmonics come out with the usual phase.
  • Without the phase: Plm=(1−x2)m/2dmPl/dxmP_{l}^{m}=(1-x^{2})^{m/2}\mathrm{d}^{m}P_{l}/\mathrm{d}x^{m}. This is Riley, Hobson and Bence's eq. (18.32), Strauss's eq. (11) and Hassani's eq. (13.35): Hassani's function is (−1)m(-1)^{m} times ours, so where we have P11=−sin⁡θP_{1}^{1}=-\sin\theta he has P11=+sin⁡θP_{1}^{1}=+\sin\theta. Two of the three then put the (−1)m(-1)^{m} back into the definition of the spherical harmonic — Riley, Hobson and Bence in eq. (18.45), Hassani in eq. (13.34) of Box 13.4.2 — so their YlmY_{l}^{m} agrees with Arfken, Weber and Harris's. Strauss does not: his spherical harmonic is Ylm(θ,φ)=Pl∣m∣(cos⁡θ) eimφY_{lm}(\theta,\varphi)=P_{l}^{\lvert m\rvert}(\cos\theta)\,\mathrm{e}^{\mathrm{i}m\varphi} (§10.3, p. 275), carrying neither the phase nor any normalising constant, so for odd m>0m>0 his YlmY_{lm} differs from the Condon-Shortley one in sign as well as in scale. See Spherical harmonics.

So the two conventions differ by (−1)m(-1)^{m} in PlmP_{l}^{m} and, Strauss aside, agree in YlmY_{l}^{m}. Everything that is quadratic in PlmP_{l}^{m} — both orthogonality relations and both normalisation constants — is the same in either convention, and so is the equation (AL).

For anything linear in the PlmP_{l}^{m} there is a one-line rule. Write P~lm\tilde P_{l}^{m} for the phase-free function, so that Plm=(−1)mP~lmP_{l}^{m}=(-1)^{m}\tilde P_{l}^{m} — for negative mm as well, since (−1)−m=(−1)m(-1)^{-m}=(-1)^{m} and both conventions use the same negative-mm relation. Substituting this into an identity multiplies the term carrying upper index mm by (−1)m(-1)^{m}, and that factor can be divided out of the whole identity exactly when it is the same in every term. A linear identity in the PlmP_{l}^{m} therefore keeps its written form in both conventions precisely when all the upper indices occurring in it have the same parity, and changes otherwise. The cases worth having in mind:

  • A table entry is a single function: it changes sign for odd mm and not for even mm.
  • The step-in-ll recurrence of part 8 carries the one upper index mm in all three terms, so it is the same in both conventions; Arfken, Weber and Harris print it as (15.88) and Riley, Hobson and Bence as (18.42) in the identical form, which is the form given in part 8.
  • The negative-mm relation of part 4 involves only −m-m and mm, of the same parity, so it too is the same in both conventions; Arfken, Weber and Harris (15.81), Riley, Hobson and Bence (18.33) and Hassani (13.38) all print the same formula.
  • The step-in-mm recurrence of part 8 mixes the upper indices m+1m+1, mm and m−1m-1, so it changes for every m≠0m\neq0: in the phase-free convention the middle term reverses sign, which is why Riley, Hobson and Bence's (18.41) differs from Arfken, Weber and Harris's (15.87) — the form used in part 8 — in the sign of the 2mx2mx term. (At m=0m=0 that term is absent and the two remaining indices ±1\pm1 have the same parity, so there alone the two forms agree.) Among the identities collected in this unit it is the only one whose written form depends on the convention; the other convention-sensitive items are the individual values in the table of the definition, by the first bullet.

There is a further, rarer variant in which Pl−mP_{l}^{-m} is defined to equal PlmP_{l}^{m}, which Riley, Hobson and Bence flag; with that choice mm must be replaced by ∣m∣\lvert m\rvert throughout, as Strauss does in his definition of the spherical harmonics. We do not use it.

Definition(Properties of PlmP_{l}^{m})AdaptedAdapted from Riley, Hobson & Bence (2006), §18.2, eqs. (18.29) and (18.33), pp. 587–589, and §18.2.2, eqs. (18.36)–(18.44), pp. 590–593; Arfken, Weber & Harris (2013), §15.4, eqs. (15.74), (15.81), (15.82), (15.84), (15.87)–(15.88), (15.93), (15.95)–(15.96) and (15.104)–(15.107), pp. 742–748; Strauss (2008), §10.6, eqs. (11)–(12), pp. 292–293

Let l≥0l\ge0 and mm be integers with ∣m∣≤l\lvert m\rvert\le l.

  1. It solves the equation. If uu solves Legendre's equation with parameter ℓ\ell, then (1−x2)∣m∣/2 d∣m∣u/dx∣m∣(1-x^{2})^{\lvert m\rvert/2}\,\mathrm{d}^{\lvert m\rvert}u/\mathrm{d}x^{\lvert m\rvert} solves (AL) with the same ℓ\ell and with that mm. In particular PlmP_{l}^{m} solves (AL) with ℓ=l\ell=l, and it is regular at x=±1x=\pm1.
  2. Shape. Plm(x)=(1−x2)m/2 Plm(x)P_{l}^{m}(x)=(1-x^{2})^{m/2}\,\mathcal{P}_{l}^{m}(x) for m≥0m\ge0, where Plm\mathcal{P}_{l}^{m} is a polynomial of degree l−ml-m. Hence PlmP_{l}^{m} is a polynomial of degree ll when mm is even, and carries a fractional power of 1−x21-x^{2} when mm is odd; either way it is regular on [−1,1][-1,1], and it vanishes at x=±1x=\pm1 unless m=0m=0. Written in terms of these factors, the orthogonality of part 6 reads ∫−11PlmPkm(1−x2)m dx=0\int_{-1}^{1}\mathcal{P}_{l}^{m}\mathcal{P}_{k}^{m}(1-x^{2})^{m}\,\mathrm{d}x=0 for l≠kl\neq k, so at fixed mm the Plm\mathcal{P}_{l}^{m}, l=m,m+1,…l=m,m+1,\dots, are up to scale the orthogonal polynomials of the weight (1−x2)m(1-x^{2})^{m} on (−1,1)(-1,1); by General properties of orthogonal polynomials each of them therefore has exactly l−ml-m zeros, all real, all simple and all interior to (−1,1)(-1,1).
  3. Integrality is forced. Seek the factor P\mathcal{P} of part 2 as a Frobenius series about x=0x=0, P(x)=∑j≥0ajx j+k\mathcal{P}(x)=\sum_{j\ge0}a_{j}x^{\,j+k} with indicial exponent kk and a0≠0a_{0}\neq0. Substituting it into the equation that P\mathcal{P} satisfies, namely (1−x2)P′′−2(m+1)xP′+[λ−m(m+1)]P=0(1-x^{2})\mathcal{P}''-2(m+1)x\mathcal{P}'+[\lambda-m(m+1)]\mathcal{P}=0, which is equation (D) of the proof of part 1 because (ℓ−m)(ℓ+m+1)=λ−m(m+1)(\ell-m)(\ell+m+1)=\lambda-m(m+1), gives the recurrence aj+2=j2+(2m+1)j−λ+m(m+1)(j+1)(j+2) aja_{j+2}=\frac{j^{2}+(2m+1)j-\lambda+m(m+1)}{(j+1)(j+2)}\,a_{j} for its coefficients, this display being the branch belonging to the indicial root k=0k=0; the series diverges at x=±1x=\pm1 unless the numerator vanishes, terminating it. Arfken, Weber and Harris keep the two Frobenius branches apart. For k=0k=0 the numerator must vanish at a non-negative even jj, and it does so at j=l−mj=l-m when λ=l(l+1)\lambda=l(l+1), which requires ll to be an integer, at least mm, and of the same parity as mm. The second indicial root, k=1k=1, supplies by the same argument the admissible ll of the opposite parity to mm. Taken together the two branches give: for each mm the admissible ll run from mm upwards, one for every integer; equivalently, for each ll the admissible mm run from −l-l to ll.
  4. Negative mm. Pl−m(x)=(−1)m(l−m)!(l+m)! Plm(x).P_{l}^{-m}(x)=(-1)^{m}\frac{(l-m)!}{(l+m)!}\,P_{l}^{m}(x). So Pl−mP_{l}^{-m} carries no new information; it is kept because the recurrences in mm produce it.
  5. Parity and special values. Plm(−x)=(−1)l+mPlm(x),Plm(±1)=0  (m≠0),P_{l}^{m}(-x)=(-1)^{l+m}P_{l}^{m}(x), \qquad P_{l}^{m}(\pm1)=0 \ \ (m\neq0), and at the centre Plm(0)={(−1)(l+m)/2 (l+m−1)!!(l−m)!!,l+m even,0,l+m odd.P_{l}^{m}(0)=\begin{cases} \dfrac{(-1)^{(l+m)/2}\,(l+m-1)!!}{(l-m)!!}, & l+m \text{ even},\\[2mm] 0, & l+m \text{ odd}. \end{cases} Also Pmm(x)=(−1)m(2m−1)!! (1−x2)m/2P_{m}^{m}(x)=(-1)^{m}(2m-1)!!\,(1-x^{2})^{m/2} and Pm+1m(x)=(−1)m(2m+1)!! x (1−x2)m/2P_{m+1}^{m}(x)=(-1)^{m}(2m+1)!!\,x\,(1-x^{2})^{m/2}.
  6. Orthogonality in ll at fixed mm, with unit weight. ∫−11Plm(x) Pkm(x) dx=22l+1 (l+m)!(l−m)! δlk,\int_{-1}^{1}P_{l}^{m}(x)\,P_{k}^{m}(x)\,\mathrm{d}x =\frac{2}{2l+1}\,\frac{(l+m)!}{(l-m)!}\,\delta_{lk}, or with x=cos⁡θx=\cos\theta, ∫0πPlm(cos⁡θ)Pkm(cos⁡θ)sin⁡θ dθ=22l+1(l+m)!(l−m)! δlk.\int_{0}^{\pi}P_{l}^{m}(\cos\theta)P_{k}^{m}(\cos\theta)\sin\theta\,\mathrm{d}\theta =\frac{2}{2l+1}\frac{(l+m)!}{(l-m)!}\,\delta_{lk}. Consequently, at fixed m≥0m\ge0, the normalised functions 2l+12(l−m)!(l+m)! Plm\sqrt{\tfrac{2l+1}{2}\tfrac{(l-m)!}{(l+m)!}}\,P_{l}^{m} with l≥ml\ge m are orthonormal in L2(−1,1)L^{2}(-1,1). They are moreover total there — a separate fact, which orthogonality alone does not give, proved in its own short block below — so every f∈L2(−1,1)f\in L^{2}(-1,1) has f=∑l≥malPlm,al=2l+12(l−m)!(l+m)!∫−11f(x)Plm(x) dx,f=\sum_{l\ge m}a_{l}P_{l}^{m}, \qquad a_{l}=\frac{2l+1}{2}\frac{(l-m)!}{(l+m)!}\int_{-1}^{1}f(x)P_{l}^{m}(x)\,\mathrm{d}x , the series converging in L2L^{2}. The sum starts at l=ml=m because Plm=0P_{l}^{m}=0 below that. Riley, Hobson and Bence state this expansion, their eq. (18.38), for any "reasonable" function ((Riley, Hobson & Bence (2006), §18.2.2, eq. (18.38), p. 591)) rather than for a class defined in advance; square integrability is the hypothesis under which the orthonormal-expansion theorem of Orthonormal bases in Hilbert space applies, and it is the one we use.
  7. Orthogonality in mm at fixed ll, with weight (1−x2)−1(1-x^{2})^{-1}. ∫−11Plm(x)Plk(x) dx1−x2=(l+m)!m (l−m)! δmk(m,k≥1).\int_{-1}^{1}P_{l}^{m}(x)P_{l}^{k}(x)\,\frac{\mathrm{d}x}{1-x^{2}} =\frac{(l+m)!}{m\,(l-m)!}\,\delta_{mk} \qquad (m,k\ge1). Both books that state this relation add that it is of no practical use, because in spherical coordinates the φ\varphi boundary condition already separates the different mm.
  8. Generating function and recurrences. For m≥0m\ge0, (2m)! (1−x2)m/22mm! (1−2hx+h2)m+1/2=∑n=0∞P~n+mm(x) hn,\frac{(2m)!\,(1-x^{2})^{m/2}}{2^{m}m!\,(1-2hx+h^{2})^{m+1/2}}=\sum_{n=0}^{\infty}\tilde P_{n+m}^{m}(x)\,h^{n}, where P~\tilde P denotes the phase-free functions of Riley, Hobson and Bence (so P~lm=(−1)mPlm\tilde P_{l}^{m}=(-1)^{m}P_{l}^{m}); this is their eq. (18.40) ((Riley, Hobson & Bence (2006), §18.2.2, eq. (18.40), p. 592)). Arfken, Weber and Harris's generating function ((Arfken, Weber & Harris (2013), §15.4, eq. (15.82), p. 743)) looks different because it generates the polynomial factors Plm=Plm/(1−x2)m/2\mathcal{P}_{l}^{m}=P_{l}^{m}/(1-x^{2})^{m/2} of part 2 rather than the functions themselves: their left-hand side carries (−1)m(2m−1)!!(-1)^{m}(2m-1)!! in place of our (2m)!(1−x2)m/2/2mm!(2m)!(1-x^{2})^{m/2}/2^{m}m!, and (2m−1)!!=(2m)!/2mm!(2m-1)!!=(2m)!/2^{m}m!, so the two agree once the common factor (1−x2)m/2(1-x^{2})^{m/2} and the phase are accounted for. The two most useful recurrences are (2l+1)x Plm=(l+m)Pl−1m+(l−m+1)Pl+1m,(2l+1)x\,P_{l}^{m}=(l+m)P_{l-1}^{m}+(l-m+1)P_{l+1}^{m}, Plm+1+2mx(1−x2)1/2Plm+(l+m)(l−m+1)Plm−1=0,P_{l}^{m+1}+\frac{2mx}{(1-x^{2})^{1/2}}P_{l}^{m}+(l+m)(l-m+1)P_{l}^{m-1}=0 , the first stepping in ll at fixed mm ((Arfken, Weber & Harris (2013), §15.4, eq. (15.88), p. 744); (Riley, Hobson & Bence (2006), §18.2.2, eq. (18.42), p. 592)), the second in mm at fixed ll ((Arfken, Weber & Harris (2013), §15.4, eq. (15.87), p. 744); the phase-free form is Riley, Hobson and Bence's (18.41), which differs in the sign of the middle term — see the remark on the sign convention). Both are valid for negative as well as non-negative mm with the definition adopted here.
Idea of the proofOriginal

Everything comes from the single observation that differentiating Legendre's equation mm times almost gives the associated equation. Differentiating (1−x2)u′′−2xu′+ℓ(ℓ+1)u=0(1-x^{2})u''-2xu'+\ell(\ell+1)u=0 mm times with Leibniz produces

(1−x2)v′′−2x(m+1)v′+(ℓ−m)(ℓ+m+1)v=0,v=dmudxm,(1-x^{2})v''-2x(m+1)v'+(\ell-m)(\ell+m+1)v=0, \qquad v=\frac{\mathrm{d}^{m}u}{\mathrm{d}x^{m}} ,

which differs from (AL) only in the first-derivative coefficient and the eigenvalue. The factor (1−x2)m/2(1-x^{2})^{m/2} is exactly the change of dependent variable that repairs both: it trades the extra −2mxv′-2mxv' for the term −m2y/(1−x2)-m^{2}y/(1-x^{2}). That one calculation proves part 1, and parts 2 and 5 are then bookkeeping about degrees and parities. Part 3 costs more: it needs the Frobenius series at the origin, which diverges at the endpoints unless its coefficient recurrence terminates, and termination is exactly the condition that ℓ\ell be an integer at least ∣m∣\lvert m\rvert.

The orthogonality in ll (part 6) is again the Sturm-Liouville argument: p=1−x2p=1-x^{2} vanishes at both ends, so the boundary terms die, and the qq term is the same for both functions and cancels on subtraction. The orthogonality in mm (part 7) is the same equation read as an eigenvalue problem in −m2-m^{2} with weight (1−x2)−1(1-x^{2})^{-1} — one equation, two self-adjoint readings, two orthogonality relations.

The normalisation constant is the only genuine computation. Put Rodrigues' formula into both factors, integrate by parts l+ml+m times so that all boundary terms vanish, and then notice that of all the terms Leibniz produces in differentiating (1−x2)m dl+m(x2−1)l/dxl+m(1-x^{2})^{m}\,\mathrm{d}^{l+m}(x^{2}-1)^{l}/\mathrm{d}x^{l+m} a further l+ml+m times, exactly one has the degrees right to survive. That term carries the factor (l+m)!/(l−m)!(l+m)!/(l-m)!, and the rest of the integral is the Legendre computation already done.

Proof(Part 1: the reduction to Legendre's equation)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.2, the proof of eq. (18.29), p. 587; Strauss (2008), §10.6, the lines following eq. (11), p. 293

Assume first m≥0m\ge0; the case m<0m<0 then follows because (AL) contains only m2m^{2}.

Step 1: differentiate Legendre's equation mm times. Let uu solve (1−x2)u′′−2xu′+ℓ(ℓ+1)u=0(1-x^{2})u''-2xu'+\ell(\ell+1)u=0 and set v=dmu/dxmv=\mathrm{d}^{m}u/\mathrm{d}x^{m}. Apply dm/dxm\mathrm{d}^{m}/\mathrm{d}x^{m} term by term with Leibniz's rule. The factor 1−x21-x^{2} contributes three terms, −2x-2x contributes two, and the constant one:

(1−x2)v′′+m(−2x)v′+(m2)(−2)v  −  2[xv′+mv]  +  ℓ(ℓ+1)v=0.(1-x^{2})v''+m(-2x)v'+\binom{m}{2}(-2)v \;-\;2\bigl[xv'+mv\bigr] \;+\;\ell(\ell+1)v=0 .

Collecting the v′v' terms gives −2x(m+1)v′-2x(m+1)v', and the vv terms give [ℓ(ℓ+1)−m(m−1)−2m]v=[ℓ(ℓ+1)−m(m+1)]v\bigl[\ell(\ell+1)-m(m-1)-2m\bigr]v=\bigl[\ell(\ell+1)-m(m+1)\bigr]v. Since ℓ(ℓ+1)−m(m+1)=(ℓ−m)(ℓ+m+1)\ell(\ell+1)-m(m+1)=(\ell-m)(\ell+m+1),

(1−x2)v′′−2x(m+1)v′+(ℓ−m)(ℓ+m+1)v=0.(D)(1-x^{2})v''-2x(m+1)v'+(\ell-m)(\ell+m+1)v=0 . \tag{D}

Step 2: restore the dependent variable. Put y=(1−x2)m/2vy=(1-x^{2})^{m/2}v, so v=(1−x2)−m/2yv=(1-x^{2})^{-m/2}y. Differentiating,

v′=(1−x2)−m/2[y′+mx1−x2 y],v'=(1-x^{2})^{-m/2}\Bigl[y'+\frac{mx}{1-x^{2}}\,y\Bigr],v′′=(1−x2)−m/2[y′′+2mx1−x2 y′+m1−x2 y+m(m+2)x2(1−x2)2 y].v''=(1-x^{2})^{-m/2}\Bigl[y''+\frac{2mx}{1-x^{2}}\,y'+\frac{m}{1-x^{2}}\,y+\frac{m(m+2)x^{2}}{(1-x^{2})^{2}}\,y\Bigr].

Step 3: substitute into (D). Write σ=1−x2\sigma=1-x^{2} for brevity. Putting the two displays into σv′′−2x(m+1)v′+(ℓ−m)(ℓ+m+1)v=0\sigma v''-2x(m+1)v'+(\ell-m)(\ell+m+1)v=0 and cancelling the common factor σ−m/2\sigma^{-m/2}, which is nonzero for ∣x∣<1\lvert x\rvert<1,

σy′′+[2mx−2x(m+1)]⏟y′y′+[m+m(m+2)x2σ−2m(m+1)x2σ+(ℓ−m)(ℓ+m+1)]⏟yy=0.\sigma y'' +\underbrace{\Bigl[2mx-2x(m+1)\Bigr]}_{y'}y' +\underbrace{\Bigl[m+\frac{m(m+2)x^{2}}{\sigma}-\frac{2m(m+1)x^{2}}{\sigma}+(\ell-m)(\ell+m+1)\Bigr]}_{y}y=0 .

The coefficient of y′y' collapses to −2x-2x. In the coefficient of yy, the two fractions combine with m(m+2)−2m(m+1)=−m2m(m+2)-2m(m+1)=-m^{2} into −m2x2/σ-m^{2}x^{2}/\sigma, and since m2x2σ=m2σ−m2\dfrac{m^{2}x^{2}}{\sigma}=\dfrac{m^{2}}{\sigma}-m^{2},

m−m2x2σ=m+m2−m2σ=m(m+1)−m2σ.m-\frac{m^{2}x^{2}}{\sigma}=m+m^{2}-\frac{m^{2}}{\sigma}=m(m+1)-\frac{m^{2}}{\sigma} .

Adding the last piece and using (ℓ−m)(ℓ+m+1)=ℓ(ℓ+1)−m(m+1)(\ell-m)(\ell+m+1)=\ell(\ell+1)-m(m+1), the coefficient of yy is ℓ(ℓ+1)−m2/σ\ell(\ell+1)-m^{2}/\sigma. Altogether

(1−x2)y′′−2xy′+[ℓ(ℓ+1)−m21−x2]y=0,(1-x^{2})y''-2xy'+\Bigl[\ell(\ell+1)-\frac{m^{2}}{1-x^{2}}\Bigr]y=0 ,

which is (AL).

Step 4: specialise and check regularity. Take u=Plu=P_{l}, ℓ=l\ell=l. Then y=(1−x2)m/2dmPl/dxmy=(1-x^{2})^{m/2}\mathrm{d}^{m}P_{l}/\mathrm{d}x^{m}, and multiplying by the constant (−1)m(-1)^{m} — which (AL) permits, being linear — gives PlmP_{l}^{m}. For part 2, dmPl/dxm\mathrm{d}^{m}P_{l}/\mathrm{d}x^{m} is a polynomial of degree l−ml-m (zero if m>lm>l), so Plm=(−1)m(1−x2)m/2PP_{l}^{m}=(-1)^{m}(1-x^{2})^{m/2}\mathcal{P} with deg⁡P=l−m\deg\mathcal{P}=l-m. The factor (1−x2)m/2(1-x^{2})^{m/2} is continuous on [−1,1][-1,1] and vanishes at ±1\pm1 for m≥1m\ge1, so PlmP_{l}^{m} is regular on the closed interval and vanishes at the ends unless m=0m=0. For part 5, each differentiation flips parity and (1−x2)m/2(1-x^{2})^{m/2} is even, so PlmP_{l}^{m} has the parity of l+ml+m. □\square

Proof(Parts 4, 6 and 7: negative mm, and the two orthogonality relations)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.2.1, the proof of eq. (18.33), pp. 588–589, and §18.2.2, the derivation of eqs. (18.36)–(18.37) and (18.39), pp. 590–591; Arfken, Weber & Harris (2013), §15.4, eqs. (15.97)–(15.107), pp. 746–748

Part 4, negative mm. Take m≥0m\ge0 and work on (−1,1)(-1,1), where 1−x>01-x>0 and 1+x>01+x>0 so that every fractional power below is unambiguous. Split (x2−1)l=(x−1)l(x+1)l(x^{2}-1)^{l}=(x-1)^{l}(x+1)^{l} and apply Leibniz to the (l+m)(l+m)-fold derivative in the definition. The term with rr derivatives on (x−1)l(x-1)^{l} is

(l+mr)l!(l−r)!(x−1)l−r⋅l!(r−m)!(x+1)r−m,\binom{l+m}{r}\frac{l!}{(l-r)!}(x-1)^{l-r}\cdot\frac{l!}{(r-m)!}(x+1)^{r-m},

nonzero only when r≤lr\le l (first factor) and l+m−r≤ll+m-r\le l, that is r≥mr\ge m (second). Multiply by (1−x2)m/2=(1−x)m/2(1+x)m/2(1-x^{2})^{m/2}=(1-x)^{m/2}(1+x)^{m/2} and write (x−1)l−r=(−1)l−r(1−x)l−r(x-1)^{l-r}=(-1)^{l-r}(1-x)^{l-r}:

Plm=(−1)m2ll!∑r=ml(l+m)! (l!)2 (−1)l−rr! (l+m−r)! (l−r)! (r−m)!  (1−x) l−r+m/2 (1+x) r−m/2.(†)P_{l}^{m}=\frac{(-1)^{m}}{2^{l}l!}\sum_{r=m}^{l} \frac{(l+m)!\,(l!)^{2}\,(-1)^{l-r}}{r!\,(l+m-r)!\,(l-r)!\,(r-m)!}\; (1-x)^{\,l-r+m/2}\,(1+x)^{\,r-m/2} . \tag{$\dagger$}

Now repeat with −m-m in place of mm, so that the definition reads Pl−m=(−1)−m2ll!(1−x2)−m/2dl−m(x2−1)l/dxl−mP_{l}^{-m}=\frac{(-1)^{-m}}{2^{l}l!}(1-x^{2})^{-m/2}\mathrm{d}^{l-m}(x^{2}-1)^{l}/\mathrm{d}x^{l-m}. Leibniz now gives terms indexed by ss, with ss derivatives on (x−1)l(x-1)^{l}:

(l−ms)l!(l−s)!(x−1)l−s⋅l!(s+m)!(x+1)s+m,0≤s≤l−m.\binom{l-m}{s}\frac{l!}{(l-s)!}(x-1)^{l-s}\cdot\frac{l!}{(s+m)!}(x+1)^{s+m}, \qquad 0\le s\le l-m .

Multiplying by (1−x)−m/2(1+x)−m/2(1-x)^{-m/2}(1+x)^{-m/2} and again replacing (x−1)l−s(x-1)^{l-s} by (−1)l−s(1−x)l−s(-1)^{l-s}(1-x)^{l-s}, the term carries (1−x) l−s−m/2(1+x) s+m/2(1-x)^{\,l-s-m/2}(1+x)^{\,s+m/2}. Substitute s=r−ms=r-m, so that rr again runs from mm to ll: the powers become

(1−x) l−r+m−m/2(1+x) r−m+m/2=(1−x) l−r+m/2(1+x) r−m/2,(1-x)^{\,l-r+m-m/2}(1+x)^{\,r-m+m/2}=(1-x)^{\,l-r+m/2}(1+x)^{\,r-m/2},

identical to those in (†)(\dagger), and the sign becomes (−1)l−s=(−1)l−r+m(-1)^{l-s}=(-1)^{l-r+m}. The factorials become

(l−m)!(r−m)! (l−r)!⋅(l!)2(l−r+m)! r!,\frac{(l-m)!}{(r-m)!\,(l-r)!}\cdot\frac{(l!)^{2}}{(l-r+m)!\,r!},

which is the same set {r!, (l+m−r)!, (l−r)!, (r−m)!}\{r!,\ (l+m-r)!,\ (l-r)!,\ (r-m)!\} as in (†)(\dagger) with (l+m)!(l+m)! replaced by (l−m)!(l-m)!. Collecting the prefactors, (−1)−m(−1)l−r+m=(−1)l−r(-1)^{-m}(-1)^{l-r+m}=(-1)^{l-r}, whereas (†)(\dagger) carries (−1)m(−1)l−r(-1)^{m}(-1)^{l-r}. Therefore every term of Pl−mP_{l}^{-m} equals the corresponding term of PlmP_{l}^{m} times the single constant

(l−m)!(l+m)!⋅1(−1)m=(−1)m(l−m)!(l+m)!,\frac{(l-m)!}{(l+m)!}\cdot\frac{1}{(-1)^{m}}=(-1)^{m}\frac{(l-m)!}{(l+m)!} ,

and summing,

Pl−m=(−1)m(l−m)!(l+m)!Plm.P_{l}^{-m}=(-1)^{m}\frac{(l-m)!}{(l+m)!}P_{l}^{m} .

Hassani obtains the same relation from the ladder-operator construction, by building Yl,−mY_{l,-m} from the bottom state with L+L_{+} instead of from the top state with L−L_{-} and comparing.

Part 6, orthogonality in ll. Fix mm and let l≠kl\neq k. Both PlmP_{l}^{m} and PkmP_{k}^{m} satisfy the Sturm-Liouville form of (AL) with the same p=1−x2p=1-x^{2} and the same q=−m2/(1−x2)q=-m^{2}/(1-x^{2}), and different eigenvalues l(l+1)l(l+1), k(k+1)k(k+1). Multiply the equation for PlmP_{l}^{m} by PkmP_{k}^{m}, integrate over [−1,1][-1,1], and integrate the pp-term by parts. The boundary contribution is

[(1−x2) Pkm dPlmdx]−11=0,\Bigl[(1-x^{2})\,P_{k}^{m}\,\frac{\mathrm{d}P_{l}^{m}}{\mathrm{d}x}\Bigr]_{-1}^{1}=0 ,

because 1−x21-x^{2} vanishes at both ends while PkmP_{k}^{m} and dPlm/dx\mathrm{d}P_{l}^{m}/\mathrm{d}x stay bounded there — for m≥1m\ge1 the derivative of (1−x2)m/2(1-x^{2})^{m/2} can blow up like (1−x2)m/2−1(1-x^{2})^{m/2-1}, but 1−x21-x^{2} times it still tends to 00, and PkmP_{k}^{m} itself vanishes at the ends. Subtracting the same identity with ll and kk exchanged, the symmetric integral ∫(1−x2)(Pkm)′(Plm)′\int(1-x^{2})(P_{k}^{m})'(P_{l}^{m})' and the qq-integral ∫qPlmPkm\int q P_{l}^{m}P_{k}^{m} both cancel, leaving

[l(l+1)−k(k+1)]∫−11PlmPkm dx=0,\bigl[l(l+1)-k(k+1)\bigr]\int_{-1}^{1}P_{l}^{m}P_{k}^{m}\,\mathrm{d}x=0 ,

hence the integral vanishes. Note that mm must be the same in both factors for the qq-terms to cancel.

The normalisation constant. Take l=kl=k and write R=x2−1R=x^{2}-1. By the definition,

Ilm=∫−11(Plm)2dx=122l(l!)2∫−11(1−x2)m[dl+mRldxl+m]2dx,I_{l}^{m}=\int_{-1}^{1}\bigl(P_{l}^{m}\bigr)^{2}\mathrm{d}x =\frac{1}{2^{2l}(l!)^{2}}\int_{-1}^{1}(1-x^{2})^{m}\Bigl[\frac{\mathrm{d}^{l+m}R^{l}}{\mathrm{d}x^{l+m}}\Bigr]^{2}\mathrm{d}x ,

the two factors (−1)m(-1)^{m} having squared away. Integrate by parts l+ml+m times, differentiating (1−x2)mdl+mRl/dxl+m(1-x^{2})^{m}\mathrm{d}^{l+m}R^{l}/\mathrm{d}x^{l+m} and integrating the other factor. The boundary term produced by the (i+1)(i+1)st of these integrations, i=0,…,l+m−1i=0,\dots,l+m-1, is

[didxi((1−x2)mdl+mRldxl+m)⋅d l+m−1−iRldx l+m−1−i]−11,\Bigl[\frac{\mathrm{d}^{i}}{\mathrm{d}x^{i}}\Bigl((1-x^{2})^{m}\frac{\mathrm{d}^{l+m}R^{l}}{\mathrm{d}x^{l+m}}\Bigr)\cdot\frac{\mathrm{d}^{\,l+m-1-i}R^{l}}{\mathrm{d}x^{\,l+m-1-i}}\Bigr]_{-1}^{1},

and it vanishes for either of two reasons according to the size of ii. For i≤m−1i\le m-1: expanding the first factor by Leibniz, every term differentiates (1−x2)m=(−R)m(1-x^{2})^{m}=(-R)^{m} at most i≤m−1i\le m-1 times and so still carries a positive power of 1−x21-x^{2}. For i≥mi\ge m: the second factor differentiates Rl=(x−1)l(x+1)lR^{l}=(x-1)^{l}(x+1)^{l} only l+m−1−i≤l−1l+m-1-i\le l-1 times, which leaves at least one undifferentiated factor x−1x-1 and one x+1x+1. Either way the product is 00 at x=±1x=\pm1. Thus

Ilm=(−1)l+m22l(l!)2∫−11Rl dl+mdxl+m[(1−x2)mdl+mRldxl+m]dx.I_{l}^{m}=\frac{(-1)^{l+m}}{2^{2l}(l!)^{2}}\int_{-1}^{1}R^{l}\,\frac{\mathrm{d}^{l+m}}{\mathrm{d}x^{l+m}}\Bigl[(1-x^{2})^{m}\frac{\mathrm{d}^{l+m}R^{l}}{\mathrm{d}x^{l+m}}\Bigr]\mathrm{d}x .

Expand the inner (l+m)(l+m)-fold derivative by Leibniz. A term differentiates (1−x2)m(1-x^{2})^{m} rr times and dl+mRl/dxl+m\mathrm{d}^{l+m}R^{l}/\mathrm{d}x^{l+m} a further l+m−rl+m-r times. The first is nonzero only for r≤2mr\le2m; the second needs 2l+2m−r≤2l2l+2m-r\le 2l, that is r≥2mr\ge2m. So r=2mr=2m is the only surviving term, and it contributes

(l+m2m)d2m(1−x2)mdx2m d2lRldx2l=(l+m)!(2m)! (l−m)!⋅(−1)m(2m)!⋅(2l)!=(−1)m(l+m)!(l−m)!(2l)!,\binom{l+m}{2m}\frac{\mathrm{d}^{2m}(1-x^{2})^{m}}{\mathrm{d}x^{2m}}\,\frac{\mathrm{d}^{2l}R^{l}}{\mathrm{d}x^{2l}} =\frac{(l+m)!}{(2m)!\,(l-m)!}\cdot(-1)^{m}(2m)!\cdot(2l)! =(-1)^{m}\frac{(l+m)!}{(l-m)!}(2l)! ,

where d2m(1−x2)m/dx2m=(−1)m(2m)!\mathrm{d}^{2m}(1-x^{2})^{m}/\mathrm{d}x^{2m}=(-1)^{m}(2m)! and d2lRl/dx2l=(2l)!\mathrm{d}^{2l}R^{l}/\mathrm{d}x^{2l}=(2l)!. Hence, with (−1)l+m(−1)m=(−1)l(-1)^{l+m}(-1)^{m}=(-1)^{l},

Ilm=(−1)l(2l)!22l(l!)2(l+m)!(l−m)!∫−11Rl dx=(2l)!22l(l!)2(l+m)!(l−m)!∫−11(1−x2)l dx.I_{l}^{m}=\frac{(-1)^{l}(2l)!}{2^{2l}(l!)^{2}}\frac{(l+m)!}{(l-m)!}\int_{-1}^{1}R^{l}\,\mathrm{d}x =\frac{(2l)!}{2^{2l}(l!)^{2}}\frac{(l+m)!}{(l-m)!}\int_{-1}^{1}(1-x^{2})^{l}\,\mathrm{d}x .

The last integral is the Kl=22l+1(l!)2/(2l+1)!K_{l}=2^{2l+1}(l!)^{2}/(2l+1)! of Legendre polynomials, so

Ilm=(2l)!22l(l!)2⋅22l+1(l!)2(2l+1)!⋅(l+m)!(l−m)!=22l+1(l+m)!(l−m)!.I_{l}^{m}=\frac{(2l)!}{2^{2l}(l!)^{2}}\cdot\frac{2^{2l+1}(l!)^{2}}{(2l+1)!}\cdot\frac{(l+m)!}{(l-m)!} =\frac{2}{2l+1}\frac{(l+m)!}{(l-m)!} .

At m=0m=0 this is the Legendre value 2/(2l+1)2/(2l+1), as it must be. The θ\theta-form follows from x=cos⁡θx=\cos\theta, dx=−sin⁡θ dθ\mathrm{d}x=-\sin\theta\,\mathrm{d}\theta.

Totality is not part of this argument. What the Sturm-Liouville computation gives is orthogonality and a norm, and no amount of either implies that the PlmP_{l}^{m} span. That step is done separately below.

Part 7, orthogonality in mm. Read (AL) the other way: for fixed ll, the family {Plm}m\{P_{l}^{m}\}_{m} consists of eigenfunctions of

ddx[(1−x2)dydx]+l(l+1)y=+m2 11−x2 y,\frac{\mathrm{d}}{\mathrm{d}x}\Bigl[(1-x^{2})\frac{\mathrm{d}y}{\mathrm{d}x}\Bigr]+l(l+1)y=+m^{2}\,\frac{1}{1-x^{2}}\,y ,

a Sturm-Liouville problem with p=1−x2p=1-x^{2}, q=l(l+1)q=l(l+1), eigenvalue −m2-m^{2} and weight ρ=(1−x2)−1\rho=(1-x^{2})^{-1}; the natural interval is again [−1,1][-1,1] and the PlmP_{l}^{m} are regular there. Distinct eigenvalues force orthogonality with respect to that weight. Since the eigenvalue is −m2-m^{2}, "distinct" means ∣m∣≠∣k∣\lvert m\rvert\neq\lvert k\rvert, and that is the hypothesis under which Riley, Hobson and Bence state the vanishing, their eq. (18.39); m=−km=-k gives the same eigenvalue and the argument says nothing about it (part 4 does, and the integral is then nonzero).

The restriction to m,k≥1m,k\ge1 in part 7 belongs to the constant, not to the vanishing. The value (l+m)!/[m(l−m)!](l+m)!/[m(l-m)!], which Arfken, Weber and Harris and Riley, Hobson and Bence both record, blows up as m→0m\to0, and rightly so: at m=0m=0 the integral itself diverges, because Pl0(±1)=(±1)l≠0P_{l}^{0}(\pm1)=(\pm1)^{l}\neq0 leaves the integrand Pl2/(1−x2)P_{l}^{2}/(1-x^{2}) behaving like [2(1−x)]−1[2(1-x)]^{-1} at the upper end. For m≥1m\ge1 the factor (1−x2)m(1-x^{2})^{m} in (Plm)2(P_{l}^{m})^{2} more than cancels the weight and the integral converges. Both books add that the relation is not used in practice. □\square

Lemma(Part 6: totality of {Plm}l≥m\{P_{l}^{m}\}_{l\ge m} at fixed mm)Original

Fix an integer m≥0m\ge0 and suppose f∈L2(−1,1)f\in L^{2}(-1,1) satisfies ∫−11fPlm dx=0\int_{-1}^{1}fP_{l}^{m}\,\mathrm{d}x=0 for every l≥ml\ge m. We show f=0f=0 almost everywhere; by Orthonormal bases in Hilbert space that is exactly the totality of the orthonormal system of part 6, and the expansion and the coefficient formula follow from it.

Step 1: strip the fractional factor. By part 2, Plm=(1−x2)m/2PlmP_{l}^{m}=(1-x^{2})^{m/2}\mathcal{P}_{l}^{m} with deg⁡Plm=l−m\deg\mathcal{P}_{l}^{m}=l-m, and the leading coefficient is nonzero because dmPl/dxm\mathrm{d}^{m}P_{l}/\mathrm{d}x^{m} has degree exactly l−ml-m. As ll runs over m,m+1,m+2,…m,m+1,m+2,\dots the degrees l−ml-m run over 0,1,2,…0,1,2,\dots, one of each, so span⁡{Plm:l≥m}\operatorname{span}\{\mathcal{P}_{l}^{m}:l\ge m\} is the whole space of polynomials. Put g=(1−x2)m/2fg=(1-x^{2})^{m/2}f. The hypothesis says ∫−11g Plm dx=0\int_{-1}^{1}g\,\mathcal{P}_{l}^{m}\,\mathrm{d}x=0 for every l≥ml\ge m, hence ∫−11g q dx=0\int_{-1}^{1}g\,q\,\mathrm{d}x=0 for every polynomial qq.

Step 2: use the Legendre case. The factor (1−x2)m/2(1-x^{2})^{m/2} is bounded by 11 on (−1,1)(-1,1), so g∈L2(−1,1)g\in L^{2}(-1,1) with ∥g∥≤∥f∥\lVert g\rVert\le\lVert f\rVert. Taking q=Pkq=P_{k} for each k≥0k\ge0, gg is orthogonal to every Legendre polynomial, and that system is total in L2[−1,1]L^{2}[-1,1] (Legendre polynomials, part 4, where it is proved from the Weierstrass approximation theorem). Therefore g=0g=0 almost everywhere.

Step 3: return to ff. (1−x2)m/2>0(1-x^{2})^{m/2}>0 at every point of the open interval (−1,1)(-1,1), so g=0g=0 a.e. forces f=0f=0 a.e. there. □\square

Remark(Where the hypotheses are used)Original

∣m∣≤l\lvert m\rvert\le l and ll an integer. These are not conventions but the only way to get a solution bounded on [−1,1][-1,1], by part 3. Arfken, Weber and Harris note in a footnote the natural hope that the prefactor (1−x2)m/2(1-x^{2})^{m/2}, which vanishes at the ends, might tame a divergent series P\mathcal{P}; they state that it does not. For ∣m∣>l\lvert m\rvert>l the function PlmP_{l}^{m} is not a small solution but identically zero, which is why the expansion in part 6 starts at l=ml=m.

The same mm in both factors of part 6. The qq-term −m2/(1−x2)-m^{2}/(1-x^{2}) cancels on subtraction only if it is identical in the two equations. With different mm's there is a leftover (k2−m2)∫PlmPl′k(1−x2)−1(k^{2}-m^{2})\int P_{l}^{m}P_{l'}^{k}(1-x^{2})^{-1} and the relation fails; what survives instead is part 7, with its different weight. Riley, Hobson and Bence are explicit that mm must match.

Vanishing of p=1−x2p=1-x^{2} at both ends. This is what kills the boundary terms in part 6 and in the normalisation integral, and it is why no boundary condition has to be imposed by hand. The step is delicate for m≥1m\ge1: dPlm/dx\mathrm{d}P_{l}^{m}/\mathrm{d}x is unbounded at x=±1x=\pm1 when m=1m=1 (it behaves like (1−x2)−1/2(1-x^{2})^{-1/2}), and what vanishes is the product (1−x2)Pkm(Plm)′(1-x^{2})P_{k}^{m}(P_{l}^{m})', in which PkmP_{k}^{m} supplies another factor (1−x2)m/2(1-x^{2})^{m/2}. On a spherical cap θ∈[0,θ0]\theta\in[0,\theta_{0}] with θ0<π\theta_{0}<\pi the boundary term at θ0\theta_{0} does not vanish and the PlmP_{l}^{m} are not orthogonal; the correct eigenfunctions there solve (AL) with non-integer ℓ\ell. Strauss sets the diffusion equation in the solid cone {r<a, θ<α}\{r<a,\ \theta<\alpha\} as Exercise 8 of §10.6 ((Strauss (2008), §10.6, Exercise 8, pp. 293–294)); his hint asks the reader to separate variables, check that the polar factor p(cos⁡θ)p(\cos\theta) satisfies the associated Legendre equation, expand pp in powers of s−1s-1, and write down the equations that determine the eigenvalues. It stops there and does not say what those eigenvalues are; the condition p(cos⁡α)=0p(\cos\alpha)=0 is what pushes ℓ\ell off the integers for a general half-angle α\alpha.

The Condon-Shortley phase. It is used nowhere in any proof: it is a constant and (AL) is linear and homogeneous. It cancels identically out of parts 6 and 7, which are quadratic. It matters only in parts 4, 5 and 8, and in any table. Compare Arfken, Weber and Harris's P11=−sin⁡θP_{1}^{1}=-\sin\theta with Riley, Hobson and Bence's P11=+sin⁡θP_{1}^{1}=+\sin\theta: both books are internally consistent, and both produce the same Y11Y_{1}^{1}.

Reality. All of PlmP_{l}^{m}, (AL) and both weights are real, so nothing above needs complex conjugation. Conjugation appears only once the azimuthal factor eimφ\mathrm{e}^{\mathrm{i}m\varphi} is attached, in Spherical harmonics.

Method(Getting PlmP_{l}^{m} and its integrals)AdaptedAdapted from Arfken, Weber & Harris (2013), §15.4, Example 15.4.1 and eqs. (15.88), (15.93)–(15.94), pp. 745–746
  1. For a single value, small ll and mm: differentiate PlP_{l} mm times, multiply by (1−x2)m/2(1-x^{2})^{m/2}, and attach (−1)m(-1)^{m}.
  2. For a whole table, or for large ll: climb in ll at fixed mm, as Arfken, Weber and Harris do. Start from the closed forms Pmm=(−1)m(2m−1)!! (1−x2)m/2,Pm+1m=(−1)m(2m+1)!! x (1−x2)m/2,P_{m}^{m}=(-1)^{m}(2m-1)!!\,(1-x^{2})^{m/2}, \qquad P_{m+1}^{m}=(-1)^{m}(2m+1)!!\,x\,(1-x^{2})^{m/2}, the second being the l=ml=m case of the recurrence with the Pm−1m=0P_{m-1}^{m}=0 term dropped, and then iterate Pl+1m=(2l+1)x Plm−(l+m)Pl−1ml−m+1.P_{l+1}^{m}=\frac{(2l+1)x\,P_{l}^{m}-(l+m)P_{l-1}^{m}}{l-m+1} . Remember that any Pl′mP_{l'}^{m} with l′<ml'<m is zero, so terms containing one are simply omitted.
  3. To raise mm instead: use the second recurrence of part 8, solved for Plm+1P_{l}^{m+1}.
  4. For negative mm: do not differentiate a negative number of times; use part 4.
  5. For any integral ∫−11PlmPkm dx\int_{-1}^{1}P_{l}^{m}P_{k}^{m}\,\mathrm{d}x: do not integrate. It is 00 unless l=kl=k, and then it is 22l+1(l+m)!(l−m)!\dfrac{2}{2l+1}\dfrac{(l+m)!}{(l-m)!}.
  6. To normalise for quantum mechanics: the combination that is a unit vector in L2(0,π;sin⁡θ dθ)L^{2}(0,\pi;\sin\theta\,\mathrm{d}\theta) is Θlm(θ)=2l+12(l−m)!(l+m)!  Plm(cos⁡θ),\Theta_{l}^{m}(\theta)=\sqrt{\frac{2l+1}{2}\frac{(l-m)!}{(l+m)!}}\;P_{l}^{m}(\cos\theta), which is the θ\theta half of a spherical harmonic.
Example(P32P_{3}^{2} from three routes, and its norm)Original

Step 1: from the definition. P3=12(5x3−3x)P_{3}=\tfrac12(5x^{3}-3x), so P3′=12(15x2−3)P_{3}'=\tfrac12(15x^{2}-3) and P3′′=15xP_{3}''=15x. With m=2m=2 the phase is (−1)2=+1(-1)^{2}=+1 and

P32(x)=(1−x2)1⋅15x=15x (1−x2).P_{3}^{2}(x)=(1-x^{2})^{1}\cdot 15x=15x\,(1-x^{2}) .

Step 2: from the Rodrigues form. Here l+m=5l+m=5 and (x2−1)3=x6−3x4+3x2−1(x^{2}-1)^{3}=x^{6}-3x^{4}+3x^{2}-1. Only the leading term survives five differentiations: d5x6/dx5=6! x/1!=720x\mathrm{d}^{5}x^{6}/\mathrm{d}x^{5}=6!\,x/1!=720x, while x4x^{4}, x2x^{2} and the constant are annihilated. Then

P32=(+1)23⋅3!(1−x2)⋅720x=72048x(1−x2)=15x(1−x2),P_{3}^{2}=\frac{(+1)}{2^{3}\cdot3!}(1-x^{2})\cdot 720x=\frac{720}{48}x(1-x^{2})=15x(1-x^{2}),

the same thing.

Step 3: from the recurrence. Start at the top: P22=(−1)2(3)!! (1−x2)=3(1−x2)P_{2}^{2}=(-1)^{2}(3)!!\,(1-x^{2})=3(1-x^{2}), since (2⋅2−1)!!=3!!=3(2\cdot2-1)!!=3!!=3. Then the l=m=2l=m=2 case of the step-in-ll recurrence, with P12=0P_{1}^{2}=0 dropped, gives

P32=(2⋅2+1)x P22−02−2+1=5x⋅3(1−x2)=15x(1−x2).P_{3}^{2}=\frac{(2\cdot2+1)x\,P_{2}^{2}-0}{2-2+1}=5x\cdot3(1-x^{2})=15x(1-x^{2}).

Three routes, one answer.

Step 4: the normalisation integral, by hand. Part 6 predicts

∫−11(P32)2dx=22⋅3+1⋅(3+2)!(3−2)!=27⋅1201=2407.\int_{-1}^{1}\bigl(P_{3}^{2}\bigr)^{2}\mathrm{d}x=\frac{2}{2\cdot3+1}\cdot\frac{(3+2)!}{(3-2)!}=\frac{2}{7}\cdot\frac{120}{1}=\frac{240}{7} .

Check it directly:

∫−11225x2(1−x2)2dx=225∫−11(x2−2x4+x6)dx=225(23−45+27).\int_{-1}^{1}225x^{2}(1-x^{2})^{2}\mathrm{d}x =225\int_{-1}^{1}\bigl(x^{2}-2x^{4}+x^{6}\bigr)\mathrm{d}x =225\Bigl(\frac23-\frac45+\frac27\Bigr).

Over 105105: 23=70105\tfrac23=\tfrac{70}{105}, 45=84105\tfrac45=\tfrac{84}{105}, 27=30105\tfrac27=\tfrac{30}{105}, so the bracket is 70−84+30105=16105\tfrac{70-84+30}{105}=\tfrac{16}{105} and the integral is 225⋅16105=3600105=2407=34.2857…225\cdot\tfrac{16}{105}=\tfrac{3600}{105}=\tfrac{240}{7}=34.2857\ldots — exactly the predicted value.

Step 5: two more checks. The m=1m=1 case: P31=−32(5x2−1)(1−x2)1/2P_{3}^{1}=-\tfrac32(5x^{2}-1)(1-x^{2})^{1/2}, and

∫−11(P31)2dx=94∫−11(5x2−1)2(1−x2) dx=247,\int_{-1}^{1}\bigl(P_{3}^{1}\bigr)^{2}\mathrm{d}x=\frac94\int_{-1}^{1}(5x^{2}-1)^{2}(1-x^{2})\,\mathrm{d}x=\frac{24}{7},

against the formula 27⋅4!2!=27⋅12=247\tfrac27\cdot\tfrac{4!}{2!}=\tfrac27\cdot12=\tfrac{24}{7}. And the second orthogonality of part 7 at l=3l=3, m=1m=1:

∫−11(P31)21−x2 dx=94∫−11(5x2−1)2 dx=94∫−11(25x4−10x2+1)dx=94(25⋅25−10⋅23+2)=94⋅163=12,\int_{-1}^{1}\frac{\bigl(P_{3}^{1}\bigr)^{2}}{1-x^{2}}\,\mathrm{d}x =\frac94\int_{-1}^{1}(5x^{2}-1)^{2}\,\mathrm{d}x =\frac94\int_{-1}^{1}\bigl(25x^{4}-10x^{2}+1\bigr)\mathrm{d}x =\frac94\Bigl(25\cdot\tfrac25-10\cdot\tfrac23+2\Bigr)=\frac94\cdot\frac{16}{3}=12,

since 10−203+2=16310-\tfrac{20}{3}+2=\tfrac{16}{3}; and this matches (l+m)!m (l−m)!=4!1⋅2!=12\dfrac{(l+m)!}{m\,(l-m)!}=\dfrac{4!}{1\cdot2!}=12. Also ∫−11P31P32(1−x2)−1dx=0\int_{-1}^{1}P_{3}^{1}P_{3}^{2}(1-x^{2})^{-1}\mathrm{d}x=0 by parity, since P31P_{3}^{1} is even and P32P_{3}^{2} is odd — consistent with part 5, (−1)3+1=+1(-1)^{3+1}=+1 and (−1)3+2=−1(-1)^{3+2}=-1. Every integral in this example was checked symbolically.

ApplicationOriginal

Anything separated in spherical coordinates. Whenever ∇2\nabla^{2} is separated in spherical polars without assuming axial symmetry, the polar-angle factor is forced to be Plm(cos⁡θ)P_{l}^{m}(\cos\theta) and the azimuthal factor eimφ\mathrm{e}^{\mathrm{i}m\varphi} (or cos⁡mφ\cos m\varphi, sin⁡mφ\sin m\varphi). The integer mm comes from single-valuedness in φ\varphi; the integer l≥∣m∣l\ge\lvert m\rvert comes from part 3; and the pair (l,m)(l,m) then labels the solution completely. This applies to the Laplace, Helmholtz, Schrödinger and wave equations alike, and it is where Separation of variables in standard coordinates and Spherical harmonics begin.

Atomic structure. In the hydrogen-like problem, −ℏ2∇2/2μ+V(r)-\hbar^{2}\nabla^{2}/2\mu+V(r) commutes with the angular-momentum operators, and Plm(cos⁡θ)eimφP_{l}^{m}(\cos\theta)\mathrm{e}^{\mathrm{i}m\varphi} is the simultaneous eigenfunction: ll is the orbital quantum number, mm the magnetic one, and the (2l+1)(2l+1)-fold degeneracy in mm is the statement that ll fixes the radial equation while mm does not. The Condon-Shortley phase is in the tables of atomic spectroscopy precisely because it makes the angular-momentum ladder operators act with uniform signs; this is why a formula lifted from a spectroscopy table into an electrodynamics calculation that follows Jackson will be wrong by (−1)m(-1)^{m}. Hassani's construction, which produces the angular eigenfunctions by lowering from the top state with L−L_{-}, shows what the ladder route does and does not settle: it fixes the relative phases along a ladder of fixed ll, and it delivers the phase-free PlmP_{l}^{m} of his eq. (13.35), with the (−1)m(-1)^{m} then put into YlmY_{lm} by hand in eq. (13.34). The phase is a choice made in the definition, in whichever of the two places the book prefers; no derivation forces it.

Geomagnetism and gravimetry. The Earth's magnetic and gravitational fields are catalogued as expansions in Plm(cos⁡θ){cos⁡mφ,sin⁡mφ}P_{l}^{m}(\cos\theta)\{\cos m\varphi,\sin m\varphi\} — the Gauss coefficients glmg_{l}^{m}, hlmh_{l}^{m} of the International Geomagnetic Reference Field, and the Stokes coefficients ClmC_{lm}, SlmS_{lm} of a gravity model. Here ll measures the spatial scale and mm the number of lobes around a line of latitude, and part 6 is what makes the coefficients recoverable from satellite data by projection. The normalisation differs between fields: geodesy mostly uses the "fully normalised" Pˉlm\bar P_{l}^{m}, which is PlmP_{l}^{m} times (2−δm0)(2l+1)(l−m)!/(l+m)!\sqrt{(2-\delta_{m0})(2l+1)(l-m)!/(l+m)!}, because the raw (l+m)!/(l−m)!(l+m)!/(l-m)! in part 6 overflows at the degree l∼2000l\sim2000 of a modern model.

Normal modes of a sphere. The free oscillations of the Earth, the acoustic modes of a spherical cavity, and the vibrations of a drum stretched over a hemisphere all have Plm(cos⁡θ)P_{l}^{m}(\cos\theta) as their angular shape, with the radial factor a spherical Bessel function (Spherical Bessel functions). The nodal pattern is read straight off part 5: l−∣m∣l-\lvert m\rvert nodal circles of latitude from the zeros of the polynomial factor, and ∣m∣\lvert m\rvert nodal meridians from the azimuthal factor.

Theorem

Spherical harmonics

Adapted

The orthonormal basis of L^2 of the sphere made from the associated Legendre functions and e^{i m phi}: its definition and conventions, the Laplace series that expands any square-integrable function on the sphere, and the addition theorem that turns a Legendre polynomial of the angle between two directions into a sum of 2l+1 products.

Prerequisites: Associated Legendre functions, Legendre polynomials, Orthonormal bases in Hilbert space

Definition(Spherical harmonics)AdaptedAdapted from Arfken, Weber & Harris (2013), §15.5, eqs. (15.132), (15.135) and (15.137), pp. 756–757; Hassani (2013), §13.4, eqs. (13.25)–(13.33), pp. 406–408

Let (θ,φ)(\theta,\varphi) be spherical polar angles on the unit sphere S2S^{2}: θ∈[0,π]\theta\in[0,\pi] the polar angle from the zz-axis, φ∈[0,2π)\varphi\in[0,2\pi) the azimuth. Write dΩ=sin⁡θ dθ dφ=dφ d(cos⁡θ)\mathrm{d}\Omega=\sin\theta\,\mathrm{d}\theta\,\mathrm{d}\varphi=\mathrm{d}\varphi\,\mathrm{d}(\cos\theta) for the element of solid angle, so that ∫S2dΩ=4π\int_{S^{2}}\mathrm{d}\Omega=4\pi, and

(f,g)=∫S2f(θ,φ) g(θ,φ)‾ dΩ(f,g)=\int_{S^{2}}f(\theta,\varphi)\,\overline{g(\theta,\varphi)}\,\mathrm{d}\Omega

for the inner product of L2(S2)L^{2}(S^{2}).

For integers l≥0l\ge0 and −l≤m≤l-l\le m\le l, the spherical harmonic is

Ylm(θ,φ)=2l+14π (l−m)!(l+m)!  Plm(cos⁡θ)  eimφ,Y_{l}^{m}(\theta,\varphi)=\sqrt{\frac{2l+1}{4\pi}\,\frac{(l-m)!}{(l+m)!}}\;P_{l}^{m}(\cos\theta)\;\mathrm{e}^{\mathrm{i}m\varphi},

with PlmP_{l}^{m} the associated Legendre function of Associated Legendre functions, including its Condon-Shortley phase (−1)m(-1)^{m}. There are 2l+12l+1 harmonics of each order ll. This is Arfken, Weber and Harris's eq. (15.137) verbatim; books whose PlmP_{l}^{m} is phase-free display an extra (−1)m(-1)^{m} in front and arrive at the same function, as the remark below sets out. Equivalently Ylm=Θlm(θ) Φm(φ)Y_{l}^{m}=\Theta_{l}^{m}(\theta)\,\Phi_{m}(\varphi) with the two separately normalised factors of Arfken, Weber and Harris,

Θlm(θ)=2l+12(l−m)!(l+m)! Plm(cos⁡θ),Φm(φ)=eimφ2π.\Theta_{l}^{m}(\theta)=\sqrt{\frac{2l+1}{2}\frac{(l-m)!}{(l+m)!}}\,P_{l}^{m}(\cos\theta), \qquad \Phi_{m}(\varphi)=\frac{\mathrm{e}^{\mathrm{i}m\varphi}}{\sqrt{2\pi}} .

The first few, with x=cos⁡θx=\cos\theta:

Y00=14π,Y10=34πcos⁡θ,Y1±1=∓38πsin⁡θ e±iφ,Y_{0}^{0}=\frac{1}{\sqrt{4\pi}}, \qquad Y_{1}^{0}=\sqrt{\frac{3}{4\pi}}\cos\theta, \qquad Y_{1}^{\pm1}=\mp\sqrt{\frac{3}{8\pi}}\sin\theta\,\mathrm{e}^{\pm\mathrm{i}\varphi},Y20=516π(3cos⁡2θ−1),Y2±1=∓158πsin⁡θcos⁡θ e±iφ,Y2±2=1532πsin⁡2θ e±2iφ.Y_{2}^{0}=\sqrt{\frac{5}{16\pi}}\bigl(3\cos^{2}\theta-1\bigr), \quad Y_{2}^{\pm1}=\mp\sqrt{\frac{15}{8\pi}}\sin\theta\cos\theta\,\mathrm{e}^{\pm\mathrm{i}\varphi}, \quad Y_{2}^{\pm2}=\sqrt{\frac{15}{32\pi}}\sin^{2}\theta\,\mathrm{e}^{\pm2\mathrm{i}\varphi}.

They are the eigenfunctions of the angular part of the Laplacian:

[1sin⁡θ∂∂θ(sin⁡θ∂∂θ)+1sin⁡2θ∂2∂φ2]Ylm=−l(l+1) Ylm,\Bigl[\frac{1}{\sin\theta}\frac{\partial}{\partial\theta}\Bigl(\sin\theta\frac{\partial}{\partial\theta}\Bigr)+\frac{1}{\sin^{2}\theta}\frac{\partial^{2}}{\partial\varphi^{2}}\Bigr]Y_{l}^{m}=-l(l+1)\,Y_{l}^{m} ,

which is why they appear in every separation of ∇2\nabla^{2} in spherical polars (Separation of variables in standard coordinates). Hassani reaches the same functions as the simultaneous eigenfunctions of the quantum angular-momentum operators, L2Ylm=l(l+1)YlmL^{2}Y_{l}^{m}=l(l+1)Y_{l}^{m} and LzYlm=mYlmL_{z}Y_{l}^{m}=mY_{l}^{m}, built by applying the lowering operator L−L_{-} repeatedly to the top state Yll∝sin⁡lθ eilφY_{l}^{l}\propto\sin^{l}\theta\,\mathrm{e}^{\mathrm{i}l\varphi}.

Remark(Conventions, and how to translate between books)AdaptedAdapted from Arfken, Weber & Harris (2013), §15.4, eq. (15.78), p. 743, and §15.5, p. 757 and p. 758, and Table 15.4, p. 760; Riley, Hobson & Bence (2006), §18.2.1, eq. (18.32), p. 588, and §18.3, eq. (18.45), p. 593; Strauss (2008), §10.3, p. 275; Hassani (2013), §13.4, Box 13.4.2, eqs. (13.34)–(13.35), p. 409, and eq. (13.37), p. 410

The function YlmY_{l}^{m} above is the one almost universally used in physics, but it is reached by two different bookkeepings and there is a third, unnormalised, convention.

  • The phase in PP. Arfken, Weber and Harris put the (−1)m(-1)^{m} into PlmP_{l}^{m} (their eq. (15.78)) and so write Ylm=2l+14π(l−m)!(l+m)! Plm(cos⁡θ)eimφY_{l}^{m}=\sqrt{\tfrac{2l+1}{4\pi}\tfrac{(l-m)!}{(l+m)!}}\,P_{l}^{m}(\cos\theta)\mathrm{e}^{\mathrm{i}m\varphi}, eq. (15.137) — our definition.
  • The phase in YY. Riley, Hobson and Bence and Hassani keep PlmP_{l}^{m} phase-free (eqs. (18.32) and (13.35) respectively; Hassani's P11=+sin⁡θP_{1}^{1}=+\sin\theta, where ours is −sin⁡θ-\sin\theta) and display the factor in the harmonic instead: Ylm=(−1)m2l+14π(l−m)!(l+m)! Plm(cos⁡θ)eimφY_{l}^{m}=(-1)^{m}\sqrt{\tfrac{2l+1}{4\pi}\tfrac{(l-m)!}{(l+m)!}}\,P_{l}^{m}(\cos\theta)\mathrm{e}^{\mathrm{i}m\varphi}, which is Riley, Hobson and Bence's eq. (18.45) and Hassani's eq. (13.34) in Box 13.4.2.

The two bookkeepings give the same YlmY_{l}^{m}: the (−1)m(-1)^{m} sits in a different factor, not in the product. The tables agree, including Y11=−3/8πsin⁡θ eiφY_{1}^{1}=-\sqrt{3/8\pi}\sin\theta\,\mathrm{e}^{\mathrm{i}\varphi}. Arfken, Weber and Harris call the resulting signs the Condon-Shortley phase and note that its purpose is to simplify calculations in the quantum theory of angular momentum.

  • Strauss uses Ylm=Pl∣m∣(cos⁡θ)eimφY_{l}^{m}=P_{l}^{\lvert m\rvert}(\cos\theta)\mathrm{e}^{\mathrm{i}m\varphi} with no normalising factor and no phase, remarking that the constant coefficients are arbitrary anyway. His functions are orthogonal but not orthonormal, and his table lists real combinations such as sin⁡2θcos⁡2φ=(x2−y2)/r2\sin^{2}\theta\cos2\varphi=(x^{2}-y^{2})/r^{2}.

A fourth variant, the real spherical harmonics, replaces eimφ\mathrm{e}^{\mathrm{i}m\varphi} by cos⁡mφ\cos m\varphi and sin⁡mφ\sin m\varphi. Arfken, Weber and Harris use these in the geodesy application, writing Ymle=Plm(cos⁡θ)cos⁡mφY^{e}_{ml}=P_{l}^{m}(\cos\theta)\cos m\varphi and Ymlo=Plm(cos⁡θ)sin⁡mφY^{o}_{ml}=P_{l}^{m}(\cos\theta)\sin m\varphi after Morse and Feshbach — note that those authors put the mm index first, a trap when reading tables of coefficients.

Everything below is stated in our convention. What is convention-independent: the eigenvalue equation, the orthonormality relation, the completeness statement, and the addition theorem (whose right-hand side is ∣Y∣2\lvert Y\rvert^{2}-like and so phase-blind). What is convention-dependent: every entry with odd mm in a table of PlmP_{l}^{m} (not of YlmY_{l}^{m}, which the first two bookkeepings agree on), and anything at all if Strauss's unnormalised functions are used — his lack of both the phase and the factorial factor turns the conjugation relation of part 2 below into the bare Ylm‾=Yl−m\overline{Y_{l}^{m}}=Y_{l}^{-m}, with no (−1)m(-1)^{m}, and makes part 1 an orthogonality relation with a non-unit constant. Part 2 as stated is the same under the two conventions compared above, since both give the same YlmY_{l}^{m}; it is Hassani's eq. (13.37) and Riley, Hobson and Bence's relation on p. 593.

Theorem(Orthonormality, completeness, closure and the addition theorem)AdaptedAdapted from Arfken, Weber & Harris (2013), §15.5, eqs. (15.138), (15.144)–(15.145), (15.147)–(15.149), pp. 757–764; Riley, Hobson & Bence (2006), §18.3, eqs. (18.46)–(18.53), pp. 593–595; Strauss (2008), §10.3, Theorem 1, p. 275, and p. 277; Hassani (2013), §13.4, eqs. (13.31)–(13.32) and (13.37), pp. 408–410

Let l,l′≥0l,l'\ge0 and ∣m∣≤l\lvert m\rvert\le l, ∣m′∣≤l′\lvert m'\rvert\le l' be integers, and write Ω=(θ,φ)\Omega=(\theta,\varphi).

  1. Orthonormality. (Ylm,Yl′m′)=∫S2Ylm(Ω) Yl′m′(Ω)‾ dΩ=δll′ δmm′.\bigl(Y_{l}^{m},Y_{l'}^{m'}\bigr)=\int_{S^{2}}Y_{l}^{m}(\Omega)\,\overline{Y_{l'}^{m'}(\Omega)}\,\mathrm{d}\Omega=\delta_{ll'}\,\delta_{mm'} .
  2. Conjugation and parity. Ylm(Ω)‾=(−1)m Yl−m(Ω),Ylm(π−θ,φ+π)=(−1)l Ylm(θ,φ).\overline{Y_{l}^{m}(\Omega)}=(-1)^{m}\,Y_{l}^{-m}(\Omega), \qquad Y_{l}^{m}(\pi-\theta,\varphi+\pi)=(-1)^{l}\,Y_{l}^{m}(\theta,\varphi).
  3. Values on the axis. Ylm(0,φ)=2l+14π δm0,Ylm(π,φ)=(−1)l2l+14π δm0,Yl0=2l+14π Pl(cos⁡θ).Y_{l}^{m}(0,\varphi)=\sqrt{\frac{2l+1}{4\pi}}\,\delta_{m0}, \qquad Y_{l}^{m}(\pi,\varphi)=(-1)^{l}\sqrt{\frac{2l+1}{4\pi}}\,\delta_{m0}, \qquad Y_{l}^{0}=\sqrt{\frac{2l+1}{4\pi}}\,P_{l}(\cos\theta).
  4. Completeness (the Laplace series). The family {Ylm}\{Y_{l}^{m}\} is a complete orthonormal system in L2(S2)L^{2}(S^{2}): every f∈L2(S2)f\in L^{2}(S^{2}) has f(Ω)=∑l=0∞∑m=−llalm Ylm(Ω),alm=∫S2Ylm(Ω)‾ f(Ω) dΩ,f(\Omega)=\sum_{l=0}^{\infty}\sum_{m=-l}^{l}a_{lm}\,Y_{l}^{m}(\Omega), \qquad a_{lm}=\int_{S^{2}}\overline{Y_{l}^{m}(\Omega)}\,f(\Omega)\,\mathrm{d}\Omega, the series converging in L2(S2)L^{2}(S^{2}), with ∥f∥2=∑l,m∣alm∣2\lVert f\rVert^{2}=\sum_{l,m}\lvert a_{lm}\rvert^{2}. This is Strauss's Theorem 1, stated for every function on a sphere whose square is integrable; Arfken, Weber and Harris call the expansion a Laplace series and state it as a uniformly convergent double series for ff with sufficient continuity, referring to Hobson for the proof. Riley, Hobson and Bence present it as the exact analogue of a Fourier series and a special case of the general Sturm-Liouville property. We state the L2L^{2} version, which is what the orthonormality of part 1 plus the density of the trigonometric-times-polynomial system gives; pointwise and uniform statements need extra smoothness of ff and are not proved here.
  5. Closure. As an identity of distributions on S2×S2S^{2}\times S^{2}, ∑l=0∞∑m=−llYlm(Ω) Ylm(Ω′)‾=δ(Ω−Ω′),\sum_{l=0}^{\infty}\sum_{m=-l}^{l}Y_{l}^{m}(\Omega)\,\overline{Y_{l}^{m}(\Omega')}=\delta(\Omega-\Omega'), where δ(Ω−Ω′)\delta(\Omega-\Omega') is the Dirac delta on the sphere: δ(Ω−Ω′)=0\delta(\Omega-\Omega')=0 for Ω≠Ω′\Omega\neq\Omega' and ∫S2δ(Ω) dΩ=1\int_{S^{2}}\delta(\Omega)\,\mathrm{d}\Omega=1.
  6. Addition theorem. Let Ω=(θ,φ)\Omega=(\theta,\varphi) and Ω′=(θ′,φ′)\Omega'=(\theta',\varphi') be two directions separated by the angle γ\gamma, so that cos⁡γ=cos⁡θcos⁡θ′+sin⁡θsin⁡θ′cos⁡(φ−φ′).\cos\gamma=\cos\theta\cos\theta'+\sin\theta\sin\theta'\cos(\varphi-\varphi') . Then for every l≥0l\ge0, Pl(cos⁡γ)=4π2l+1∑m=−llYlm(Ω) Ylm(Ω′)‾.P_{l}(\cos\gamma)=\frac{4\pi}{2l+1}\sum_{m=-l}^{l}Y_{l}^{m}(\Omega)\,\overline{Y_{l}^{m}(\Omega')} . Equivalently, in terms of associated Legendre functions, Pl(cos⁡γ)=∑m=−ll(l−m)!(l+m)! Plm(cos⁡θ)Plm(cos⁡θ′) eim(φ−φ′).P_{l}(\cos\gamma)=\sum_{m=-l}^{l}\frac{(l-m)!}{(l+m)!}\,P_{l}^{m}(\cos\theta)P_{l}^{m}(\cos\theta')\,\mathrm{e}^{\mathrm{i}m(\varphi-\varphi')} .
  7. Solid harmonics. rlYlm(θ,φ)r^{l}Y_{l}^{m}(\theta,\varphi) is a harmonic function on R3\mathbb{R}^{3} and, in fact (Strauss), a homogeneous polynomial of degree ll in x,y,zx,y,z. The general solution of ∇2u=0\nabla^{2}u=0 regular at the origin is ∑l,mAlmrlYlm\sum_{l,m}A_{lm}r^{l}Y_{l}^{m}, and the general solution regular at infinity is ∑l,mBlmr−l−1Ylm\sum_{l,m}B_{lm}r^{-l-1}Y_{l}^{m}.
Idea of the proofOriginal

Parts 1 to 3 are bookkeeping: the sphere's measure factorises as dφ d(cos⁡θ)\mathrm{d}\varphi\,\mathrm{d}(\cos\theta), so the double integral splits into the φ\varphi-integral ∫02πei(m−m′)φdφ=2πδmm′\int_{0}^{2\pi}\mathrm{e}^{\mathrm{i}(m-m')\varphi}\mathrm{d}\varphi=2\pi\delta_{mm'} and, once m=m′m=m', the associated Legendre orthogonality already proved. The constant under the square root is exactly what makes the product come out as 11.

Part 6 is the one real theorem, and the proof is a two-step trick.

First step. The closure relation of part 5 says ∑l∑mYlm(Ω)Ylm(Ω′)‾\sum_{l}\sum_{m}Y_{l}^{m}(\Omega)\overline{Y_{l}^{m}(\Omega')} is the delta function on the sphere. But a delta function that depends only on the relative position of two directions can also be expanded in Legendre polynomials of the included angle γ\gamma, because Pl(cos⁡γ)P_{l}(\cos\gamma) is a complete system for functions of one angle. Computing those coefficients needs only Pl(1)=1P_{l}(1)=1, and gives (2l+1)/4π(2l+1)/4\pi. So the two expansions are

∑l[∑mYlm(Ω)Ylm(Ω′)‾]=∑l2l+14πPl(cos⁡γ),\sum_{l}\Bigl[\sum_{m}Y_{l}^{m}(\Omega)\overline{Y_{l}^{m}(\Omega')}\Bigr] =\sum_{l}\frac{2l+1}{4\pi}P_{l}(\cos\gamma),

two sums over ll with equal totals.

Second step. Equal totals are not equal terms. What makes the sums match term by term is that each ll-block is rotation invariant: the 2l+12l+1 functions YlmY_{l}^{m} of a fixed ll span a space that a rotation of the axes maps into itself and never mixes with another ll. Choose the new polar axis along Ω′\Omega'; then Pl(cos⁡γ)P_{l}(\cos\gamma) is a multiple of Yl0Y_{l}^{0} in the new frame, hence a combination of the YlmY_{l}^{m} of the old frame with the same ll. So the llth term on the right lives in the llth block, matching the llth term on the left, and the sums agree block by block.

Proof(Parts 1–3, and part 5)AdaptedAdapted from Arfken, Weber & Harris (2013), §15.5, eqs. (15.134), (15.136), (15.138) and (15.148)–(15.149), pp. 757–764; Riley, Hobson & Bence (2006), §18.3, eqs. (18.46) and (18.51), pp. 593–594; Hassani (2013), §13.4, eqs. (13.30)–(13.32) and (13.37), pp. 408–410

Part 1. Write the solid-angle element as dΩ=dφ d(cos⁡θ)\mathrm{d}\Omega=\mathrm{d}\varphi\,\mathrm{d}(\cos\theta) and put x=cos⁡θx=\cos\theta. Then

(Ylm,Yl′m′)=NlmNl′m′∫02πei(m−m′)φdφ∫−11Plm(x)Pl′m′(x) dx,\bigl(Y_{l}^{m},Y_{l'}^{m'}\bigr) =N_{lm}N_{l'm'}\int_{0}^{2\pi}\mathrm{e}^{\mathrm{i}(m-m')\varphi}\mathrm{d}\varphi \int_{-1}^{1}P_{l}^{m}(x)P_{l'}^{m'}(x)\,\mathrm{d}x ,

with NlmN_{lm} the square root in the definition, which is real so no conjugation touches it; the PlmP_{l}^{m} are real as well. The φ\varphi-integral is 2πδmm′2\pi\delta_{mm'} — this is where the azimuthal orthogonality lives, and it is the reason m=m′m=m' may be assumed in the xx-integral. With m=m′m=m' the xx-integral is part 6 of Associated Legendre functions, namely 22l+1(l+m)!(l−m)!δll′\dfrac{2}{2l+1}\dfrac{(l+m)!}{(l-m)!}\delta_{ll'}. Hence

(Ylm,Yl′m′)=δmm′δll′⋅2π⋅2l+14π(l−m)!(l+m)!⋅22l+1(l+m)!(l−m)!=δll′δmm′,\bigl(Y_{l}^{m},Y_{l'}^{m'}\bigr) =\delta_{mm'}\delta_{ll'}\cdot 2\pi\cdot\frac{2l+1}{4\pi}\frac{(l-m)!}{(l+m)!}\cdot\frac{2}{2l+1}\frac{(l+m)!}{(l-m)!} =\delta_{ll'}\delta_{mm'} ,

since 2π⋅2/(4π)=12\pi\cdot 2/(4\pi)=1 and the factorials cancel.

Part 2. Conjugating, Ylm‾=NlmPlm(cos⁡θ)e−imφ\overline{Y_{l}^{m}}=N_{lm}P_{l}^{m}(\cos\theta)\mathrm{e}^{-\mathrm{i}m\varphi}. On the other hand, by the negative-mm relation of Associated Legendre functions,

Yl−m=Nl,−mPl−m(cos⁡θ)e−imφ=2l+14π(l+m)!(l−m)!⋅(−1)m(l−m)!(l+m)!Plm(cos⁡θ) e−imφ.Y_{l}^{-m}=N_{l,-m}P_{l}^{-m}(\cos\theta)\mathrm{e}^{-\mathrm{i}m\varphi} =\sqrt{\frac{2l+1}{4\pi}\frac{(l+m)!}{(l-m)!}}\cdot(-1)^{m}\frac{(l-m)!}{(l+m)!}P_{l}^{m}(\cos\theta)\,\mathrm{e}^{-\mathrm{i}m\varphi} .

The product of the two mm-dependent factors is (−1)m2l+14π(l−m)!(l+m)!=(−1)mNlm(-1)^{m}\sqrt{\tfrac{2l+1}{4\pi}\tfrac{(l-m)!}{(l+m)!}}=(-1)^{m}N_{lm}, so Yl−m=(−1)mYlm‾Y_{l}^{-m}=(-1)^{m}\overline{Y_{l}^{m}}, which rearranges to the stated relation because (−1)m=(−1)−m(-1)^{m}=(-1)^{-m}. Hassani records exactly this as his eq. (13.37). For the parity statement, the antipodal map is θ↦π−θ\theta\mapsto\pi-\theta, φ↦φ+π\varphi\mapsto\varphi+\pi, that is x↦−xx\mapsto-x and eimφ↦(−1)meimφ\mathrm{e}^{\mathrm{i}m\varphi}\mapsto(-1)^{m}\mathrm{e}^{\mathrm{i}m\varphi}; combining with Plm(−x)=(−1)l+mPlm(x)P_{l}^{m}(-x)=(-1)^{l+m}P_{l}^{m}(x) gives the factor (−1)l+m(−1)m=(−1)l(-1)^{l+m}(-1)^{m}=(-1)^{l}, with no mm left.

Part 3. At θ=0\theta=0 the azimuth is meaningless, so any harmonic with a genuine φ\varphi-dependence must vanish there; concretely Plm(±1)=0P_{l}^{m}(\pm1)=0 for m≠0m\neq0, leaving only m=0m=0. For m=0m=0, Nl0=(2l+1)/4πN_{l0}=\sqrt{(2l+1)/4\pi} and Pl(1)=1P_{l}(1)=1, Pl(−1)=(−1)lP_{l}(-1)=(-1)^{l}, which gives the two displayed values. The identity Yl0=(2l+1)/4π Pl(cos⁡θ)Y_{l}^{0}=\sqrt{(2l+1)/4\pi}\,P_{l}(\cos\theta) is the same statement without evaluating θ\theta; it is Hassani's eq. (13.32) and it is the bridge used repeatedly below.

Part 5. Substitute the coefficient formula of part 4 into the expansion of part 4 and exchange the order of summation and integration:

f(Ω)=∑l,m[∫S2Ylm(Ω′)‾f(Ω′) dΩ′]Ylm(Ω)=∫S2f(Ω′)[∑l,mYlm(Ω)Ylm(Ω′)‾]dΩ′.f(\Omega)=\sum_{l,m}\Bigl[\int_{S^{2}}\overline{Y_{l}^{m}(\Omega')}f(\Omega')\,\mathrm{d}\Omega'\Bigr]Y_{l}^{m}(\Omega) =\int_{S^{2}}f(\Omega')\Bigl[\sum_{l,m}Y_{l}^{m}(\Omega)\overline{Y_{l}^{m}(\Omega')}\Bigr]\mathrm{d}\Omega' .

The bracket therefore reproduces ff at Ω\Omega when integrated against ff, for every ff; that is the defining property of δ(Ω−Ω′)\delta(\Omega-\Omega'). (Read as an identity of distributions: the double series does not converge pointwise, exactly as ∑neinθ/2π\sum_{n}\mathrm{e}^{\mathrm{i}n\theta}/2\pi does not. Riley, Hobson and Bence derive the relation in this way and use it immediately in the next proof.) □\square

Proof(Part 6, the addition theorem)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.3, the proof of eq. (18.49) via eqs. (18.51)–(18.53), pp. 594–595; Arfken, Weber & Harris (2013), §15.5, eq. (15.147) and the surrounding rotation argument, p. 763

We follow Riley, Hobson and Bence. Fix Ω′\Omega' and let γ\gamma be the polar angle of Ω\Omega measured from Ω′\Omega'.

Step 1: expand the delta function in Legendre polynomials of γ\gamma. The function δ(Ω−Ω′)\delta(\Omega-\Omega') can only depend on Ω\Omega and Ω′\Omega' through their relative position, which on a sphere means through γ\gamma alone. Expanding a function of the single variable cos⁡γ∈[−1,1]\cos\gamma\in[-1,1] in Legendre polynomials (part 4 of Legendre polynomials),

δ(Ω−Ω′)=∑l≥0bl Pl(cos⁡γ),bl=2l+12∫−11δ(Ω−Ω′) Pl(cos⁡γ) d(cos⁡γ).\delta(\Omega-\Omega')=\sum_{l\ge0}b_{l}\,P_{l}(\cos\gamma), \qquad b_{l}=\frac{2l+1}{2}\int_{-1}^{1}\delta(\Omega-\Omega')\,P_{l}(\cos\gamma)\,\mathrm{d}(\cos\gamma).

Step 2: evaluate blb_{l}. The integrand does not depend on the azimuth ψ\psi about the axis Ω′\Omega', so multiplying by 2π/2π2\pi/2\pi and inserting ∫02πdψ=2π\int_{0}^{2\pi}\mathrm{d}\psi=2\pi turns the one-dimensional integral into an integral over the whole sphere in the coordinates (γ,ψ)(\gamma,\psi):

bl=2l+14π∫02π ⁣ ⁣∫−11δ(Ω−Ω′)Pl(cos⁡γ) d(cos⁡γ) dψ=2l+14π Pl(cos⁡γ)∣Ω=Ω′.b_{l}=\frac{2l+1}{4\pi}\int_{0}^{2\pi}\!\!\int_{-1}^{1}\delta(\Omega-\Omega')P_{l}(\cos\gamma)\,\mathrm{d}(\cos\gamma)\,\mathrm{d}\psi =\frac{2l+1}{4\pi}\,P_{l}(\cos\gamma)\Big|_{\Omega=\Omega'} .

At Ω=Ω′\Omega=\Omega' the separation is γ=0\gamma=0, so cos⁡γ=1\cos\gamma=1 and Pl(1)=1P_{l}(1)=1. Hence

bl=2l+14πfor every l.b_{l}=\frac{2l+1}{4\pi}\qquad\text{for every } l .

Step 3: compare with the closure relation. Part 5 gives a second expansion of the same object, so

∑l∑m=−llYlm(Ω)Ylm(Ω′)‾=∑l2l+14πPl(cos⁡γ).(♯)\sum_{l}\sum_{m=-l}^{l}Y_{l}^{m}(\Omega)\overline{Y_{l}^{m}(\Omega')} =\sum_{l}\frac{2l+1}{4\pi}P_{l}(\cos\gamma) . \tag{$\sharp$}

Both sides are sums over ll. It remains to show that the llth summands are separately equal; that is the whole content of the theorem, and it does not follow from (♯)(\sharp) alone.

Step 4: each ll-block is closed under rotations. Rotate the coordinate axes rigidly and call the new angular coordinates Ωˉ\bar\Omega (the same physical direction, new numbers). The angular eigenvalue equation in the definition above is written in terms of the Laplacian on the sphere, which is rotation invariant; so Ylm(Ωˉ)Y_{l}^{m}(\bar\Omega), as a function of direction, is again an eigenfunction with eigenvalue −l(l+1)-l(l+1). The eigenspace for that eigenvalue is spanned by the 2l+12l+1 functions Ylm′(Ω)Y_{l}^{m'}(\Omega) of the same ll and no other, because distinct ll give distinct eigenvalues −l(l+1)-l(l+1). Therefore

Ylm(Ωˉ)=∑m′=−llDm′ml Ylm′(Ω)Y_{l}^{m}(\bar\Omega)=\sum_{m'=-l}^{l}D^{l}_{m'm}\,Y_{l}^{m'}(\Omega)

for coefficients Dm′mlD^{l}_{m'm} depending only on the rotation — Arfken, Weber and Harris's eq. (15.147), and the reason a rotation cannot change the radial behaviour of a solution of Laplace's equation.

Step 5: identify Pl(cos⁡γ)P_{l}(\cos\gamma) inside the ll-block. Choose the rotation that sends the polar axis to Ω′\Omega'. In the new frame the polar angle of Ω\Omega is exactly γ\gamma, so by part 3,

Pl(cos⁡γ)=4π2l+1  Yl0(Ωˉ)=4π2l+1∑m′=−llDm′0l Ylm′(Ω).P_{l}(\cos\gamma)=\sqrt{\frac{4\pi}{2l+1}}\;Y_{l}^{0}(\bar\Omega) =\sqrt{\frac{4\pi}{2l+1}}\sum_{m'=-l}^{l}D^{l}_{m'0}\,Y_{l}^{m'}(\Omega) .

So, as a function of Ω\Omega for fixed Ω′\Omega', the llth term on the right of (♯)(\sharp) lies in the span of {Ylm′(Ω)}m′\{Y_{l}^{m'}(\Omega)\}_{m'}; and the llth term on the left obviously does too. Since the spans for distinct ll are mutually orthogonal in L2(S2)L^{2}(S^{2}) by part 1, the two sums in (♯)(\sharp) can be matched term by term: project both sides onto the ll-block. This gives

∑m=−llYlm(Ω)Ylm(Ω′)‾=2l+14πPl(cos⁡γ)\sum_{m=-l}^{l}Y_{l}^{m}(\Omega)\overline{Y_{l}^{m}(\Omega')}=\frac{2l+1}{4\pi}P_{l}(\cos\gamma)

for each ll separately, which is the addition theorem.

Step 6: the PP-form. Insert the definition and use eimφ′‾=e−imφ′\overline{\mathrm{e}^{\mathrm{i}m\varphi'}}=\mathrm{e}^{-\mathrm{i}m\varphi'}:

4π2l+1∑m2l+14π(l−m)!(l+m)!Plm(cos⁡θ)Plm(cos⁡θ′)eim(φ−φ′),\frac{4\pi}{2l+1}\sum_{m}\frac{2l+1}{4\pi}\frac{(l-m)!}{(l+m)!}P_{l}^{m}(\cos\theta)P_{l}^{m}(\cos\theta')\mathrm{e}^{\mathrm{i}m(\varphi-\varphi')},

and the prefactors cancel, giving the stated second form. Note that the Condon-Shortley phase appears squared and so drops out: the theorem holds verbatim in either convention of Associated Legendre functions. □\square

Remark(Where the hypotheses are used)Original

mm integral, and ∣m∣≤l\lvert m\rvert\le l. Single-valuedness of the solution in φ\varphi forces eim(φ+2π)=eimφ\mathrm{e}^{\mathrm{i}m(\varphi+2\pi)}=\mathrm{e}^{\mathrm{i}m\varphi}, hence m∈Zm\in\mathbb{Z}; and regularity on the polar axis forces l∈Z≥0l\in\mathbb{Z}_{\ge0} with l≥∣m∣l\ge\lvert m\rvert (part 3 of Associated Legendre functions). Hassani is careful about the first point: in classical physics the function must agree at φ\varphi and φ+2π\varphi+2\pi, which gives integer mm; in quantum mechanics only the absolute value of the function is measurable, so on that criterion alone mm could also be a half-integer. He goes no further with it, and neither do we: his Box 13.4.1 assumes mm integral from that point on, and so does everything here. (Half-integer angular-momentum quantum numbers do occur in physics, but not as solutions of this equation on the sphere — single-valuedness of YlmY_{l}^{m} still excludes them for orbital angular momentum.)

Both factors normalised, and PlmP_{l}^{m} real. Part 1 needs the constant NlmN_{lm} to be real so that it is unaffected by the conjugation, and the xx-integral to be the unit-weight orthogonality relation. The second orthogonality of the associated Legendre functions, the one with weight (1−x2)−1(1-x^{2})^{-1}, plays no part here: the φ\varphi-integral has already separated the different mm, which is why both books that state it call it useless in practice.

Rotation invariance (Step 4). This is the indispensable ingredient, and it is where the proof would fail if attempted by brute force. Without it one has only the aggregate identity (♯)(\sharp), which is far weaker: it would be consistent with the ll-terms being unequal and compensating. Arfken, Weber and Harris make the same point from the other direction, observing that the three Y1mY_{1}^{m} are proportional to z/rz/r and (x±iy)/r(x\pm\mathrm{i}y)/r, so that a rotation of the axes can only produce combinations of them — never a Y2mY_{2}^{m}.

Ω\Omega and Ω′\Omega' enter only through γ\gamma (Step 1). This is what licenses a one-variable expansion of a two-variable object. It holds because δ(Ω−Ω′)\delta(\Omega-\Omega') is defined by a rotation-invariant integral; it would fail for a kernel with a preferred direction, and then no addition theorem of this shape exists.

L2L^{2} rather than pointwise convergence in part 4. Orthonormality plus density gives convergence in norm and nothing more. The uniformly convergent double series that Arfken, Weber and Harris quote carries an unquantified continuity requirement on ff, for which they refer to Hobson rather than proving it; Strauss states the theorem for square-integrable ff, which is the version we take. A discontinuous ff — say a hemisphere at one potential and the other at another, Strauss's Exercise 7 — gives a series that converges in L2L^{2} and oscillates near the equator, exactly as a Fourier series does at a jump.

Finitely many terms per ll. The inner sum in part 4 is finite (2l+12l+1 terms) and this is used tacitly in Step 5, where "the ll-block" is a finite-dimensional space, so that projecting onto it is unproblematic.

Method(Expanding a function on a sphere)AdaptedAdapted from Arfken, Weber & Harris (2013), §15.5, Example 15.5.1, pp. 761–762; Strauss (2008), §10.3, Example 2, pp. 276–277

To expand data g(θ,φ)g(\theta,\varphi) on a sphere, or to solve a boundary-value problem with that data:

  1. Reduce by symmetry first. If gg is independent of φ\varphi, only m=0m=0 survives and the expansion is in Pl(cos⁡θ)P_{l}(\cos\theta). If gg depends on φ\varphi only through cos⁡m0φ\cos m_{0}\varphi or sin⁡m0φ\sin m_{0}\varphi, only m=±m0m=\pm m_{0} survive. If gg is a polynomial in x,y,zx,y,z restricted to the sphere, the expansion terminates, by part 7.
  2. Identify, do not integrate. Write gg in terms of cos⁡θ\cos\theta, sin⁡θ\sin\theta and e±imφ\mathrm{e}^{\pm\mathrm{i}m\varphi} and match it against the table of YlmY_{l}^{m}. This is exact and much faster than evaluating ∫Ylm‾g dΩ\int\overline{Y_{l}^{m}}g\,\mathrm{d}\Omega.
  3. Otherwise project. alm=∫02π ⁣∫0πYlm(θ,φ)‾ g(θ,φ)sin⁡θ dθ dφa_{lm}=\int_{0}^{2\pi}\!\int_{0}^{\pi}\overline{Y_{l}^{m}(\theta,\varphi)}\,g(\theta,\varphi)\sin\theta\,\mathrm{d}\theta\,\mathrm{d}\varphi. Do the φ\varphi-integral first: it is a Fourier coefficient and kills all but one mm.
  4. For the interior Dirichlet problem ∇2u=0\nabla^{2}u=0 in r<ar<a with u=gu=g on r=ar=a: discard the r−l−1r^{-l-1} solutions, which are singular at the origin, and write u(r,θ,φ)=∑l,malm(ra)lYlm(θ,φ),u(r,\theta,\varphi)=\sum_{l,m}a_{lm}\Bigl(\frac{r}{a}\Bigr)^{l}Y_{l}^{m}(\theta,\varphi) , which matches gg at r=ar=a term by term. For the exterior problem, keep r−l−1r^{-l-1} instead and use (a/r)l+1(a/r)^{l+1}.
  5. For the Helmholtz or Schrödinger equation, replace rlr^{l} by the appropriate radial solution — jl(kr)j_{l}(kr) and yl(kr)y_{l}(kr) of Spherical Bessel functions for Helmholtz (The Helmholtz equation), the Laguerre-type radial function for the Coulomb problem — and keep the angular part unchanged. The angular factor never depends on which of these equations is being solved.
  6. To collapse a double sum over mm that has the structure ∑mYlm(Ω)Ylm(Ω′)‾\sum_{m}Y_{l}^{m}(\Omega)\overline{Y_{l}^{m}(\Omega')}, use the addition theorem backwards: it becomes 2l+14πPl(cos⁡γ)\frac{2l+1}{4\pi}P_{l}(\cos\gamma), a single term.
Example(A Dirichlet problem on a ball, and the addition theorem checked numerically)Original

Part 1: the boundary-value problem. Solve ∇2u=0\nabla^{2}u=0 inside the ball r<ar<a with

u(a,θ,φ)=sin⁡2θ cos⁡2φ.u(a,\theta,\varphi)=\sin^{2}\theta\,\cos2\varphi .

Step 1: expand the data. Write cos⁡2φ=12(e2iφ+e−2iφ)\cos2\varphi=\tfrac12(\mathrm{e}^{2\mathrm{i}\varphi}+\mathrm{e}^{-2\mathrm{i}\varphi}). From the table, Y2±2=15/32π sin⁡2θ e±2iφY_{2}^{\pm2}=\sqrt{15/32\pi}\,\sin^{2}\theta\,\mathrm{e}^{\pm2\mathrm{i}\varphi}, so

sin⁡2θ e±2iφ=32π15  Y2±2,\sin^{2}\theta\,\mathrm{e}^{\pm2\mathrm{i}\varphi}=\sqrt{\frac{32\pi}{15}}\;Y_{2}^{\pm2},

and therefore

g=sin⁡2θcos⁡2φ=c (Y22+Y2−2),c=1232π15=8π15=1.2944173….g=\sin^{2}\theta\cos2\varphi=c\,\bigl(Y_{2}^{2}+Y_{2}^{-2}\bigr), \qquad c=\frac12\sqrt{\frac{32\pi}{15}}=\sqrt{\frac{8\pi}{15}}=1.2944173\ldots .

Only l=2l=2, m=±2m=\pm2 occur: by step 1 of the method this had to happen, since gg carries cos⁡2φ\cos 2\varphi and is a quadratic polynomial on the sphere.

Step 2: check the coefficient by projection. By part 4, a22=∫S2Y22‾ g dΩa_{22}=\int_{S^{2}}\overline{Y_{2}^{2}}\,g\,\mathrm{d}\Omega. Substituting the expansion and using orthonormality, a22=ca_{22}=c, and likewise a2,−2=ca_{2,-2}=c. As an independent check, Parseval must give

∥g∥2=∣a22∣2+∣a2,−2∣2=2c2=16π15.\lVert g\rVert^{2}=\lvert a_{22}\rvert^{2}+\lvert a_{2,-2}\rvert^{2}=2c^{2}=\frac{16\pi}{15} .

Directly:

∫S2sin⁡4θcos⁡22φ dΩ=∫02πcos⁡22φ dφ∫0πsin⁡5θ dθ=π⋅1615=16π15,\int_{S^{2}}\sin^{4}\theta\cos^{2}2\varphi\,\mathrm{d}\Omega =\int_{0}^{2\pi}\cos^{2}2\varphi\,\mathrm{d}\varphi\int_{0}^{\pi}\sin^{5}\theta\,\mathrm{d}\theta =\pi\cdot\frac{16}{15}=\frac{16\pi}{15},

using ∫02πcos⁡22φ dφ=π\int_{0}^{2\pi}\cos^{2}2\varphi\,\mathrm{d}\varphi=\pi and ∫0πsin⁡5θ dθ=∫−11(1−x2)2dx=2−43+25=1615\int_{0}^{\pi}\sin^{5}\theta\,\mathrm{d}\theta=\int_{-1}^{1}(1-x^{2})^{2}\mathrm{d}x=2-\tfrac43+\tfrac25=\tfrac{16}{15}. The two agree.

Step 3: solve. By step 4 of the method, the interior solution is

u(r,θ,φ)=c(ra)2(Y22+Y2−2)=(ra)2sin⁡2θ cos⁡2φ.u(r,\theta,\varphi)=c\Bigl(\frac{r}{a}\Bigr)^{2}\bigl(Y_{2}^{2}+Y_{2}^{-2}\bigr) =\Bigl(\frac{r}{a}\Bigr)^{2}\sin^{2}\theta\,\cos2\varphi .

Step 4: the checkable number, and a sanity check. At r=a/2r=a/2, θ=π/2\theta=\pi/2, φ=0\varphi=0:

u=(12)2⋅1⋅1=14.u=\Bigl(\tfrac12\Bigr)^{2}\cdot 1\cdot 1=\frac14 .

That the answer is harmonic is visible in Cartesian form, which is part 7 at work: r2sin⁡2θcos⁡2φ=r2sin⁡2θ(cos⁡2φ−sin⁡2φ)=x2−y2r^{2}\sin^{2}\theta\cos2\varphi=r^{2}\sin^{2}\theta(\cos^{2}\varphi-\sin^{2}\varphi)=x^{2}-y^{2}, so

u=x2−y2a2,∇2u=2−2a2=0,u=\frac{x^{2}-y^{2}}{a^{2}}, \qquad \nabla^{2}u=\frac{2-2}{a^{2}}=0 ,

and at the test point (x,y,z)=(a/2,0,0)(x,y,z)=(a/2,0,0) indeed u=(a/2)2/a2=1/4u=(a/2)^{2}/a^{2}=1/4.

Part 2: the addition theorem, numerically. Take

Ω=(θ,φ)=(π/3, 0),Ω′=(θ′,φ′)=(π/4, π/2).\Omega=(\theta,\varphi)=(\pi/3,\,0), \qquad \Omega'=(\theta',\varphi')=(\pi/4,\,\pi/2).

From part 6,

cos⁡γ=cos⁡π3cos⁡π4+sin⁡π3sin⁡π4cos⁡(−π2)=12⋅22+0=24=0.3535534…,\cos\gamma=\cos\tfrac{\pi}{3}\cos\tfrac{\pi}{4}+\sin\tfrac{\pi}{3}\sin\tfrac{\pi}{4}\cos\bigl(-\tfrac{\pi}{2}\bigr) =\tfrac12\cdot\tfrac{\sqrt2}{2}+0=\frac{\sqrt2}{4}=0.3535534\ldots,

the second term vanishing because cos⁡(φ−φ′)=cos⁡(−π/2)=0\cos(\varphi-\varphi')=\cos(-\pi/2)=0. Now evaluate both sides for l=2l=2. The right-hand side of the theorem is a sum of five terms; computing them from the table gives

4π5∑m=−22Y2m(Ω)Y2m(Ω′)‾=−0.3125000,\frac{4\pi}{5}\sum_{m=-2}^{2}Y_{2}^{m}(\Omega)\overline{Y_{2}^{m}(\Omega')}=-0.3125000 ,

while the left-hand side is

P2(cos⁡γ)=3cos⁡2γ−12=3⋅18−12=0.375−12=−0.3125P_{2}(\cos\gamma)=\frac{3\cos^{2}\gamma-1}{2}=\frac{3\cdot\tfrac18-1}{2}=\frac{0.375-1}{2}=-0.3125

exactly, since cos⁡2γ=2/16=1/8\cos^{2}\gamma=2/16=1/8. The same test at l=1l=1 gives 0.35355340.3535534 on both sides (P1(cos⁡γ)=cos⁡γP_{1}(\cos\gamma)=\cos\gamma), and at l=3l=3 gives −0.4198447-0.4198447 on both. Note how much the theorem saves: the five-term sum on the right collapses to one evaluation of a quadratic. All numbers in this example were checked symbolically or to ten digits numerically.

ApplicationAdaptedAdapted from Arfken, Weber & Harris (2013), §15.5, Example 15.5.2 with eq. (15.146) and Table 15.5, p. 762, and eq. (15.152) with the dipole selection rule, p. 764; Strauss (2008), §10.3, Example 2, pp. 276–277

Geodesy and planetary science. Arfken, Weber and Harris give the gravitational potential of the Earth, Moon and Mars as a Laplace series in the real spherical harmonics, their eq. (15.146),

U(r,θ,φ)=GMR[−Rr+∑l≥2∑m=0l(Rr)l+1(ClmYmle+SlmYmlo)],U(r,\theta,\varphi)=\frac{GM}{R}\Bigl[-\frac{R}{r}+\sum_{l\ge2}\sum_{m=0}^{l}\Bigl(\frac{R}{r}\Bigr)^{l+1}\bigl(C_{lm}Y^{e}_{ml}+S_{lm}Y^{o}_{ml}\bigr)\Bigr],

with MM the mass, RR the equatorial radius. The coefficient C20C_{20} measures the equatorial bulge and C22C_{22}, S22S_{22} the azimuthal asymmetry. Their measured values illustrate how the shapes differ: C20=1.083×10−3C_{20}=1.083\times10^{-3} for the Earth, (0.200±0.002)×10−3(0.200\pm0.002)\times10^{-3} for the Moon, (1.96±0.01)×10−3(1.96\pm0.01)\times10^{-3} for Mars. Every satellite geodesy mission, and every orbit propagator, is an exercise in estimating these numbers; the l=1l=1 terms are absent because the origin is taken at the centre of mass.

Multipole expansions and the addition theorem. Combining the addition theorem with the inverse-distance expansion of Legendre polynomials turns

1∣r−r′∣=∑lr′ lrl+1Pl(cos⁡γ)into1∣r−r′∣=∑l,m4π2l+1r′ lrl+1Ylm(Ω)Ylm(Ω′)‾,\frac{1}{\lvert\mathbf{r}-\mathbf{r}'\rvert}=\sum_{l}\frac{r'^{\,l}}{r^{l+1}}P_{l}(\cos\gamma) \quad\text{into}\quad \frac{1}{\lvert\mathbf{r}-\mathbf{r}'\rvert}=\sum_{l,m}\frac{4\pi}{2l+1}\frac{r'^{\,l}}{r^{l+1}}Y_{l}^{m}(\Omega)\overline{Y_{l}^{m}(\Omega')} ,

in which the two directions are separated: the source variables (r′,Ω′)(r',\Omega') and the field variables (r,Ω)(r,\Omega) no longer appear in a single entangled factor Pl(cos⁡γ)P_{l}(\cos\gamma). This factorisation is what makes a multipole expansion computable at all (Multipole expansions) and is the mathematical core of the fast multipole method, which reduces an NN-body potential evaluation from O(N2)O(N^{2}) to O(N)O(N) work.

Quantum mechanics. YlmY_{l}^{m} is the angular part of every central-field wavefunction: the atomic orbitals s,p,d,fs,p,d,f are l=0,1,2,3l=0,1,2,3, the (2l+1)(2l+1)-fold degeneracy in mm is the rotational symmetry, and the Condon-Shortley phase is in the tables so that the ladder operators act with uniform signs. Arfken, Weber and Harris's recurrence in cos⁡θ\cos\theta gives at once

∫Yl′m′‾ cos⁡θ  Ylm dΩ=[(l−m+1)(l+m+1)(2l+1)(2l+3)]1/2δm′mδl′,l+1+[(l−m)(l+m)(2l−1)(2l+1)]1/2δm′mδl′,l−1,\int\overline{Y_{l'}^{m'}}\,\cos\theta\;Y_{l}^{m}\,\mathrm{d}\Omega =\Bigl[\tfrac{(l-m+1)(l+m+1)}{(2l+1)(2l+3)}\Bigr]^{1/2}\delta_{m'm}\delta_{l',l+1} +\Bigl[\tfrac{(l-m)(l+m)}{(2l-1)(2l+1)}\Bigr]^{1/2}\delta_{m'm}\delta_{l',l-1},

which is the electric-dipole selection rule Δl=±1\Delta l=\pm1, Δm=0\Delta m=0 for zz-polarised light: the integral vanishes unless ll changes by one.

Application(Spherical harmonics as a storage format)Original

Geophysics, cosmology and graphics. The cosmic microwave background temperature map is reported as the power spectrum Cl=12l+1∑m∣alm∣2C_{l}=\frac{1}{2l+1}\sum_{m}\lvert a_{lm}\rvert^{2} of its Laplace series, and the rotation invariance of each ll-block (Step 4 above) is precisely why ClC_{l}, and not the individual alma_{lm}, is the statistically meaningful quantity. Global seismic tomography, atmospheric and ocean circulation models, and the ambient-light "irradiance probes" of real-time computer graphics all store a field on a sphere as the first few dozen of its spherical-harmonic coefficients, for the same reason a signal is stored as Fourier coefficients: the expansion is orthonormal, it is truncatable with controlled L2L^{2} error, and the truncation is rotation-covariant.

Definition

Hermite polynomials

Adapted

The polynomials generated by e−t2+2tx\mathrm{e}^{-t^{2}+2tx}: their Rodrigues formula, three-term recurrence, Hermite equation and orthogonality with weight e−x2\mathrm{e}^{-x^{2}} on the whole line, the Hermite functions they build, and the two incompatible conventions (physicists' HnH_n, probabilists' HenHe_n).

Prerequisites: General properties of orthogonal polynomials, Sturm-Liouville eigenvalue problems, Orthonormal bases in Hilbert space, Fubini and Tonelli theorems

Definition(Hermite polynomials)AdaptedAdapted from Arfken, Weber & Harris (2013), §18.1, eq. (18.2), p. 872, and eq. (18.8), p. 874; Hassani (2013), §8.4.1, the standardisation $K_n=(-1)^n$ and eq. (8.13), p. 248; Riley, Hobson & Bence (2006), §18.9.1, eq. (18.130), p. 626; Kreyszig (1978), §3.7-2, eq. (7b), p. 182

The Hermite polynomials H0,H1,H2,…H_{0},H_{1},H_{2},\dots are the coefficients in the expansion of

g(x,t)  =  e−t2+2tx  =  ∑n=0∞Hn(x) tnn!,g(x,t) \;=\; \mathrm{e}^{-t^{2}+2tx} \;=\; \sum_{n=0}^{\infty} H_{n}(x)\,\frac{t^{n}}{n!} ,

which converges for every x∈Rx\in\mathbb{R} and every t∈Ct\in\mathbb{C}, because t↦g(x,t)t\mapsto g(x,t) is entire. Equivalently, Hn(x)=∂tng(x,t)∣t=0H_{n}(x)=\partial_{t}^{n}g(x,t)\big|_{t=0}.

Arfken, Weber and Harris take this generating function as the definition, so that nothing has to be imported from the general theory; Hassani, Riley, Hobson and Bence, and Kreyszig instead start from the Rodrigues formula

Hn(x)  =  (−1)n ex2 dndxn e−x2,H_{n}(x) \;=\; (-1)^{n}\,\mathrm{e}^{x^{2}}\,\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\,\mathrm{e}^{-x^{2}} ,

and the two agree: the statement below derives each from the other. In the language of General properties of orthogonal polynomials, HnH_n is the classical family belonging to the weight w(x)=e−x2w(x)=\mathrm{e}^{-x^{2}} on (−∞,∞)(-\infty,\infty), with the scale fixed by the choice Hassani writes as Kn=(−1)nK_{n}=(-1)^{n}. That scale is a convention and nothing more; it makes the leading coefficient of HnH_{n} equal to 2n2^{n} rather than 11.

The first few are

H0=1,H1=2x,H2=4x2−2,H3=8x3−12x,H4=16x4−48x2+12,H5=32x5−160x3+120x.\begin{aligned} H_{0}&=1, & H_{1}&=2x, & H_{2}&=4x^{2}-2,\\ H_{3}&=8x^{3}-12x, & H_{4}&=16x^{4}-48x^{2}+12, & H_{5}&=32x^{5}-160x^{3}+120x . \end{aligned}
Definition(Hermite functions)AdaptedAdapted from Arfken, Weber & Harris (2013), §18.1, eqs. (18.12)–(18.13), pp. 875–876; Kreyszig (1978), §3.7-2, eq. (7a), p. 181; Reed & Simon (1980), Appendix to §V.3, the display defining $\phi_n$ and the eigenvalue equation below it, p. 142

The Hermite functions are the Hermite polynomials with half of the weight absorbed into them. Arfken, Weber and Harris write the unnormalised version

φn(x)  =  e−x2/2Hn(x),\varphi_{n}(x) \;=\; \mathrm{e}^{-x^{2}/2}H_{n}(x) ,

which is no longer a polynomial. Kreyszig and Reed and Simon use the normalised version, which we write

ψn(x)  =  (2n n! π )−1/2 e−x2/2Hn(x),n=0,1,2,…\psi_{n}(x) \;=\; \bigl(2^{n}\,n!\,\sqrt{\pi}\,\bigr)^{-1/2}\,\mathrm{e}^{-x^{2}/2}H_{n}(x) , \qquad n=0,1,2,\dots

Kreyszig obtains exactly these functions by running Gram-Schmidt (Orthonormal bases and Gram-Schmidt) on the sequence tne−t2/2t^{n}\mathrm{e}^{-t^{2}/2} in the real space L2(−∞,∞)L^{2}(-\infty,\infty); Reed and Simon obtain them instead by repeated application of A†=2−1/2(x−d/dx)A^{\dagger}=2^{-1/2}(x-\mathrm{d}/\mathrm{d}x) to the Gaussian ψ0=π−1/4e−x2/2\psi_{0}=\pi^{-1/4}\mathrm{e}^{-x^{2}/2}, as ψn=(n!)−1/2(A†)nψ0\psi_{n}=(n!)^{-1/2}(A^{\dagger})^{n}\psi_{0}. The three descriptions give the same functions, with the same signs.

Definition(What the Hermite polynomials satisfy)AdaptedAdapted from Arfken, Weber & Harris (2013), §18.1, eqs. (18.1), (18.3)–(18.9), (18.11), (18.13) and (18.15), pp. 871–876; Riley, Hobson & Bence (2006), §18.9, eqs. (18.126) and (18.129), pp. 624–625, and §18.9.1, eqs. (18.130), (18.132)–(18.135), pp. 626–628; Hassani (2013), §8.4.1, eqs. (8.13)–(8.15) and the values $k_n^{(n)}=2^{n}$ and $h_n=\sqrt{\pi}2^{n}n!$, pp. 248–249; Kreyszig (1978), §3.7-2, eqs. (7a)–(7c), the orthonormality (8) and the Hermite equation (9), pp. 181–184; Reed & Simon (1980), Appendix to §V.3, Lemma 3 and the eigenvalue equation above it, p. 142

Let HnH_{n} be defined by the generating function above and let ψn\psi_{n} be the normalised Hermite functions. Write H−1≡0H_{-1}\equiv0.

  1. Explicit form. For n≥0n\ge0, Hn(x)  =  ∑s=0⌊n/2⌋(−1)s n!(n−2s)!  s! (2x)n−2s.H_{n}(x) \;=\; \sum_{s=0}^{\lfloor n/2\rfloor}\frac{(-1)^{s}\,n!}{(n-2s)!\;s!}\,(2x)^{n-2s} . In particular HnH_{n} is a polynomial of degree exactly nn with leading coefficient 2n2^{n}, the coefficient of xn−1x^{n-1} vanishes, and Hn(−x)=(−1)nHn(x),H2m(0)=(−1)m(2m)!m!,H2m+1(0)=0.H_{n}(-x)=(-1)^{n}H_{n}(x), \qquad H_{2m}(0)=(-1)^{m}\frac{(2m)!}{m!}, \qquad H_{2m+1}(0)=0 .
  2. Rodrigues formula. Hn(x)=(−1)nex2dndxne−x2\displaystyle H_{n}(x)=(-1)^{n}\mathrm{e}^{x^{2}}\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\mathrm{e}^{-x^{2}}.
  3. Recurrences. For n≥0n\ge0, Hn+1(x)=2x Hn(x)−2n Hn−1(x),Hn′(x)=2n Hn−1(x).H_{n+1}(x) = 2x\,H_{n}(x) - 2n\,H_{n-1}(x), \qquad H_{n}'(x) = 2n\,H_{n-1}(x) .
  4. Hermite's equation. HnH_{n} solves y′′−2x y′+2n y=0.y'' - 2x\,y' + 2n\,y = 0 . Riley, Hobson and Bence state this equation with a general real parameter ν\nu in place of nn; it is exactly when ν\nu is a non-negative integer that it has a polynomial solution.
  5. Orthogonality and norm. For all m,n≥0m,n\ge0, ∫−∞∞e−x2Hm(x)Hn(x) dx  =  2n n! π  δmn.\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}H_{m}(x)H_{n}(x)\,\mathrm{d}x \;=\; 2^{n}\,n!\,\sqrt{\pi}\;\delta_{mn} . Equivalently the ψn\psi_{n} are orthonormal in L2(R)L^{2}(\mathbb{R}) with unit weight.
  6. The eigenvalue problem solved. Each ψn\psi_{n} satisfies −ψn′′(x)+x2ψn(x)  =  (2n+1) ψn(x),-\psi_{n}''(x) + x^{2}\psi_{n}(x) \;=\; (2n+1)\,\psi_{n}(x) , which is the self-adjoint form Arfken, Weber and Harris write as φn′′+(2n+1−x2)φn=0\varphi_{n}''+(2n+1-x^{2})\varphi_{n}=0 and Reed and Simon as (−d2/dx2+x2)ϕn=(2n+1)ϕn(-\mathrm{d}^{2}/\mathrm{d}x^{2}+x^{2})\phi_{n}=(2n+1)\phi_{n}.
  7. Completeness. {ψn}n≥0\{\psi_{n}\}_{n\ge0} is an orthonormal basis of L2(R)L^{2}(\mathbb{R}): this is Lemma 3 of the appendix to Reed and Simon's §V.3, and it is the totality statement Kreyszig records (and uses to conclude that L2(R)L^{2}(\mathbb{R}) is separable). We state it here on their authority and do not prove it; items 1–6 are proved below. Consequently every f∈L2(R)f\in L^{2}(\mathbb{R}) has the mean-square convergent expansion f=∑n(f,ψn)ψnf=\sum_{n}(f,\psi_{n})\psi_{n}, which Riley, Hobson and Bence write in the unnormalised form f=∑nanHnf=\sum_{n}a_{n}H_{n} with an=12nn!π∫−∞∞f(x)Hn(x) e−x2 dx.a_{n}=\frac{1}{2^{n}n!\sqrt{\pi}}\int_{-\infty}^{\infty}f(x)H_{n}(x)\,\mathrm{e}^{-x^{2}}\,\mathrm{d}x .
Idea of the proofOriginal

Everything follows from one algebraic accident. The exponent −t2+2tx-t^{2}+2tx completes the square two different ways: as x2−(t−x)2x^{2}-(t-x)^{2}, and as a product of two exponentials e2txe−t2\mathrm{e}^{2tx}\mathrm{e}^{-t^{2}} whose separate series are known.

The first completion makes tt and xx interchangeable up to a sign, because ∂te−(t−x)2=−∂xe−(t−x)2\partial_{t}\mathrm{e}^{-(t-x)^{2}}=-\partial_{x}\mathrm{e}^{-(t-x)^{2}}; differentiating nn times in tt and setting t=0t=0 therefore turns into differentiating nn times in xx, which is the Rodrigues formula. The second completion multiplies out to the explicit coefficient sum, and the sum shows at a glance that HnH_{n} has degree nn, which powers it contains and what it is at the origin.

The recurrences come from the two first-order partial differential equations ∂tg=(2x−2t)g\partial_{t}g=(2x-2t)g and ∂xg=2tg\partial_{x}g=2tg that the exponential satisfies, read coefficient by coefficient; eliminating Hn−1H_{n-1} between them and differentiating once gives Hermite's equation. Orthogonality is the same trick applied twice: multiply two copies of the generating function, in variables ss and tt, by the weight e−x2\mathrm{e}^{-x^{2}}, and the xx-integral is a Gaussian that evaluates to π e2st\sqrt{\pi}\,\mathrm{e}^{2st}. A function of the product stst alone has no smtns^{m}t^{n} term with m≠nm\neq n, so orthogonality and the value of the norm drop out of the same line.

ProofAdaptedAdapted from Arfken, Weber & Harris (2013), §18.1, the derivations surrounding eqs. (18.3)–(18.4), p. 872, eq. (18.8), p. 874, and eqs. (18.14)–(18.15), p. 876; Riley, Hobson & Bence (2006), §18.9.1, the worked derivation of the recurrences and the normalisation, pp. 626–628; Hassani (2013), §8.7, Problem 8.13, p. 260

Item 1, the explicit sum. Multiply the two absolutely convergent exponential series

e2tx=∑j≥0(2x)jj!tj,e−t2=∑s≥0(−1)ss!t2s,\mathrm{e}^{2tx}=\sum_{j\ge0}\frac{(2x)^{j}}{j!}t^{j}, \qquad \mathrm{e}^{-t^{2}}=\sum_{s\ge0}\frac{(-1)^{s}}{s!}t^{2s},

and collect the terms with tnt^{n}, which are those with j+2s=nj+2s=n, that is 0≤s≤⌊n/2⌋0\le s\le\lfloor n/2\rfloor and j=n−2sj=n-2s:

g(x,t)=∑n≥0tn∑s=0⌊n/2⌋(−1)s(2x)n−2s(n−2s)! s!.g(x,t)=\sum_{n\ge0}t^{n}\sum_{s=0}^{\lfloor n/2\rfloor}\frac{(-1)^{s}(2x)^{n-2s}}{(n-2s)!\,s!} .

Comparing with ∑nHntn/n!\sum_n H_n t^{n}/n! and multiplying by n!n! gives the stated sum. The rearrangement is legitimate because both series converge absolutely for every xx and tt.

Reading the sum: the s=0s=0 term contributes (2x)n(2x)^{n} and no other term reaches degree nn, so the degree is nn and the leading coefficient 2n2^{n}; every exponent n−2sn-2s has the parity of nn, which gives both the absence of an xn−1x^{n-1} term and the parity relation Hn(−x)=(−1)nHn(x)H_{n}(-x)=(-1)^{n}H_{n}(x); and at x=0x=0 the only surviving term is the one with n−2s=0n-2s=0, which exists only for even n=2mn=2m and equals (−1)m(2m)!/m!(-1)^{m}(2m)!/m!.

We also record a bound for later use. Replacing every term of the double sum by its absolute value and running the computation backwards,

∑n≥0∣Hn(x)∣∣t∣nn!  ≤  ∑j,s≥0(2∣x∣)j∣t∣jj!∣t∣2ss!  =  e2∣x∣∣t∣+t2.(∗)\sum_{n\ge0}\bigl|H_{n}(x)\bigr|\frac{|t|^{n}}{n!} \;\le\;\sum_{j,s\ge0}\frac{(2|x|)^{j}|t|^{j}}{j!}\frac{|t|^{2s}}{s!} \;=\;\mathrm{e}^{2|x||t|+t^{2}} . \tag{$\ast$}

Item 2, the Rodrigues formula. Complete the square in the exponent the other way:

−t2+2tx=x2−(t−x)2,sog(x,t)=ex2 e−(t−x)2.-t^{2}+2tx = x^{2}-(t-x)^{2}, \qquad\text{so}\qquad g(x,t)=\mathrm{e}^{x^{2}}\,\mathrm{e}^{-(t-x)^{2}} .

The function e−(t−x)2\mathrm{e}^{-(t-x)^{2}} depends on tt and xx only through t−xt-x, so ∂t\partial_{t} acting on it equals −∂x-\partial_{x} acting on it. Applying this nn times, and noting that ex2\mathrm{e}^{x^{2}} is a constant for ∂t\partial_{t},

Hn(x)=∂tng∣t=0=ex2 ∂tne−(t−x)2∣t=0=(−1)nex2 ∂xne−(t−x)2∣t=0=(−1)nex2dndxne−x2,H_{n}(x)=\partial_{t}^{n}g\big|_{t=0} =\mathrm{e}^{x^{2}}\,\partial_{t}^{n}\mathrm{e}^{-(t-x)^{2}}\Big|_{t=0} =(-1)^{n}\mathrm{e}^{x^{2}}\,\partial_{x}^{n}\mathrm{e}^{-(t-x)^{2}}\Big|_{t=0} =(-1)^{n}\mathrm{e}^{x^{2}}\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\mathrm{e}^{-x^{2}} ,

where the last step sets t=0t=0 before differentiating in xx, which is allowed because e−(t−x)2\mathrm{e}^{-(t-x)^2} is jointly smooth and the tt-derivatives have already been taken.

Item 3, the recurrences. The exponential satisfies

∂g∂t=(2x−2t) g,∂g∂x=2t g,\frac{\partial g}{\partial t}=(2x-2t)\,g , \qquad \frac{\partial g}{\partial x}=2t\,g ,

by direct differentiation. Insert the series into the first identity. On the left, ∂t∑nHntn/n!=∑n≥0Hn+1tn/n!\partial_{t}\sum_{n}H_{n}t^{n}/n!=\sum_{n\ge0}H_{n+1}t^{n}/n! after shifting the index. On the right,

2x∑n≥0Hntnn!−2∑n≥0Hntn+1n!=2x∑n≥0Hntnn!−2∑n≥1n Hn−1tnn!,2x\sum_{n\ge0}H_{n}\frac{t^{n}}{n!}-2\sum_{n\ge0}H_{n}\frac{t^{n+1}}{n!} =2x\sum_{n\ge0}H_{n}\frac{t^{n}}{n!}-2\sum_{n\ge1}n\,H_{n-1}\frac{t^{n}}{n!} ,

where the second sum was re-indexed using tn+1/n!=n′ tn′/n′!t^{n+1}/n!=n'\,t^{n'}/n'! with n′=n+1n'=n+1. Term-by-term differentiation and re-indexing are legitimate because the series is a power series with infinite radius of convergence. Two power series in tt agree only if their coefficients agree, so Hn+1=2xHn−2nHn−1H_{n+1}=2xH_{n}-2nH_{n-1} for n≥1n\ge1; for n=0n=0 the second sum is empty and the identity reads H1=2xH0H_{1}=2xH_{0}, which is the same statement with the convention H−1=0H_{-1}=0.

The second identity gives, in the same way,

∑n≥0Hn′tnn!=2∑n≥1n Hn−1tnn!,\sum_{n\ge0}H_{n}'\frac{t^{n}}{n!}=2\sum_{n\ge1}n\,H_{n-1}\frac{t^{n}}{n!},

hence Hn′=2nHn−1H_{n}'=2nH_{n-1} for n≥1n\ge1; at n=0n=0 the right-hand sum is empty and the identity reads H0′=0H_{0}'=0, which is again the stated relation with the convention H−1=0H_{-1}=0. Both recurrences therefore hold for every n≥0n\ge0, as item 3 asserts.

Item 4, Hermite's equation. Use 2nHn−1=Hn′2nH_{n-1}=H_{n}' to eliminate Hn−1H_{n-1} from the three-term recurrence:

Hn+1=2xHn−Hn′.H_{n+1}=2xH_{n}-H_{n}' .

Differentiate this in xx:

Hn+1′=2Hn+2xHn′−Hn′′.H_{n+1}'=2H_{n}+2xH_{n}'-H_{n}'' .

On the left, the derivative recurrence at index n+1n+1 says Hn+1′=2(n+1)HnH_{n+1}'=2(n+1)H_{n}. Equating and cancelling 2Hn2H_{n} from both sides leaves

2n Hn=2x Hn′−Hn′′,2n\,H_{n}=2x\,H_{n}'-H_{n}'' ,

which is Hermite's equation. Each hypothesis used here is one of the two recurrences, so this holds for every n≥0n\ge0.

Item 5, orthogonality and the norm. Take two copies of the generating function, in variables ss and tt, multiply by the weight, and integrate:

∫−∞∞e−x2 g(x,s) g(x,t) dx=∫−∞∞e−x2−s2+2sx−t2+2tx dx.\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}\,g(x,s)\,g(x,t)\,\mathrm{d}x =\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}-s^{2}+2sx-t^{2}+2tx}\,\mathrm{d}x .

The exponent is −(x−s−t)2+2st-\bigl(x-s-t\bigr)^{2}+2st, as one checks by expanding the square. So the integral is e2st\mathrm{e}^{2st} times a shifted Gaussian integral, and

∫−∞∞e−x2g(x,s)g(x,t) dx=π e2st=π∑k≥02ksktkk!.\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}g(x,s)g(x,t)\,\mathrm{d}x=\sqrt{\pi}\,\mathrm{e}^{2st} =\sqrt{\pi}\sum_{k\ge0}\frac{2^{k}s^{k}t^{k}}{k!} .

Now expand the left-hand side instead. The double series may be integrated term by term because, by the bound (∗)(\ast),

∫−∞∞e−x2∑m,n∣Hm(x)Hn(x)∣∣s∣m∣t∣nm! n! dx  ≤  ∫−∞∞e−x2e2∣x∣(∣s∣+∣t∣)+s2+t2 dx  <  ∞,\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}\sum_{m,n}\bigl|H_{m}(x)H_{n}(x)\bigr|\frac{|s|^{m}|t|^{n}}{m!\,n!}\,\mathrm{d}x \;\le\;\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}\mathrm{e}^{2|x|(|s|+|t|)+s^{2}+t^{2}}\,\mathrm{d}x \;<\;\infty ,

the last integral being finite because the Gaussian beats the exponential; Tonelli's theorem then licenses the interchange (Fubini and Tonelli theorems). Therefore

∑m,n≥0smtnm! n!∫−∞∞e−x2HmHn dx=π∑k≥02kk!sktk.\sum_{m,n\ge0}\frac{s^{m}t^{n}}{m!\,n!}\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}H_{m}H_{n}\,\mathrm{d}x =\sqrt{\pi}\sum_{k\ge0}\frac{2^{k}}{k!}s^{k}t^{k} .

Both sides are power series in the two variables ss and tt, convergent on a neighbourhood of the origin, so their coefficients match one by one. The right-hand side has no smtns^{m}t^{n} term with m≠nm\neq n, which gives orthogonality; and matching the sntns^{n}t^{n} terms gives

1(n!)2∫−∞∞e−x2Hn2 dx=π 2nn!,that is∫−∞∞e−x2Hn2 dx=2nn!π.\frac{1}{(n!)^{2}}\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}H_{n}^{2}\,\mathrm{d}x=\frac{\sqrt{\pi}\,2^{n}}{n!}, \qquad\text{that is}\qquad \int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}H_{n}^{2}\,\mathrm{d}x=2^{n}n!\sqrt{\pi} .

Dividing by the two normalising constants shows (ψm,ψn)=δmn(\psi_{m},\psi_{n})=\delta_{mn}.

Item 6, the eigenvalue equation. Put ψn=cne−x2/2Hn\psi_{n}=c_{n}\mathrm{e}^{-x^{2}/2}H_{n} with cnc_{n} the constant above. Then

ψn′=cne−x2/2(Hn′−xHn),ψn′′=cne−x2/2(Hn′′−2xHn′+(x2−1)Hn).\psi_{n}'=c_{n}\mathrm{e}^{-x^{2}/2}\bigl(H_{n}'-xH_{n}\bigr), \qquad \psi_{n}''=c_{n}\mathrm{e}^{-x^{2}/2}\bigl(H_{n}''-2xH_{n}'+(x^{2}-1)H_{n}\bigr).

By item 4, Hn′′−2xHn′=−2nHnH_{n}''-2xH_{n}'=-2nH_{n}, so ψn′′=cne−x2/2(x2−1−2n)Hn=(x2−2n−1)ψn\psi_{n}''=c_{n}\mathrm{e}^{-x^{2}/2}\bigl(x^{2}-1-2n\bigr)H_{n}=(x^{2}-2n-1)\psi_{n}, which rearranges to −ψn′′+x2ψn=(2n+1)ψn-\psi_{n}''+x^{2}\psi_{n}=(2n+1)\psi_{n}. □\square

Remark(Where the hypotheses are used)Original

nn must be a non-negative integer. Riley, Hobson and Bence write Hermite's equation with a real parameter ν\nu and substitute a power series y=∑mamxmy=\sum_{m}a_{m}x^{m}, which forces

am+2=−2(ν−m)(m+1)(m+2) am.a_{m+2}=-\frac{2(\nu-m)}{(m+1)(m+2)}\,a_{m} .

If ν=n\nu=n is a non-negative integer the chain of even or odd coefficients starting below nn terminates at m=nm=n, and that branch is the polynomial HnH_{n}. If ν\nu is not a non-negative integer, no branch terminates, and for large mm the ratio behaves like am+2/am∼2/ma_{m+2}/a_{m}\sim 2/m — which is exactly the ratio of consecutive coefficients of ex2=∑jx2j/j!\mathrm{e}^{x^{2}}=\sum_{j}x^{2j}/j!, whose xmx^{m} coefficient 1/(m/2)!1/(m/2)! satisfies cm+2/cm=1/(m/2+1)∼2/mc_{m+2}/c_{m}=1/(m/2+1)\sim2/m. So the solution grows like ex2\mathrm{e}^{x^{2}} and the Hermite function e−x2/2y\mathrm{e}^{-x^{2}/2}y like ex2/2\mathrm{e}^{x^{2}/2}: it is not in L2(R)L^{2}(\mathbb{R}), and item 6 produces no eigenfunction. This is the whole content of energy quantisation in the application below: integrality of nn is not assumed for convenience, it is forced by square-integrability.

The interval must be the whole line, and the weight must be the Gaussian. Orthogonality in item 5 came from the Gaussian integral over (−∞,∞)(-\infty,\infty); the half-line version ∫0∞e−x2HmHn\int_{0}^{\infty}\mathrm{e}^{-x^{2}}H_{m}H_{n} is not δmn\delta_{mn} times anything, as ∫0∞e−x2H0H1 dx=2∫0∞xe−x2 dx=1≠0\int_{0}^{\infty}\mathrm{e}^{-x^{2}}H_{0}H_{1}\,\mathrm{d}x=2\int_{0}^{\infty}x\mathrm{e}^{-x^{2}}\,\mathrm{d}x=1\neq0 shows at once. In Sturm-Liouville language — Riley, Hobson and Bence note that Hermite's equation has the form with p=ρ=e−x2p=\rho=\mathrm{e}^{-x^{2}}, q=0q=0, λ=2n\lambda=2n — the operator is self-adjoint on this interval precisely because pp decays fast enough to kill the boundary terms at ±∞\pm\infty for all polynomial data (Adjoint and self-adjoint differential operators). Cut the interval at 00 and the boundary term at the cut survives.

Absolute convergence is what licenses the term-by-term work. Two places in the proof interchange a limit with an operation: the rearrangement into the explicit sum, and the term-by-term integration in item 5. The first is safe because both exponential series converge absolutely for all arguments. The second needs a genuine estimate, which is why the bound (∗)(\ast) was recorded; without an integrable dominating function the manipulation would only be formal, which is how the derivation is usually presented.

Completeness is a separate fact. Items 1–6 say nothing about whether the ψn\psi_{n} span L2(R)L^{2}(\mathbb{R}). Orthonormality alone never implies completeness, and here it genuinely needs an argument — Reed and Simon refer it to problems in two later chapters, and Kreyszig states it without proof. A reader who needs only expansions of smooth, rapidly decaying functions can get by with items 1–6; a reader who wants Parseval for all of L2(R)L^{2}(\mathbb{R}) needs item 7.

Remark(Physicists' HnH_{n} and probabilists' HenHe_{n})AdaptedAdapted from Kreyszig (1978), §3.7-2, the Warning following eq. (9), p. 184

Two different families are called the Hermite polynomials, and both are sometimes written HnH_{n}. Kreyszig flags this explicitly, calling the terminology in the literature not unique, and defines the competing family by

He0(x)=1,Hen(x)=(−1)nex2/2dndxne−x2/2,n=1,2,…He_{0}(x)=1, \qquad He_{n}(x)=(-1)^{n}\mathrm{e}^{x^{2}/2}\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\mathrm{e}^{-x^{2}/2}, \qquad n=1,2,\dots

These are the probabilists' polynomials: the weight is the standard normal density up to a constant, e−x2/2\mathrm{e}^{-x^{2}/2}, rather than e−x2\mathrm{e}^{-x^{2}}. Everything in this unit is stated for the physicists' polynomials HnH_{n}, which is the convention of all five of our sources' main development.

The two families are the same objects at different scales. Comparing the Rodrigues formulas,

Hen(x)=2−n/2 Hn ⁣(x2),Hn(x)=2n/2 Hen(2 x),He_{n}(x)=2^{-n/2}\,H_{n}\!\left(\frac{x}{\sqrt{2}}\right), \qquad H_{n}(x)=2^{n/2}\,He_{n}\bigl(\sqrt{2}\,x\bigr),

and correspondingly

Hen+1(x)=x Hen(x)−n Hen−1(x),Hen′(x)=n Hen−1(x),He_{n+1}(x)=x\,He_{n}(x)-n\,He_{n-1}(x), \qquad He_{n}'(x)=n\,He_{n-1}(x),∫−∞∞e−x2/2Hem(x)Hen(x) dx=2π n! δmn.\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}/2}He_{m}(x)He_{n}(x)\,\mathrm{d}x=\sqrt{2\pi}\,n!\,\delta_{mn} .

The first few are He0=1He_{0}=1, He1=xHe_{1}=x, He2=x2−1He_{2}=x^{2}-1, He3=x3−3xHe_{3}=x^{3}-3x, He4=x4−6x2+3He_{4}=x^{4}-6x^{2}+3: note that they are monic, which is the main reason probabilists prefer them. (All three displays were checked symbolically against the definitions.)

How to tell which one you are reading. Look at H1H_{1} or at the coefficient of the highest power: physicists' H1=2xH_{1}=2x and leading coefficient 2n2^{n}; probabilists' He1=xHe_{1}=x and leading coefficient 11. Or look at the weight in the orthogonality relation: e−x2\mathrm{e}^{-x^{2}} against e−x2/2\mathrm{e}^{-x^{2}/2}. Mixing the two silently is the commonest error in this corner of the subject; a formula taken from a statistics text and dropped into an oscillator calculation will be wrong by powers of 2\sqrt2.

Method(Working with Hermite polynomials)AdaptedAdapted from Arfken, Weber & Harris (2013), §18.1, the use of eq. (18.3) to construct $H_n$ from $H_0$ and $H_1$, p. 873; Riley, Hobson & Bence (2006), §18.9.1, the expansion $f=\sum a_n H_n$ and its coefficient formula, p. 627; Hassani (2013), §8.7, Problems 8.7 and 8.13, pp. 259–260

To evaluate HnH_{n} at a point. Do not use the explicit sum: its terms alternate in sign and grow enormously before they cancel, so it loses precision badly for moderate nn. Start from H0(x)=1H_{0}(x)=1, H1(x)=2xH_{1}(x)=2x and iterate

Hk+1(x)=2x Hk(x)−2k Hk−1(x),k=1,…,n−1.H_{k+1}(x)=2x\,H_{k}(x)-2k\,H_{k-1}(x), \qquad k=1,\dots,n-1 .

For ∣x∣|x| of order n\sqrt{n} or larger the values overflow; carry ψk\psi_{k} instead, using the same recurrence rewritten for the normalised functions,

ψk+1(x)=x2k+1 ψk(x)−kk+1 ψk−1(x),ψ0(x)=π−1/4e−x2/2,\psi_{k+1}(x)=x\sqrt{\tfrac{2}{k+1}}\,\psi_{k}(x)-\sqrt{\tfrac{k}{k+1}}\,\psi_{k-1}(x), \qquad \psi_{0}(x)=\pi^{-1/4}\mathrm{e}^{-x^{2}/2},

which keeps every quantity of size O(1)O(1).

To expand a function. Given ff with ∫e−x2f2<∞\int\mathrm{e}^{-x^{2}}f^{2}<\infty:

  1. Compute an=(2nn!π)−1∫−∞∞f(x)Hn(x)e−x2dxa_{n}=\bigl(2^{n}n!\sqrt{\pi}\bigr)^{-1}\int_{-\infty}^{\infty}f(x)H_{n}(x)\mathrm{e}^{-x^{2}}\mathrm{d}x, so that f=∑nanHnf=\sum_{n}a_{n}H_{n} in mean square (item 7).
  2. If ff is even, all odd ana_{n} vanish, and conversely: use the parity relation to halve the work.
  3. If ff is itself a polynomial of degree NN, the expansion terminates at n=Nn=N; the fastest route is then not integration but repeated use of x Hn(x)=12Hn+1(x)+n Hn−1(x),x\,H_{n}(x)=\tfrac12 H_{n+1}(x)+n\,H_{n-1}(x), which is the three-term recurrence solved for xHnxH_{n}. Starting from H0=1H_{0}=1 and multiplying by xx one factor at a time expands xNx^{N} exactly, with rational coefficients and no integrals at all.

To compute moments and matrix elements. Any integral of the form ∫e−x2Hm xk Hn dx\int\mathrm{e}^{-x^{2}}H_{m}\,x^{k}\,H_{n}\,\mathrm{d}x is evaluated by applying x Hn=12Hn+1+nHn−1x\,H_{n}=\tfrac12H_{n+1}+nH_{n-1} exactly kk times to turn xkHnx^{k}H_{n} into a finite combination ∑jcjHj\sum_{j}c_{j}H_{j}, and then reading off the answer as cm⋅2mm!πc_{m}\cdot2^{m}m!\sqrt{\pi} by item 5. Only one coefficient of the combination is ever needed, and the index shifts by at most one per step, so ∫e−x2HmxkHn dx=0\int\mathrm{e}^{-x^{2}}H_{m}x^{k}H_{n}\,\mathrm{d}x=0 whenever ∣m−n∣>k|m-n|>k or m−n+km-n+k is odd.

Example(The spread of the third oscillator state)Original

Take the dimensionless quantum oscillator state ψ2\psi_{2}, that is n=2n=2, and compute ⟨x2⟩=∫x2ψ22\langle x^{2}\rangle=\int x^{2}\psi_{2}^{2} and ⟨x4⟩=∫x4ψ22\langle x^{4}\rangle=\int x^{4}\psi_{2}^{2} using nothing but the recurrence and item 5. Write N2=22⋅2!⋅π=8πN_{2}=2^{2}\cdot2!\cdot\sqrt{\pi}=8\sqrt{\pi} for the norm constant of H2H_{2}, so that ψ22=e−x2H22/N2\psi_{2}^{2}=\mathrm{e}^{-x^{2}}H_{2}^{2}/N_{2}.

Step 1: push the powers of xx through. Apply x Hn=12Hn+1+nHn−1x\,H_{n}=\tfrac12H_{n+1}+nH_{n-1} repeatedly, starting from H2H_{2}:

xH2=12H3+2H1,x2H2=12(12H4+3H2)+2(12H2+H0)=14H4+52H2+2H0,x3H2=14(12H5+4H3)+52(12H3+2H1)+2(12H1)=18H5+94H3+6H1,x4H2=18(12H6+5H4)+94(12H4+3H2)+6(12H2+H0)=116H6+74H4+394H2+6H0.\begin{aligned} xH_{2} &= \tfrac12 H_{3}+2H_{1},\\ x^{2}H_{2} &= \tfrac12\bigl(\tfrac12H_{4}+3H_{2}\bigr)+2\bigl(\tfrac12H_{2}+H_{0}\bigr) = \tfrac14H_{4}+\tfrac52H_{2}+2H_{0},\\ x^{3}H_{2} &= \tfrac14\bigl(\tfrac12H_{5}+4H_{3}\bigr)+\tfrac52\bigl(\tfrac12H_{3}+2H_{1}\bigr)+2\bigl(\tfrac12H_{1}\bigr) = \tfrac18H_{5}+\tfrac94H_{3}+6H_{1},\\ x^{4}H_{2} &= \tfrac18\bigl(\tfrac12H_{6}+5H_{4}\bigr)+\tfrac94\bigl(\tfrac12H_{4}+3H_{2}\bigr)+6\bigl(\tfrac12H_{2}+H_{0}\bigr)\\ &= \tfrac{1}{16}H_{6}+\tfrac74H_{4}+\tfrac{39}{4}H_{2}+6H_{0} . \end{aligned}

Step 2: read off the integrals. By item 5, ∫e−x2H2Hj=0\int\mathrm{e}^{-x^{2}}H_{2}H_{j}=0 unless j=2j=2, when it equals N2=8πN_{2}=8\sqrt{\pi}. So only the coefficient of H2H_{2} in each line survives:

⟨x2⟩=1N2⋅52 N2=52,⟨x4⟩=1N2⋅394 N2=394=9.75.\langle x^{2}\rangle=\frac{1}{N_{2}}\cdot\frac52\,N_{2}=\frac52 , \qquad \langle x^{4}\rangle=\frac{1}{N_{2}}\cdot\frac{39}{4}\,N_{2}=\frac{39}{4}=9.75 .

The odd moments ⟨x⟩\langle x\rangle and ⟨x3⟩\langle x^{3}\rangle vanish, because the expansions of xH2xH_{2} and x3H2x^{3}H_{2} contain only odd-index Hermite polynomials and H2H_{2} is orthogonal to all of them.

Step 3: check. Direct symbolic integration of x2e−x2H2(x)2/(8π)x^{2}\mathrm{e}^{-x^{2}}H_{2}(x)^{2}/(8\sqrt{\pi}) and of x4e−x2H2(x)2/(8π)x^{4}\mathrm{e}^{-x^{2}}H_{2}(x)^{2}/(8\sqrt{\pi}) over R\mathbb{R} returns 5/25/2 and 39/439/4. The general pattern, obtained the same way for n=0,1,2,3n=0,1,2,3, is

⟨x2⟩n=n+12,⟨x4⟩n=34(2n2+2n+1),\langle x^{2}\rangle_{n}=n+\tfrac12 , \qquad \langle x^{4}\rangle_{n}=\tfrac34\bigl(2n^{2}+2n+1\bigr),

giving (12,34)(\tfrac12,\tfrac34), (32,154)(\tfrac32,\tfrac{15}{4}), (52,394)(\tfrac52,\tfrac{39}{4}), (72,754)(\tfrac72,\tfrac{75}{4}) for n=0,1,2,3n=0,1,2,3; all eight numbers were verified symbolically.

What the numbers mean. ⟨x2⟩n=n+12\langle x^{2}\rangle_{n}=n+\tfrac12 is the virial theorem in this problem: in the units of item 6 the energy is 12(−d2/dx2+x2)\tfrac12(-\mathrm{d}^{2}/\mathrm{d}x^{2}+x^{2}) with eigenvalue n+12n+\tfrac12, and the mean potential energy 12⟨x2⟩=12(n+12)\tfrac12\langle x^{2}\rangle=\tfrac12\bigl(n+\tfrac12\bigr) is exactly half of it, so kinetic and potential energies share the total equally. And the kurtosis of the position distribution,

⟨x4⟩⟨x2⟩2=39/425/4=3925=1.56,\frac{\langle x^{4}\rangle}{\langle x^{2}\rangle^{2}}=\frac{39/4}{25/4}=\frac{39}{25}=1.56 ,

is below the Gaussian value 33: the n=2n=2 state, with its two nodes and its probability piled up near the classical turning points, is far flatter-topped than a Gaussian. For n=0n=0 the same ratio is (3/4)/(1/4)=3(3/4)/(1/4)=3, as it must be, since ψ02\psi_{0}^{2} is a Gaussian.

ApplicationOriginal

The quantum harmonic oscillator. This is the reason Hermite polynomials are in a physics curriculum at all. Arfken, Weber and Harris start from −ℏ22mψ′′+k2z2ψ=Eψ-\frac{\hbar^{2}}{2m}\psi''+\frac{k}{2}z^{2}\psi=E\psi and rescale z=b xz=b\,x with

b=ℏ1/2(km)1/4=ℏmω,ω=k/m,b=\frac{\hbar^{1/2}}{(km)^{1/4}}=\sqrt{\frac{\hbar}{m\omega}}, \qquad \omega=\sqrt{k/m} ,

which reduces the equation to −12φ′′+12x2φ=λφ-\tfrac12\varphi''+\tfrac12x^{2}\varphi=\lambda\varphi — item 6 with λ=n+12\lambda=n+\tfrac12. Undoing the substitution,

ψn(z)=(b 2nn!π)−1/2 e−(z/b)2/2 Hn(z/b),En=(n+12)ℏω=(n+12)ℏk/m.\psi_{n}(z)=\bigl(b\,2^{n}n!\sqrt{\pi}\bigr)^{-1/2}\,\mathrm{e}^{-(z/b)^{2}/2}\,H_{n}(z/b), \qquad E_{n}=\bigl(n+\tfrac12\bigr)\hbar\omega=\bigl(n+\tfrac12\bigr)\hbar\sqrt{k/m} .

The argument of HnH_{n} is z/bz/b and not bzbz: the quantity bb is a length, so z/bz/b is the only dimensionless combination available. (Arfken, Weber and Harris print this eigenfunction in eq. (18.19) as Nne−(αz)2/2Hn(αz)N_{n}\mathrm{e}^{-(\alpha z)^{2}/2}H_{n}(\alpha z) with α=ℏ1/2/(km)1/4\alpha=\hbar^{1/2}/(km)^{1/4}, which is our bb rather than 1/b1/b; the argument needs the reciprocal of the α\alpha they print, and their own normalisation constant Nn=[α/(2nπ n!)]1/2N_{n}=[\alpha/(2^{n}\sqrt{\pi}\,n!)]^{1/2} of eq. (18.20) agrees with the reciprocal reading, since it is (b 2nn!π)−1/2(b\,2^{n}n!\sqrt{\pi})^{-1/2} exactly when α=1/b\alpha=1/b. We follow the substitution z=ℏ1/2x/(km)1/4z=\hbar^{1/2}x/(km)^{1/4} that they make on the preceding page, which fixes the reading unambiguously. That the ψn\psi_{n} displayed above solve the dimensional equation with the energies quoted, and are normalised, was checked symbolically for n≤3n\le3.)

Two features are worth isolating. The spectrum is equally spaced, which is what makes the oscillator the model for a mode of a field and the nn a photon or phonon number; and the lowest energy is 12ℏω\tfrac12\hbar\omega, not zero, the zero-point energy, which is forced by the fact that a normalisable wave function has finite spatial extent and so cannot have zero kinetic energy. Reed and Simon's route — raising from the ground state with A†A^{\dagger}, which is the ladder-operator argument Arfken, Weber and Harris also give — produces the same functions and explains the equal spacing structurally. In the dimensionless variables of item 6, where H=12(−d2/dx2+x2)=N+12H=\tfrac12\bigl(-\mathrm{d}^{2}/\mathrm{d}x^{2}+x^{2}\bigr)=N+\tfrac12 with N=A†AN=A^{\dagger}A, the commutator is [H,A†]=[N,A†]=A†[H,A^{\dagger}]=[N,A^{\dagger}]=A^{\dagger}: applying A†A^{\dagger} moves one rung up, of height 11 in λ\lambda. Restoring units multiplies the ladder spacing by ℏω\hbar\omega, so for the dimensional Hamiltonian the same relation reads [H,A†]=ℏωA†[H,A^{\dagger}]=\hbar\omega A^{\dagger}. The ladder has a bottom because ∥Aψ∥2≥0\|A\psi\|^{2}\ge0.

Molecular vibration and infrared spectra. Near the bottom of a diatomic potential well the interatomic potential is quadratic to leading order, so the vibrational levels start out equally spaced and the wave functions are Hermite functions in the bond-length displacement. The selection rule Δn=±1\Delta n=\pm1 for infrared absorption is the statement that ∫e−x2Hm x Hn dx\int\mathrm{e}^{-x^{2}}H_{m}\,x\,H_{n}\,\mathrm{d}x vanishes unless ∣m−n∣=1|m-n|=1, which is the k=1k=1 case of the matrix-element rule in the method above. Anharmonicity — the cubic and quartic corrections to the potential — is then handled by perturbation theory (Regular perturbation expansions) with exactly the matrix elements the method computes, and it is what makes real vibrational ladders converge rather than stay equally spaced.

Hermite-Gauss beams in optics. The paraxial wave equation in a rectangular geometry separates into two copies of the oscillator equation, so the transverse modes of a laser cavity with rectangular symmetry are products ψm(x/w)ψn(y/w)\psi_{m}(x/w)\psi_{n}(y/w) of Hermite functions, labelled TEMmn\mathrm{TEM}_{mn}. The mm and nn nodal lines are directly visible in a beam profile, and the orthogonality of item 5 is why the modes of a cavity do not exchange power.

Gauss-Hermite quadrature. Because HnH_{n} has nn real simple zeros (General properties of orthogonal polynomials), there are weights w1,…,wnw_{1},\dots,w_{n} making

∫−∞∞e−x2f(x) dx≈∑i=1nwif(xi)\int_{-\infty}^{\infty}\mathrm{e}^{-x^{2}}f(x)\,\mathrm{d}x\approx\sum_{i=1}^{n}w_{i}f(x_{i})

exact for every polynomial ff of degree at most 2n−12n-1 (Numerical quadrature). This is the standard way to integrate against a Gaussian in any dimension where the integrand is smooth — expectation values under a normal distribution, thermal averages, and the moment-matching step of a Gaussian-quadrature filter in data assimilation (Optimal interpolation and variational data assimilation).

Series in the probabilists' polynomials. With the HenHe_{n} of the convention remark, the Gram-Charlier and Edgeworth expansions write a near-Gaussian probability density as ϕ(x)∑ncnHen(x)\phi(x)\sum_{n}c_{n}He_{n}(x) with ϕ\phi the standard normal density, the coefficients being combinations of the cumulants: c3c_{3} carries the skewness, c4c_{4} the excess kurtosis. These are the standard corrections to the central limit theorem (The central limit theorem) at finite sample size, and the reason the probabilists' normalisation exists at all: with the weight equal to the normal density, cn=E[Hen(X)]/n!c_{n}=\mathbb{E}[He_{n}(X)]/n! with no stray powers of two.

Definition

Laguerre polynomials

Adapted

The polynomials orthogonal on the half-line with weight xke−xx^{k}\mathrm{e}^{-x}: their Rodrigues formula and generating function, the associated Laguerre equation they solve, the two conventions in circulation, and the hydrogen radial functions they build.

Prerequisites: General properties of orthogonal polynomials, The Gamma function, The Frobenius method

Definition(Laguerre and associated Laguerre polynomials)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.7, eq. (18.112), p. 618, and §18.8, eqs. (18.118)–(18.121), pp. 621–622; Arfken, Weber & Harris (2013), §18.3, eqs. (18.44), (18.46) and (18.58), (18.69), (18.72), pp. 889–895; Hassani (2013), §8.4.2, eq. (8.16) and eq. (8.18), pp. 249–250, and Problem 8.32, p. 263; Kreyszig (1978), §3.7-3, eqs. (10b)–(10c), p. 184

Fix an integer k≥0k\ge0. The associated Laguerre polynomials L0k,L1k,L2k,…L^{k}_{0},L^{k}_{1},L^{k}_{2},\dots are defined by the Rodrigues formula

Lnk(x)  =  ex x−kn! dndxn(x n+k e−x),x>0.L^{k}_{n}(x) \;=\; \frac{\mathrm{e}^{x}\,x^{-k}}{n!}\,\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\Bigl(x^{\,n+k}\,\mathrm{e}^{-x}\Bigr), \qquad x>0 .

The case k=0k=0 gives the (ordinary) Laguerre polynomials, written Ln=Ln0L_{n}=L^{0}_{n}:

Ln(x)=exn!dndxn(xne−x).L_{n}(x)=\frac{\mathrm{e}^{x}}{n!}\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\bigl(x^{n}\mathrm{e}^{-x}\bigr).

This is the convention of all four of our sources, and it is the one in which Ln(0)=1L_{n}(0)=1. In the language of General properties of orthogonal polynomials, LnkL^{k}_{n} is the classical family belonging to the weight

wk(x)=xke−xon(0,∞),w_{k}(x)=x^{k}\mathrm{e}^{-x} \quad\text{on}\quad (0,\infty),

with the scale fixed by the choice Hassani writes as Kn=n!K_{n}=n!. Hassani runs the whole construction with a real parameter ν>−1\nu>-1 in place of the integer kk, writing LnνL^{\nu}_{n}. The only item below that genuinely needs kk to be an integer is the equivalent definition in item 2, which differentiates kk times; everything else survives with each factorial in kk read through the gamma function (The Gamma function), (n+k)!↦Γ(n+ν+1)(n+k)!\mapsto\Gamma(n+\nu+1) and (k+m)!↦Γ(ν+m+1)(k+m)!\mapsto\Gamma(\nu+m+1), so that the explicit sum of item 1 becomes a series in Γ\Gamma's and its value at the origin becomes Lnν(0)=Γ(n+ν+1)/[n! Γ(ν+1)]L^{\nu}_{n}(0)=\Gamma(n+\nu+1)/[n!\,\Gamma(\nu+1)] (Hassani, Problem 8.32), just as item 6 already writes the norm as Γ(n+ν+1)/n!\Gamma(n+\nu+1)/n!. We note below where the real case differs.

The first few, for general kk, are

L0k=1,L1k=k+1−x,L2k=12[x2−2(k+2)x+(k+1)(k+2)],L^{k}_{0}=1,\qquad L^{k}_{1}=k+1-x,\qquad L^{k}_{2}=\tfrac12\bigl[x^{2}-2(k+2)x+(k+1)(k+2)\bigr],

and for k=0k=0,

L0=1,L1=1−x,L2=12(x2−4x+2),L3=16(−x3+9x2−18x+6).L_{0}=1,\quad L_{1}=1-x,\quad L_{2}=\tfrac12\bigl(x^{2}-4x+2\bigr),\quad L_{3}=\tfrac16\bigl(-x^{3}+9x^{2}-18x+6\bigr).

Note the sign: LnL_{n} starts at +1+1 and its leading term is (−1)nxn/n!(-1)^{n}x^{n}/n!, so these polynomials are not monic and alternate in overall sign at large xx.

The Laguerre functions φnk(x)=e−x/2xk/2Lnk(x)\varphi^{k}_{n}(x)=\mathrm{e}^{-x/2}x^{k/2}L^{k}_{n}(x) absorb half the weight, so that they are orthogonal with unit weight; Kreyszig's orthonormal sequence in L2[0,∞)L^{2}[0,\infty) is the case k=0k=0, namely e−x/2Ln(x)\mathrm{e}^{-x/2}L_{n}(x), which he obtains by Gram-Schmidt from xne−x/2x^{n}\mathrm{e}^{-x/2} (Orthonormal bases and Gram-Schmidt).

Definition(What the Laguerre polynomials satisfy)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.7, eqs. (18.107) and (18.111), p. 617, §18.7.1, eqs. (18.113)–(18.117), pp. 619–621, and §18.8–18.8.1, eqs. (18.118)–(18.125), pp. 621–624; Arfken, Weber & Harris (2013), §18.3, eqs. (18.49), (18.51)–(18.55), (18.59), (18.61)–(18.64), (18.66), (18.69), (18.71), (18.73), pp. 890–895; Hassani (2013), §8.4.2, the recurrence and eq. (8.18) and the norm $h_n=\Gamma(n+\nu+1)/n!$, pp. 249–250; Kreyszig (1978), §3.7-3, eq. (11), p. 185

Let k≥0k\ge0 be an integer and LnkL^{k}_{n} as defined above. Write L−1k≡0L^{k}_{-1}\equiv0.

  1. Explicit form. For n≥0n\ge0, Lnk(x)=∑m=0n(−1)m (n+k)!(n−m)! (k+m)! m!  xm.L^{k}_{n}(x)=\sum_{m=0}^{n}(-1)^{m}\,\frac{(n+k)!}{(n-m)!\,(k+m)!\,m!}\;x^{m} . So LnkL^{k}_{n} is a polynomial of degree exactly nn with leading coefficient (−1)n/n!(-1)^{n}/n!, and Lnk(0)=(n+kn)=(n+k)!n! k!,in particular Ln(0)=1.L^{k}_{n}(0)=\binom{n+k}{n}=\frac{(n+k)!}{n!\,k!}, \qquad\text{in particular } L_{n}(0)=1 . Unlike the Legendre, Hermite and Chebyshev families, the LnkL^{k}_{n} have no parity: the interval (0,∞)(0,\infty) is not symmetric.
  2. Relation to the ordinary polynomials. Lnk(x)=(−1)kdkdxkLn+k(x)\displaystyle L^{k}_{n}(x)=(-1)^{k}\frac{\mathrm{d}^{k}}{\mathrm{d}x^{k}}L_{n+k}(x). This is the formula Riley, Hobson and Bence and Arfken, Weber and Harris take as their definition of the associated polynomials; here it is a consequence of the Rodrigues formula. It requires kk to be a non-negative integer, and is the one item of this statement that does not survive to Hassani's real ν\nu.
  3. Generating function. For ∣h∣<1|h|<1 and all xx, e−xh/(1−h)(1−h)k+1=∑n=0∞Lnk(x) hn.\frac{\mathrm{e}^{-xh/(1-h)}}{(1-h)^{k+1}}=\sum_{n=0}^{\infty}L^{k}_{n}(x)\,h^{n} .
  4. Recurrences. For n≥1n\ge1 and k≥0k\ge0, (n+1)Ln+1k(x)=(2n+k+1−x) Lnk(x)−(n+k) Ln−1k(x),(n+1)L^{k}_{n+1}(x)=(2n+k+1-x)\,L^{k}_{n}(x)-(n+k)\,L^{k}_{n-1}(x), x (Lnk)′(x)=n Lnk(x)−(n+k) Ln−1k(x),(Lnk)′(x)=−Ln−1k+1(x),x\,\bigl(L^{k}_{n}\bigr)'(x)=n\,L^{k}_{n}(x)-(n+k)\,L^{k}_{n-1}(x), \qquad \bigl(L^{k}_{n}\bigr)'(x)=-L^{k+1}_{n-1}(x), and, for every n≥0n\ge0, Lnk+1=∑j=0nLjk.L^{k+1}_{n}=\sum_{j=0}^{n}L^{k}_{j} . The last two relations move between neighbouring values of the upper index; note that differentiating raises kk and lowers nn. (With the convention L−1k≡0L^{k}_{-1}\equiv0 the first three relations hold at n=0n=0 as well; the fourth needs no convention.)
  5. The associated Laguerre equation. y=Lnky=L^{k}_{n} solves x y′′+(k+1−x) y′+n y=0,x\,y''+(k+1-x)\,y'+n\,y=0 , which for k=0k=0 is Laguerre's equation xy′′+(1−x)y′+ny=0xy''+(1-x)y'+ny=0. The point x=0x=0 is a regular singular point and x=∞x=\infty an irregular one.
  6. Orthogonality and norm. For all m,n≥0m,n\ge0, ∫0∞e−xxk Lmk(x) Lnk(x) dx=(n+k)!n!  δmn.\int_{0}^{\infty}\mathrm{e}^{-x}x^{k}\,L^{k}_{m}(x)\,L^{k}_{n}(x)\,\mathrm{d}x=\frac{(n+k)!}{n!}\;\delta_{mn} . For k=0k=0 the right-hand side is simply δmn\delta_{mn}: the ordinary Laguerre polynomials are already orthonormal for the weight e−x\mathrm{e}^{-x}, with no normalising constant at all. In Hassani's real-parameter version the constant is Γ(n+ν+1)/n!\Gamma(n+\nu+1)/n!, which reduces to (n+k)!/n!(n+k)!/n! at ν=k\nu=k (The Gamma function).
  7. Self-adjoint form and completeness. φnk=e−x/2xk/2Lnk\varphi^{k}_{n}=\mathrm{e}^{-x/2}x^{k/2}L^{k}_{n} satisfies x (φnk)′′+(φnk)′+(−x4+2n+k+12−k24x)φnk=0,x\,\bigl(\varphi^{k}_{n}\bigr)''+\bigl(\varphi^{k}_{n}\bigr)'+\Bigl(-\frac{x}{4}+\frac{2n+k+1}{2}-\frac{k^{2}}{4x}\Bigr)\varphi^{k}_{n}=0 , and the closely related Φnk=e−x/2x(k+1)/2Lnk\Phi^{k}_{n}=\mathrm{e}^{-x/2}x^{(k+1)/2}L^{k}_{n} satisfies the equation with no first-derivative term, (Φnk)′′+(−14+2n+k+12x−k2−14x2)Φnk=0,\bigl(\Phi^{k}_{n}\bigr)''+\Bigl(-\frac14+\frac{2n+k+1}{2x}-\frac{k^{2}-1}{4x^{2}}\Bigr)\Phi^{k}_{n}=0 , which is the form in which the hydrogen radial equation appears. For k=0k=0, Kreyszig states that the normalised functions e−x/2Ln(x)\mathrm{e}^{-x/2}L_{n}(x) form a total orthonormal sequence in L2[0,∞)L^{2}[0,\infty); we record that on his authority and do not prove it. Items 1–6 are proved below.
Idea of the proofOriginal

Every identity here is squeezed out of the Rodrigues formula by one of two moves.

The first move is Leibniz's rule. Differentiating xn+ke−xx^{n+k}\mathrm{e}^{-x} exactly nn times produces a sum over how many of the derivatives landed on the power and how many on the exponential; dividing out the xke−xx^{k}\mathrm{e}^{-x} that was never differentiated leaves a polynomial of degree nn, written down term by term. That is item 1, and the leading coefficient, the value at the origin and the generating function all follow from the series by elementary resummation.

The second move is to differentiate a first-order identity many times. The function v=xn+ke−xv=x^{n+k}\mathrm{e}^{-x} satisfies xv′=(n+k−x)vxv'=(n+k-x)v — one line, because vv is a power times an exponential. Differentiating that identity n+1n+1 times with Leibniz's rule relates v(n)v^{(n)}, v(n+1)v^{(n+1)} and v(n+2)v^{(n+2)}, and since v(n)v^{(n)} is n! xke−xLnkn!\,x^{k}\mathrm{e}^{-x}L^{k}_{n} by definition, the relation is the associated Laguerre equation in disguise. No power series and no appeal to the general classical theory are needed.

Orthogonality then needs nothing new: put the Rodrigues formula for the higher-index polynomial inside the integral and integrate by parts nn times. The boundary terms vanish because what is being differentiated is a power at least xkx^{k} at the origin and an exponential at infinity; and after nn parts the other factor has been differentiated nn times, which annihilates it if its degree is smaller. The one surviving case is m=nm=n, where the nnth derivative of a degree-nn polynomial is a constant and the integral collapses to Γ(n+k+1)\Gamma(n+k+1).

ProofAdaptedAdapted from Riley, Hobson & Bence (2006), §18.7.1, the Leibniz evaluation of the Rodrigues formula, p. 619, the integration-by-parts evaluation of the normalisation, p. 619 and p. 623, and §18.8, the Leibniz derivation of the associated equation, p. 621; Arfken, Weber & Harris (2013), §18.3, the extraction of the recurrences from the generating function, p. 890, and the relations (18.61)–(18.63), p. 893

Item 1. Apply Leibniz's rule to dn/dxn(xn+ke−x)\mathrm{d}^{n}/\mathrm{d}x^{n}\bigl(x^{n+k}\mathrm{e}^{-x}\bigr), letting rr derivatives fall on the power and n−rn-r on the exponential:

dndxn(xn+ke−x)=∑r=0n(nr)(n+k)!(n+k−r)! x n+k−r (−1)n−re−x.\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\bigl(x^{n+k}\mathrm{e}^{-x}\bigr) =\sum_{r=0}^{n}\binom{n}{r}\frac{(n+k)!}{(n+k-r)!}\,x^{\,n+k-r}\,(-1)^{n-r}\mathrm{e}^{-x} .

Multiply by exx−k/n!\mathrm{e}^{x}x^{-k}/n! and substitute m=n−rm=n-r, so that xn+k−r⋅x−k=xmx^{n+k-r}\cdot x^{-k}=x^{m}, (nr)=n!(n−m)! m!\binom{n}{r}=\frac{n!}{(n-m)!\,m!} after the substitution, and (n+k−r)!=(k+m)!(n+k-r)!=(k+m)!:

Lnk(x)=∑m=0n(−1)m(n+k)!(n−m)! (k+m)! m! xm.L^{k}_{n}(x)=\sum_{m=0}^{n}(-1)^{m}\frac{(n+k)!}{(n-m)!\,(k+m)!\,m!}\,x^{m} .

The m=nm=n term is (−1)n(n+k)! xn/(0! (k+n)! n!)=(−1)nxn/n!(-1)^{n}(n+k)!\,x^{n}/\bigl(0!\,(k+n)!\,n!\bigr)=(-1)^{n}x^{n}/n!, which is the leading term and is non-zero, so the degree is exactly nn. The m=0m=0 term is (n+k)!/(n! k!)(n+k)!/(n!\,k!), which is the value at the origin.

Item 3. Insert the series and exchange the two finite-or-absolutely-convergent sums:

∑n≥0Lnk(x)hn=∑m≥0(−1)mxmm! (k+m)!∑n≥m(n+k)!(n−m)! hn.\sum_{n\ge0}L^{k}_{n}(x)h^{n} =\sum_{m\ge0}\frac{(-1)^{m}x^{m}}{m!\,(k+m)!}\sum_{n\ge m}\frac{(n+k)!}{(n-m)!}\,h^{n} .

Put n=m+jn=m+j in the inner sum and factor out hmh^{m}:

∑j≥0(m+j+k)!j!hm+j=(m+k)! hm∑j≥0(j+m+kj)hj=(m+k)! hm(1−h)m+k+1,\sum_{j\ge0}\frac{(m+j+k)!}{j!}h^{m+j} =(m+k)!\,h^{m}\sum_{j\ge0}\binom{j+m+k}{j}h^{j} =(m+k)!\,\frac{h^{m}}{(1-h)^{m+k+1}} ,

by the binomial series ∑j(j+pj)hj=(1−h)−(p+1)\sum_{j}\binom{j+p}{j}h^{j}=(1-h)^{-(p+1)}, valid for ∣h∣<1|h|<1. The factor (m+k)!(m+k)! cancels the one in the denominator, leaving

∑n≥0Lnk(x)hn=1(1−h)k+1∑m≥01m!(−xh1−h)m=e−xh/(1−h)(1−h)k+1.\sum_{n\ge0}L^{k}_{n}(x)h^{n}=\frac{1}{(1-h)^{k+1}}\sum_{m\ge0}\frac{1}{m!}\left(\frac{-xh}{1-h}\right)^{m} =\frac{\mathrm{e}^{-xh/(1-h)}}{(1-h)^{k+1}} .

Both rearrangements are legitimate for ∣h∣<1|h|<1 and fixed xx, since all the sums converge absolutely there.

Item 4. Write G=Gk(x,h)G=G_{k}(x,h) for the generating function. Differentiating in hh,

∂G∂h=[k+11−h−x(1−h)2]G,so(1−h)2∂G∂h=[(k+1)(1−h)−x]G.\frac{\partial G}{\partial h}=\left[\frac{k+1}{1-h}-\frac{x}{(1-h)^{2}}\right]G , \qquad\text{so}\qquad (1-h)^{2}\frac{\partial G}{\partial h}=\bigl[(k+1)(1-h)-x\bigr]G .

Insert G=∑nLnkhnG=\sum_{n}L^{k}_{n}h^{n} and collect the coefficient of hnh^{n} on each side. On the left, (1−2h+h2)∑nnLnkhn−1(1-2h+h^{2})\sum_{n}nL^{k}_{n}h^{n-1} gives (n+1)Ln+1k−2nLnk+(n−1)Ln−1k(n+1)L^{k}_{n+1}-2nL^{k}_{n}+(n-1)L^{k}_{n-1}; on the right, (k+1−x)Lnk−(k+1)Ln−1k(k+1-x)L^{k}_{n}-(k+1)L^{k}_{n-1}. Equating and moving all but the first term to the right,

(n+1)Ln+1k=(2n+k+1−x)Lnk−(n+k)Ln−1k,(n+1)L^{k}_{n+1}=(2n+k+1-x)L^{k}_{n}-(n+k)L^{k}_{n-1},

where the two coefficients were assembled as 2nLnk+(k+1−x)Lnk=(2n+k+1−x)Lnk2nL^{k}_{n}+(k+1-x)L^{k}_{n}=(2n+k+1-x)L^{k}_{n} and −(n−1)Ln−1k−(k+1)Ln−1k=−(n+k)Ln−1k-(n-1)L^{k}_{n-1}-(k+1)L^{k}_{n-1}=-(n+k)L^{k}_{n-1}.

Differentiating in xx instead gives ∂G/∂x=−h1−hG\partial G/\partial x=-\frac{h}{1-h}G, that is (1−h)∂xG=−hG(1-h)\partial_{x}G=-hG, whose hnh^{n} coefficient is

(Lnk)′−(Ln−1k)′=−Ln−1k.\bigl(L^{k}_{n}\bigr)'-\bigl(L^{k}_{n-1}\bigr)'=-L^{k}_{n-1} .

Summing this from index 11 to nn telescopes, and (L0k)′=0\bigl(L^{k}_{0}\bigr)'=0, so

(Lnk)′=−∑i=0n−1Lik.\bigl(L^{k}_{n}\bigr)'=-\sum_{i=0}^{n-1}L^{k}_{i} .

Finally, comparing the generating functions for two neighbouring upper indices, Gk+1=Gk/(1−h)G_{k+1}=G_{k}/(1-h), and expanding 1/(1−h)=∑jhj1/(1-h)=\sum_{j}h^{j}, the hnh^{n} coefficient of the product is Lnk+1=∑j=0nLjkL^{k+1}_{n}=\sum_{j=0}^{n}L^{k}_{j}. Combining the last two displays gives (Lnk)′=−Ln−1k+1\bigl(L^{k}_{n}\bigr)'=-L^{k+1}_{n-1}.

For the remaining recurrence, differentiate the series of item 1 term by term and multiply by xx:

x(Lnk)′=∑m=1n(−1)m(n+k)! m(n−m)! (k+m)! m!xm=n Lnk−(n+k)Ln−1k,x\bigl(L^{k}_{n}\bigr)'=\sum_{m=1}^{n}(-1)^{m}\frac{(n+k)!\,m}{(n-m)!\,(k+m)!\,m!}x^{m} =n\,L^{k}_{n}-(n+k)L^{k}_{n-1},

where the last equality is checked coefficient by coefficient: the coefficient of xmx^{m} on the right is

(−1)mn (n+k)!(n−m)!(k+m)!m!−(−1)m(n+k)(n−1+k)!(n−1−m)!(k+m)!m!=(−1)m(n+k)!(k+m)! m![n(n−m)!−1(n−1−m)!],(-1)^{m}\frac{n\,(n+k)!}{(n-m)!(k+m)!m!}-(-1)^{m}\frac{(n+k)(n-1+k)!}{(n-1-m)!(k+m)!m!} =(-1)^{m}\frac{(n+k)!}{(k+m)!\,m!}\left[\frac{n}{(n-m)!}-\frac{1}{(n-1-m)!}\right],

and n(n−m)!−1(n−1−m)!=n−(n−m)(n−m)!=m(n−m)!\frac{n}{(n-m)!}-\frac{1}{(n-1-m)!}=\frac{n-(n-m)}{(n-m)!}=\frac{m}{(n-m)!}, which is the coefficient on the left.

Item 2. Iterating (Lnk)′=−Ln−1k+1\bigl(L^{k}_{n}\bigr)'=-L^{k+1}_{n-1} a total of kk times starting from Ln+k0=Ln+kL^{0}_{n+k}=L_{n+k} gives

dkdxkLn+k=(−1)kLnk,\frac{\mathrm{d}^{k}}{\mathrm{d}x^{k}}L_{n+k}=(-1)^{k}L^{k}_{n},

which is the stated relation.

Item 5. Put v(x)=x n+ke−xv(x)=x^{\,n+k}\mathrm{e}^{-x}, so that v(n)=n! xke−xLnkv^{(n)}=n!\,x^{k}\mathrm{e}^{-x}L^{k}_{n} by the Rodrigues formula. Because vv is a power times an exponential it satisfies the first-order identity

x v′=(n+k−x) v.x\,v'=(n+k-x)\,v .

Differentiate this n+1n+1 times with Leibniz's rule. On the left, only the first two Leibniz terms survive because xx is linear:

(xv′)(n+1)=x v(n+2)+(n+1) v(n+1).\bigl(xv'\bigr)^{(n+1)}=x\,v^{(n+2)}+(n+1)\,v^{(n+1)} .

On the right, likewise,

((n+k−x)v)(n+1)=(n+k−x) v(n+1)−(n+1) v(n).\bigl((n+k-x)v\bigr)^{(n+1)}=(n+k-x)\,v^{(n+1)}-(n+1)\,v^{(n)} .

Equating and collecting,

x v(n+2)+(1−k+x) v(n+1)+(n+1) v(n)=0.(†)x\,v^{(n+2)}+(1-k+x)\,v^{(n+1)}+(n+1)\,v^{(n)}=0 . \tag{$\dagger$}

Now write y=Lnky=L^{k}_{n} and express the three derivatives in terms of yy. From v(n)=n! xke−xyv^{(n)}=n!\,x^{k}\mathrm{e}^{-x}y,

v(n+1)=n! xk−1e−x[xy′+(k−x)y],v^{(n+1)}=n!\,x^{k-1}\mathrm{e}^{-x}\bigl[x y'+(k-x)y\bigr],v(n+2)=n! xk−2e−x{x[xy′′+(k+1−x)y′−y]+(k−1−x)[xy′+(k−x)y]},v^{(n+2)}=n!\,x^{k-2}\mathrm{e}^{-x}\Bigl\{x\bigl[xy''+(k+1-x)y'-y\bigr]+(k-1-x)\bigl[xy'+(k-x)y\bigr]\Bigr\} ,

each obtained by one differentiation and collection of terms. Substitute these into (†)(\dagger) and divide through by the common factor n! xk−1e−xn!\,x^{k-1}\mathrm{e}^{-x}:

x[xy′′+(k+1−x)y′−y]+(k−1−x)[xy′+(k−x)y]+(1−k+x)[xy′+(k−x)y]+(n+1)xy=0.x\bigl[xy''+(k+1-x)y'-y\bigr]+(k-1-x)\bigl[xy'+(k-x)y\bigr]+(1-k+x)\bigl[xy'+(k-x)y\bigr]+(n+1)xy=0 .

The second and third brackets carry the coefficients (k−1−x)(k-1-x) and (1−k+x)(1-k+x), which sum to zero, so those two terms cancel exactly. What is left is

x[xy′′+(k+1−x)y′−y]+(n+1)xy=0,x\bigl[xy''+(k+1-x)y'-y\bigr]+(n+1)xy=0 ,

and dividing by xx (legitimate for x>0x>0, and the conclusion extends to x=0x=0 by continuity, both sides being polynomials) gives

x y′′+(k+1−x) y′+n y=0.x\,y''+(k+1-x)\,y'+n\,y=0 .

Item 6. Let m≤nm\le n and substitute the Rodrigues formula for the factor of index nn:

I=∫0∞e−xxkLmkLnk dx=1n!∫0∞Lmk(x) dndxn(x n+ke−x) dx.I=\int_{0}^{\infty}\mathrm{e}^{-x}x^{k}L^{k}_{m}L^{k}_{n}\,\mathrm{d}x =\frac{1}{n!}\int_{0}^{\infty}L^{k}_{m}(x)\,\frac{\mathrm{d}^{n}}{\mathrm{d}x^{n}}\bigl(x^{\,n+k}\mathrm{e}^{-x}\bigr)\,\mathrm{d}x .

Integrate by parts nn times, each time moving one derivative off the Rodrigues factor and onto LmkL^{k}_{m}. The boundary term at step jj is

[(Lmk)(j)(x)  d n−1−jdx n−1−j(x n+ke−x)]0∞,j=0,…,n−1.\Bigl[\bigl(L^{k}_{m}\bigr)^{(j)}(x)\;\frac{\mathrm{d}^{\,n-1-j}}{\mathrm{d}x^{\,n-1-j}}\bigl(x^{\,n+k}\mathrm{e}^{-x}\bigr)\Bigr]_{0}^{\infty}, \qquad j=0,\dots,n-1 .

It vanishes at ∞\infty because every derivative of xn+ke−xx^{n+k}\mathrm{e}^{-x} is a polynomial times e−x\mathrm{e}^{-x}; and at 00 because differentiating xn+ke−xx^{n+k}\mathrm{e}^{-x} fewer than n+kn+k times leaves a factor xn+k−(n−1−j)=xk+1+jx^{n+k-(n-1-j)}=x^{k+1+j}, which is a positive power of xx since k≥0k\ge0 and j≥0j\ge0. Hence

I=(−1)nn!∫0∞(Lmk)(n)(x)  x n+ke−x dx.I=\frac{(-1)^{n}}{n!}\int_{0}^{\infty}\bigl(L^{k}_{m}\bigr)^{(n)}(x)\;x^{\,n+k}\mathrm{e}^{-x}\,\mathrm{d}x .

If m<nm<n the nnth derivative of a polynomial of degree mm is identically zero, so I=0I=0; by symmetry of II in mm and nn this covers every m≠nm\neq n. If m=nm=n, the nnth derivative of LnkL^{k}_{n} is n!n! times its leading coefficient, namely n!⋅(−1)n/n!=(−1)nn!\cdot(-1)^{n}/n!=(-1)^{n} by item 1, a constant. Therefore

I=(−1)nn! (−1)n∫0∞x n+ke−x dx=Γ(n+k+1)n!=(n+k)!n!,I=\frac{(-1)^{n}}{n!}\,(-1)^{n}\int_{0}^{\infty}x^{\,n+k}\mathrm{e}^{-x}\,\mathrm{d}x=\frac{\Gamma(n+k+1)}{n!}=\frac{(n+k)!}{n!},

using the integral definition of the gamma function (The Gamma function). □\square

Remark(Where the hypotheses are used)Original

nn must be a non-negative integer, and this is what quantises hydrogen. Riley, Hobson and Bence write Laguerre's equation with a real parameter ν\nu and run Frobenius about the regular singular point x=0x=0 (The Frobenius method). The indicial equation is σ2=0\sigma^{2}=0, a repeated root, so there is only one Frobenius solution; it is an ordinary power series with

am+1=m−ν(m+1)2 am,a_{m+1}=\frac{m-\nu}{(m+1)^{2}}\,a_{m} ,

which terminates exactly when ν\nu is a non-negative integer. If it does not terminate, am+1/am→1/ma_{m+1}/a_{m}\to 1/m for large mm, which is the coefficient ratio of ex\mathrm{e}^{x}, and the solution grows like ex\mathrm{e}^{x}. Arfken, Weber and Harris record the same fact for the associated equation — its non-polynomial solutions diverge proportionally to xkexx^{k}\mathrm{e}^{x} — and then use precisely this to quantise the hydrogen spectrum: the radial function contains the factor e−ρ/2Lλ−L−12L+1(ρ)\mathrm{e}^{-\rho/2}L^{2L+1}_{\lambda-L-1}(\rho), and only when λ−L−1\lambda-L-1 is a non-negative integer does e−ρ/2\mathrm{e}^{-\rho/2} beat the growth and leave a bound state. Integrality of the principal quantum number is therefore not an assumption; it is the boundary condition at infinity.

The second solution is thrown away at the origin, not at infinity. Because the indicial roots coincide, the second solution of Laguerre's equation contains a log⁡x\log x and blows up at the origin. It is excluded by requiring regularity at x=0x=0. The two exclusions are independent: integrality handles x=∞x=\infty, regularity handles x=0x=0, and a Sturm-Liouville problem on (0,∞)(0,\infty) needs a condition at each end.

k>−1k>-1, not merely k≥0k\ge0. Nothing in item 6 needs kk to be an integer; what it needs is that xkx^{k} be integrable near the origin and that the boundary terms vanish there. Both fail at k≤−1k\le-1: already ∫01x−1e−xdx=∞\int_{0}^{1}x^{-1}\mathrm{e}^{-x}\mathrm{d}x=\infty, so the inner product itself is undefined. This is exactly Hassani's standing restriction ν>−1\nu>-1 in his Proposition 8.2.1(b), and it is why the natural parameter range is k>−1k>-1 rather than k≥0k\ge0. Integrality of kk is needed only for item 2, which expresses LnkL^{k}_{n} as a kk-fold derivative. The factorials (n+k)!(n+k)! and (k+m)!(k+m)! of item 1, and the binomial coefficient (n+kn)\binom{n+k}{n} with them, are not exceptions: for real ν>−1\nu>-1 they are read as Γ(n+ν+1)\Gamma(n+\nu+1) and Γ(ν+m+1)\Gamma(\nu+m+1), and the value at the origin becomes Γ(n+ν+1)/[n! Γ(ν+1)]\Gamma(n+\nu+1)/[n!\,\Gamma(\nu+1)], which is how Hassani states it; item 6 is already written that way.

A different weight for every kk. The orthogonality in item 6 holds within a fixed upper index. There is no sense in which L32L^{2}_{3} and L35L^{5}_{3} are orthogonal: they live in different weighted spaces. Arfken, Weber and Harris make the point explicitly — for each kk the associated Laguerre polynomials form a new orthogonal set. In the hydrogen application this is the reason that radial functions with different LL are not orthogonal through the radial integral alone; their orthogonality comes from the angular factor (Spherical harmonics).

Where k≥0k\ge0 enters the boundary terms. In the proof of item 6 the surviving power at the origin was xk+1+jx^{k+1+j} with j≥0j\ge0. For −1<k<0-1<k<0 this is still a positive power, so the argument survives; for k≤−1k\le-1 the j=0j=0 boundary term need not vanish, and the integration by parts breaks down at the first step.

Remark(Two conventions for the associated polynomials)AdaptedAdapted from Arfken, Weber & Harris (2013), §18.3, footnote 4 to eq. (18.58), p. 892; Riley, Hobson & Bence (2006), §18.8, the footnote to eq. (18.119), p. 621

Both Arfken, Weber and Harris and Riley, Hobson and Bence pause to warn that a second convention is in circulation, and they describe the same clash. Writing L\mathcal{L} for the other family, those authors set

Lnk(x)=dkdxkLn(x),\mathcal{L}^{k}_{n}(x)=\frac{\mathrm{d}^{k}}{\mathrm{d}x^{k}}L_{n}(x) ,

indexing by the degree of the parent Laguerre polynomial rather than by the degree of the result. The two are related by

Lnk(x)=(−1)k Ln+kk(x),L^{k}_{n}(x)=(-1)^{k}\,\mathcal{L}^{k}_{n+k}(x) ,

which is Riley, Hobson and Bence's footnote verbatim in content, and Arfken, Weber and Harris's footnote 4 in the same form. Note that both things change: the sign (−1)k(-1)^{k} and the subscript, by kk. Lnk\mathcal{L}^{k}_{n} has degree n−kn-k, while LnkL^{k}_{n} has degree nn.

This unit uses LnkL^{k}_{n} throughout, the convention of all four sources' main text and of the tables of Abramowitz and Stegun that Arfken, Weber and Harris say they follow. The practical test: in our convention Lnk(0)=(n+kn)>0L^{k}_{n}(0)=\binom{n+k}{n}>0 and the polynomial has degree nn; in the other, Lnk\mathcal{L}^{k}_{n} has degree n−kn-k and Lnk(0)=(−1)k(nn−k)\mathcal{L}^{k}_{n}(0)=(-1)^{k}\binom{n}{n-k} alternates in sign with kk. A hydrogen radial function written with the wrong one goes wrong twice over. It acquires a sign (−1)2L+1=−1(-1)^{2L+1}=-1, which a normalisation constant hides; and — far worse — the subscript is wrong by 2L+12L+1. In our convention the correct factor for the state with principal number NN and orbital number LL is LN−L−12L+1L^{2L+1}_{N-L-1}, of degree N−L−1N-L-1, and the other convention calls that same object −LN+L2L+1-\mathcal{L}^{2L+1}_{N+L}; so a formula carried across without translating the subscript becomes LN+L2L+1L^{2L+1}_{N+L}, of degree N+LN+L. Since LmkL^{k}_{m} has exactly mm simple zeros in (0,∞)(0,\infty), the symptom is a radial function with N+LN+L nodes instead of N−L−1N-L-1 — for 2p2\mathrm{p} (N=2N=2, L=1L=1) three nodes where the true function has none, and for 3d3\mathrm{d} (N=3N=3, L=2L=2) five instead of none. (Degrees and zero counts here were checked symbolically.)

A third difference, independent of this one, is the overall scale: Hassani's standardisation Kn=n!K_{n}=n! is what puts the 1/n!1/n! in front of the Rodrigues formula. Dropping it (as some older texts do) multiplies LnkL^{k}_{n} by n!n! and changes the norm in item 6 from (n+k)!/n!(n+k)!/n! to n! (n+k)!n!\,(n+k)!.

Method(Recognising and solving an associated Laguerre problem)AdaptedAdapted from Arfken, Weber & Harris (2013), §18.3, the substitutions (18.73)–(18.75) and Example 18.3.1, the reduction of the hydrogen radial equation (18.76)–(18.83), pp. 895–897; Riley, Hobson & Bence (2006), §18.8.1, the expansion $f=\sum a_n L^m_n$ and its coefficient formula, p. 623

Many bound-state radial problems arrive in a form with a 1/x1/x and a 1/x21/x^{2} term and no obvious polynomial in sight. The recipe is to strip off the asymptotics at both ends and read off what is left.

  1. Scale so that the exponential decay rate is 1/21/2. If the equation is u′′+(−α24+βr−γr2)u=0u''+\bigl(-\tfrac{\alpha^{2}}{4}+\tfrac{\beta}{r}-\tfrac{\gamma}{r^{2}}\bigr)u=0, put x=αrx=\alpha r. The equation becomes u′′+(−14+β/αx−γx2)u=0u''+\bigl(-\tfrac14+\tfrac{\beta/\alpha}{x}-\tfrac{\gamma}{x^{2}}\bigr)u=0 in the new variable.
  2. Match to the normal form of item 7. Compare with Φ′′+(−14+2n+k+12x−k2−14x2)Φ=0.\Phi''+\Bigl(-\frac14+\frac{2n+k+1}{2x}-\frac{k^{2}-1}{4x^{2}}\Bigr)\Phi=0 . Reading the 1/x21/x^{2} coefficient gives k=4γ+1k=\sqrt{4\gamma+1} — take the positive root, since the other sign makes x(k+1)/2x^{(k+1)/2} singular at the origin. Reading the 1/x1/x coefficient gives n=12(2β/α−k−1)n=\tfrac12\bigl(2\beta/\alpha-k-1\bigr).
  3. Impose integrality. The solution is bounded at infinity only if nn is a non-negative integer. This is the quantisation condition, and solving it for whatever physical parameter sits inside α\alpha or β\beta gives the spectrum.
  4. Write down the solution. Φ=e−x/2x(k+1)/2Lnk(x)\Phi=\mathrm{e}^{-x/2}x^{(k+1)/2}L^{k}_{n}(x), and the original unknown is recovered by undoing step 1.
  5. Normalise. Use item 6: ∫0∞e−xxk(Lnk)2dx=(n+k)!/n!\int_{0}^{\infty}\mathrm{e}^{-x}x^{k}\bigl(L^{k}_{n}\bigr)^{2}\mathrm{d}x=(n+k)!/n!, and multiply by whatever Jacobian the change of variable in step 1 introduced.

To expand a function on (0,∞)(0,\infty): by item 6 the coefficients in f=∑nanLnkf=\sum_{n}a_{n}L^{k}_{n} are

an=n!(n+k)!∫0∞f(x) Lnk(x) xke−x dx.a_{n}=\frac{n!}{(n+k)!}\int_{0}^{\infty}f(x)\,L^{k}_{n}(x)\,x^{k}\mathrm{e}^{-x}\,\mathrm{d}x .

For k=0k=0 the prefactor is 11. Evaluate LnkL^{k}_{n} at a point by the three-term recurrence of item 4, starting from L0k=1L^{k}_{0}=1 and L1k=k+1−xL^{k}_{1}=k+1-x, never by the explicit sum of item 1, whose terms alternate and grow before cancelling.

Example(The hydrogen 2p2\mathrm{p} radial function)Original

Take hydrogen (Z=1Z=1) in the state with principal quantum number N=2N=2 and orbital quantum number L=1L=1, and build the radial function from scratch by the recipe above, then compute ⟨r⟩\langle r\rangle and ⟨1/r⟩\langle 1/r\rangle.

Step 1: reduce the radial equation. With u(r)=r R(r)u(r)=r\,R(r) the radial Schrodinger equation for a Coulomb potential is

u′′+(−α24+2a0r−L(L+1)r2)u=0,α2=−8mEℏ2,u''+\left(-\frac{\alpha^{2}}{4}+\frac{2}{a_{0}r}-\frac{L(L+1)}{r^{2}}\right)u=0, \qquad \alpha^{2}=-\frac{8mE}{\hbar^{2}},

with a0a_{0} the Bohr radius. Set x=αrx=\alpha r. Comparing with the normal form, the 1/x21/x^{2} coefficient gives k2−1=4L(L+1)k^{2}-1=4L(L+1), that is k=2L+1=3k=2L+1=3; the 1/x1/x coefficient gives 2n+k+1=4/(αa0)2n+k+1=4/(\alpha a_{0}), which with k=2L+1k=2L+1 reads 2(n+L+1)=4/(αa0)2(n+L+1)=4/(\alpha a_{0}), that is α=2/(Na0)\alpha=2/(Na_{0}) with N=n+L+1N=n+L+1 the principal quantum number. For N=2N=2, L=1L=1 we get n=N−L−1=0n=N-L-1=0 and α=1/a0\alpha=1/a_{0}.

Step 2: write the function. With n=0n=0 the Laguerre factor is L03=1L^{3}_{0}=1, so

u(r)∝e−x/2x(k+1)/2=e−r/2a0(ra0)2,R21(r)=ur∝ra0 e−r/2a0.u(r)\propto \mathrm{e}^{-x/2}x^{(k+1)/2}=\mathrm{e}^{-r/2a_{0}}\Bigl(\frac{r}{a_{0}}\Bigr)^{2}, \qquad R_{21}(r)=\frac{u}{r}\propto \frac{r}{a_{0}}\,\mathrm{e}^{-r/2a_{0}} .

Step 3: normalise. Require ∫0∞R212r2dr=1\int_{0}^{\infty}R_{21}^{2}r^{2}\mathrm{d}r=1. Writing R21=C r e−r/2a0R_{21}=C\,r\,\mathrm{e}^{-r/2a_{0}},

C2∫0∞r4e−r/a0 dr=C2 4! a05=24 C2a05=1,C^{2}\int_{0}^{\infty}r^{4}\mathrm{e}^{-r/a_{0}}\,\mathrm{d}r=C^{2}\,4!\,a_{0}^{5}=24\,C^{2}a_{0}^{5}=1,

using ∫0∞rpe−r/a0dr=p! a0p+1\int_{0}^{\infty}r^{p}\mathrm{e}^{-r/a_{0}}\mathrm{d}r=p!\,a_{0}^{p+1}. Hence C=1/(26 a05/2)C=1/\bigl(2\sqrt{6}\,a_{0}^{5/2}\bigr) and

R21(r)=126 a03/2 ra0 e−r/2a0.R_{21}(r)=\frac{1}{2\sqrt{6}\,a_{0}^{3/2}}\,\frac{r}{a_{0}}\,\mathrm{e}^{-r/2a_{0}} .

Step 4: the two expectation values. With the same integral formula,

⟨r⟩=∫0∞r R212 r2dr=C2∫0∞r5e−r/a0dr=5! a0624 a05=12024a0=5a0,\langle r\rangle=\int_{0}^{\infty}r\,R_{21}^{2}\,r^{2}\mathrm{d}r=C^{2}\int_{0}^{\infty}r^{5}\mathrm{e}^{-r/a_{0}}\mathrm{d}r =\frac{5!\,a_{0}^{6}}{24\,a_{0}^{5}}=\frac{120}{24}a_{0}=5a_{0} ,⟨1r⟩=C2∫0∞r3e−r/a0dr=3! a0424 a05=624a0=14a0.\Bigl\langle \frac1r\Bigr\rangle=C^{2}\int_{0}^{\infty}r^{3}\mathrm{e}^{-r/a_{0}}\mathrm{d}r=\frac{3!\,a_{0}^{4}}{24\,a_{0}^{5}}=\frac{6}{24a_{0}}=\frac{1}{4a_{0}} .

Step 5: check. Arfken, Weber and Harris quote, for the normalised hydrogenic radial function, ⟨r⟩=a02[3N2−L(L+1)]\langle r\rangle=\tfrac{a_{0}}{2}\bigl[3N^{2}-L(L+1)\bigr] and ⟨r−1⟩=1/(N2a0)\langle r^{-1}\rangle=1/(N^{2}a_{0}). At N=2N=2, L=1L=1 these give a02(12−2)=5a0\tfrac{a_{0}}{2}(12-2)=5a_{0} and 1/(4a0)1/(4a_{0}): both match. The normalisation, ⟨r⟩\langle r\rangle and ⟨1/r⟩\langle 1/r\rangle were also recomputed symbolically from the general formula

RNL(r)=α3/2[(N−L−1)!2N (N+L)!]1/2e−αr/2(αr)LLN−L−12L+1(αr),α=2Na0,R_{NL}(r)=\alpha^{3/2}\left[\frac{(N-L-1)!}{2N\,(N+L)!}\right]^{1/2}\mathrm{e}^{-\alpha r/2}(\alpha r)^{L}L^{2L+1}_{N-L-1}(\alpha r), \qquad \alpha=\frac{2}{Na_{0}} ,

and agree; the same check at N=3N=3, L=0L=0 returns ⟨r⟩=272a0\langle r\rangle=\tfrac{27}{2}a_{0}, again matching the quoted formula.

Reading the numbers. ⟨1/r⟩=1/(N2a0)\langle 1/r\rangle=1/(N^{2}a_{0}) depends only on NN, not on LL — which is why the Coulomb energy −12e2⟨1/r⟩/(4πϵ0)-\tfrac12 e^{2}\langle 1/r\rangle/(4\pi\epsilon_{0}) is degenerate in LL, the accidental degeneracy of the hydrogen atom. By contrast ⟨r⟩=5a0\langle r\rangle=5a_{0} does depend on LL: the 2s2\mathrm{s} state (N=2N=2, L=0L=0) has ⟨r⟩=6a0\langle r\rangle=6a_{0}, further out on average than 2p2\mathrm{p} even though 2s2\mathrm{s} has more weight very close to the nucleus. The two facts together are the whole story of why ⟨r⟩\langle r\rangle is a poor predictor of chemical behaviour and ⟨1/r⟩\langle 1/r\rangle is a good one.

ApplicationOriginal

The hydrogen atom. Separating the Schrodinger equation for a charge ZeZe in spherical coordinates leaves the angular part to the spherical harmonics (Spherical harmonics) and the radial part to the associated Laguerre equation with k=2L+1k=2L+1, as in the example. The bound states are labelled by N=n+L+1=1,2,3,…N=n+L+1=1,2,3,\dots, the energies are

EN=−Z2m2N2ℏ2(e24πϵ0)2,E_{N}=-\frac{Z^{2}m}{2N^{2}\hbar^{2}}\left(\frac{e^{2}}{4\pi\epsilon_{0}}\right)^{2},

and the radial function has n=N−L−1n=N-L-1 nodes, which is the degree of its Laguerre factor. Three structural facts come straight out of the mathematics: the energy depends on NN alone, because the quantisation condition in step 3 of the method involves only the combination 2n+k+1=2N2n+k+1=2N; the degeneracy of level NN is ∑L=0N−1(2L+1)=N2\sum_{L=0}^{N-1}(2L+1)=N^{2}; and the node count is N−L−1N-L-1, so the 1s1\mathrm{s}, 2p2\mathrm{p}, 3d3\mathrm{d}, … states are nodeless. Arfken, Weber and Harris note that the bound states alone are not complete — the positive-energy continuum states are needed for that — which matters whenever one expands a scattering state or a tightly localised wave packet in hydrogenic orbitals.

Gauss-Laguerre quadrature. The nn zeros of LnkL^{k}_{n} are real, simple and in (0,∞)(0,\infty) (General properties of orthogonal polynomials), and there are weights wiw_{i} making

∫0∞xke−xf(x) dx≈∑i=1nwif(xi)\int_{0}^{\infty}x^{k}\mathrm{e}^{-x}f(x)\,\mathrm{d}x\approx\sum_{i=1}^{n}w_{i}f(x_{i})

exact for all polynomials ff of degree at most 2n−12n-1 (Numerical quadrature). This is the standard tool for integrals over a half-line with exponential decay: partition functions and thermal averages, the energy integrals of radiative transfer, and lifetime integrals in reliability work. The decay rate must be matched to the weight first — an integrand behaving like e−3x\mathrm{e}^{-3x} should be rescaled to e−x\mathrm{e}^{-x} before the rule is applied, or the quadrature converges slowly.

Laguerre-Gaussian laser modes. In a cylindrically symmetric cavity the paraxial equation separates in polar coordinates, and the transverse modes are

unℓ(r,ϕ)  ∝  (r2w)∣ℓ∣Ln∣ℓ∣ ⁣(2r2w2)e−r2/w2eiℓϕ.u_{n\ell}(r,\phi)\;\propto\;\left(\frac{r\sqrt2}{w}\right)^{|\ell|}L^{|\ell|}_{n}\!\left(\frac{2r^{2}}{w^{2}}\right)\mathrm{e}^{-r^{2}/w^{2}}\mathrm{e}^{\mathrm{i}\ell\phi} .

The index nn counts radial rings, the index ℓ\ell is the azimuthal winding, and each photon in such a beam carries orbital angular momentum ℓℏ\ell\hbar. These are the beams used in optical tweezers to spin trapped particles and in mode-division multiplexing to carry independent channels on one wavelength; the independence of the channels is item 6, the orthogonality of Ln∣ℓ∣L^{|\ell|}_{n} for different nn at fixed ℓ\ell, together with the orthogonality of eiℓϕ\mathrm{e}^{\mathrm{i}\ell\phi}.

Sonine polynomials in kinetic theory. In the Chapman-Enskog solution of the Boltzmann equation the departure of the velocity distribution from the local Maxwellian is expanded in associated Laguerre polynomials of the reduced squared peculiar speed xx, called Sonine polynomials in that literature and written Sν(r)S^{(r)}_{\nu}, with the two indices in the opposite positions to ours: Sν(r)S^{(r)}_{\nu} is a multiple of our LrνL^{\nu}_{r}, of degree rr in xx. The two tensor ranks of the perturbation use different ν\nu. The vector part, driven by the temperature gradient, is expanded in the S3/2(r)S^{(r)}_{3/2}, orthogonal for the weight x3/2e−xx^{3/2}\mathrm{e}^{-x}, and it is this part that gives the thermal conductivity; the traceless-tensor part, driven by the rate of strain, is expanded in the S5/2(r)S^{(r)}_{5/2}, orthogonal for the weight x5/2e−xx^{5/2}\mathrm{e}^{-x}, and it is this part that gives the viscosity. Both are half-integer values of the parameter, perfectly legal because the construction needs only ν>−1\nu>-1. Each coefficient then comes out as a ratio of determinants built from the matrix elements of the collision operator in its own basis, and the first-order truncation is already accurate to a couple of per cent: for a hard-sphere gas the first Sonine approximation to the viscosity is low by under two per cent, and the one to the thermal conductivity by somewhat more.

Expansions in numerical work on a half-line. Because e−x/2Ln(x)\mathrm{e}^{-x/2}L_{n}(x) is a complete orthonormal sequence in L2[0,∞)L^{2}[0,\infty) (Kreyszig's totality statement), Laguerre functions are the natural spectral basis for problems on a semi-infinite domain — boundary layers, radial transport, the outer region of an atmospheric column — where a Chebyshev or Fourier basis would need an artificial truncation of the domain (Spectral methods). The practical caveat is the same as for the quadrature: the method converges fast only if the solution's decay rate is close to e−x/2\mathrm{e}^{-x/2}, so a scaling parameter must be fitted to the problem.

Definition

Chebyshev polynomials

Adapted

Disguised cosines: Tn(cos⁡θ)=cos⁡nθT_n(\cos\theta)=\cos n\theta gives polynomials with a three-term recurrence, orthogonality for the weight (1−x2)−1/2(1-x^{2})^{-1/2}, and the minimax property that makes 21−nTn2^{1-n}T_n the monic polynomial of degree nn that stays closest to zero on [−1,1][-1,1].

Prerequisites: General properties of orthogonal polynomials, Fourier series

Definition(Chebyshev polynomials: both kinds)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.4, eqs. (18.54)–(18.55) and (18.58), pp. 595–596; Arfken, Weber & Harris (2013), §18.4, eqs. (18.104)–(18.110), p. 904; Kreyszig (1978), §6.4, eq. (6) and footnote 1, p. 348; Hassani (2013), Table 8.1, p. 245

Substituting x=cos⁡θx=\cos\theta into Chebyshev's equation

(1−x2) y′′−x y′+n2y=0(1-x^{2})\,y''-x\,y'+n^{2}y=0

turns it, because d/dx=(−1/sin⁡θ) d/dθ\mathrm{d}/\mathrm{d}x=(-1/\sin\theta)\,\mathrm{d}/\mathrm{d}\theta, into d2y/dθ2+n2y=0\mathrm{d}^{2}y/\mathrm{d}\theta^{2}+n^{2}y=0, whose solutions are cos⁡nθ\cos n\theta and sin⁡nθ\sin n\theta. The Chebyshev polynomials of the first kind are the first of these written back in xx:

Tn(x)=cos⁡(narccos⁡x),−1≤x≤1,n=0,1,2,…T_{n}(x)=\cos\bigl(n\arccos x\bigr),\qquad -1\le x\le 1,\qquad n=0,1,2,\dots

The second solution, Vn(x)=sin⁡(narccos⁡x)V_{n}(x)=\sin(n\arccos x), is not a polynomial. Riley, Hobson and Bence and Arfken, Weber and Harris both define two further functions by dividing out the square root,

Un(x)=Vn+1(x)1−x2=sin⁡((n+1)θ)sin⁡θ,Wn(x)=Tn+1(x)1−x2=cos⁡((n+1)θ)sin⁡θ,U_{n}(x)=\frac{V_{n+1}(x)}{\sqrt{1-x^{2}}}=\frac{\sin\bigl((n+1)\theta\bigr)}{\sin\theta}, \qquad W_{n}(x)=\frac{T_{n+1}(x)}{\sqrt{1-x^{2}}}=\frac{\cos\bigl((n+1)\theta\bigr)}{\sin\theta},

of which UnU_{n} is a polynomial of degree nn — the Chebyshev polynomial of the second kind — and WnW_{n} is not. Exactly two of the four are polynomials, and they are the two that are used.

In the language of General properties of orthogonal polynomials, Hassani places both among the Jacobi polynomials: TnT_{n} is the case μ=ν=−12\mu=\nu=-\tfrac12, with weight (1−x2)−1/2(1-x^{2})^{-1/2} on [−1,1][-1,1], and UnU_{n} the case μ=ν=+12\mu=\nu=+\tfrac12, with weight (1−x2)+1/2(1-x^{2})^{+1/2}. Arfken, Weber and Harris instead obtain them as the ultraspherical polynomials Cn(α)C^{(\alpha)}_{n} at α=0\alpha=0 and α=1\alpha=1, the first of those needing a limit because the generating function degenerates at α=0\alpha=0; the Legendre polynomials (Legendre polynomials) sit between them at α=12\alpha=\tfrac12.

The first few are

T0=1,T1=x,T2=2x2−1,T3=4x3−3x,T4=8x4−8x2+1,U0=1,U1=2x,U2=4x2−1,U3=8x3−4x,U4=16x4−12x2+1.\begin{aligned} T_{0}&=1, & T_{1}&=x, & T_{2}&=2x^{2}-1, & T_{3}&=4x^{3}-3x, & T_{4}&=8x^{4}-8x^{2}+1,\\ U_{0}&=1, & U_{1}&=2x, & U_{2}&=4x^{2}-1, & U_{3}&=8x^{3}-4x, & U_{4}&=16x^{4}-12x^{2}+1 . \end{aligned}

Outside [−1,1][-1,1] the definition is continued by Tn(x)=cosh⁡(n arccosh x)T_{n}(x)=\cosh(n\,\mathrm{arccosh}\,x) for x>1x>1, which is the same polynomial; Stakgold and Holst call this extending the Chebyshev polynomials in the natural way, and it is what makes them useful for iteration counts, where the spectral interval is mapped to [−1,1][-1,1] and the convergence rate is read off at a point outside it.

Definition(What the Chebyshev polynomials satisfy)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.4, the special values and the ODE (18.59), pp. 597–598, and §18.4.1, eqs. (18.61)–(18.72), pp. 599–602; Arfken, Weber & Harris (2013), §18.4, eqs. (18.86), (18.91)–(18.93), (18.97)–(18.98), (18.100)–(18.101), (18.114)–(18.115), pp. 899–905, and the orthogonality weights, p. 906, with the normalisations of Exercises 5.2.5–5.2.6, p. 274; Kreyszig (1978), §6.4, the induction (4) and the recursion (10), pp. 347–349; Hassani (2013), Table 8.2, p. 258, and Problems 8.36–8.37, p. 263

For n≥0n\ge0 and x∈[−1,1]x\in[-1,1], with x=cos⁡θx=\cos\theta and θ∈[0,π]\theta\in[0,\pi]:

  1. Polynomial character. TnT_{n} is a polynomial of degree nn; its leading coefficient is 2n−12^{n-1} for n≥1n\ge1 and 11 for n=0n=0. UnU_{n} is a polynomial of degree nn with leading coefficient 2n2^{n}. Both have the parity of nn: Tn(−x)=(−1)nTn(x)T_{n}(-x)=(-1)^{n}T_{n}(x) and Un(−x)=(−1)nUn(x)U_{n}(-x)=(-1)^{n}U_{n}(x).
  2. Recurrences. With T0=1T_{0}=1, T1=xT_{1}=x, U0=1U_{0}=1, U1=2xU_{1}=2x, Tn+1(x)=2x Tn(x)−Tn−1(x),Un+1(x)=2x Un(x)−Un−1(x).T_{n+1}(x)=2x\,T_{n}(x)-T_{n-1}(x), \qquad U_{n+1}(x)=2x\,U_{n}(x)-U_{n-1}(x) . The two families obey the same recurrence and differ only in the starting pair.
  3. Special values. Tn(1)=1,Tn(−1)=(−1)n,T2m(0)=(−1)m,T2m+1(0)=0,T_{n}(1)=1,\quad T_{n}(-1)=(-1)^{n},\quad T_{2m}(0)=(-1)^{m},\quad T_{2m+1}(0)=0, Un(1)=n+1,Un(−1)=(−1)n(n+1),U2m(0)=(−1)m,U2m+1(0)=0.U_{n}(1)=n+1,\quad U_{n}(-1)=(-1)^{n}(n+1),\quad U_{2m}(0)=(-1)^{m},\quad U_{2m+1}(0)=0 .
  4. Differential equations. TnT_{n} solves (1−x2)y′′−xy′+n2y=0(1-x^{2})y''-xy'+n^{2}y=0, and UnU_{n} solves (1−x2)y′′−3xy′+n(n+2)y=0(1-x^{2})y''-3xy'+n(n+2)y=0.
  5. Zeros and extrema (for n≥1n\ge1). TnT_{n} has nn simple zeros, all in (−1,1)(-1,1), xj=cos⁡(2j−1)π2n,j=1,…,n,x_{j}=\cos\frac{(2j-1)\pi}{2n},\qquad j=1,\dots,n , and n+1n+1 points of extreme value in [−1,1][-1,1], yj=cos⁡jπn,j=0,…,n,withTn(yj)=(−1)j.y_{j}=\cos\frac{j\pi}{n},\qquad j=0,\dots,n, \qquad\text{with}\qquad T_{n}(y_{j})=(-1)^{j} . In particular ∣Tn(x)∣≤1\lvert T_{n}(x)\rvert\le1 on [−1,1][-1,1], with equality exactly at the yjy_{j}.
  6. Orthogonality. ∫−11Tm(x)Tn(x)1−x2 dx={0,m≠n,π,m=n=0,π/2,m=n≥1,∫−11Um(x)Un(x)1−x2 dx=π2 δmn.\int_{-1}^{1}\frac{T_{m}(x)T_{n}(x)}{\sqrt{1-x^{2}}}\,\mathrm{d}x= \begin{cases} 0, & m\neq n,\\ \pi, & m=n=0,\\ \pi/2, & m=n\ge1, \end{cases} \qquad \int_{-1}^{1}U_{m}(x)U_{n}(x)\sqrt{1-x^{2}}\,\mathrm{d}x=\frac{\pi}{2}\,\delta_{mn}. The exceptional value at m=n=0m=n=0 in the first kind is not a blemish; it is the reason the constant term of a Chebyshev series is conventionally written 12a0\tfrac12 a_{0}.
  7. Generating functions. For ∣h∣<1|h|<1 and ∣x∣≤1|x|\le1, 1−xh1−2xh+h2=∑n=0∞Tn(x)hn,11−2xh+h2=∑n=0∞Un(x)hn.\frac{1-xh}{1-2xh+h^{2}}=\sum_{n=0}^{\infty}T_{n}(x)h^{n}, \qquad \frac{1}{1-2xh+h^{2}}=\sum_{n=0}^{\infty}U_{n}(x)h^{n} . Arfken, Weber and Harris and Hassani give the first in the equivalent form 1−h21−2xh+h2=T0(x)+2∑n≥1Tn(x)hn\dfrac{1-h^{2}}{1-2xh+h^{2}}=T_{0}(x)+2\sum_{n\ge1}T_{n}(x)h^{n}, which is twice the display above minus one.
  8. Relations between the two kinds. Tn′(x)=n Un−1(x),Tn(x)=Un(x)−x Un−1(x),(1−x2)Un−1(x)=x Tn(x)−Tn+1(x).T_{n}'(x)=n\,U_{n-1}(x), \qquad T_{n}(x)=U_{n}(x)-x\,U_{n-1}(x), \qquad (1-x^{2})U_{n-1}(x)=x\,T_{n}(x)-T_{n+1}(x).
Definition(The minimax property)AdaptedAdapted from Kreyszig (1978), §6.4, Theorem 6.4-3 and the surrounding reduction of the problem, pp. 346–349, and §6.3, the Haar condition 6.3-2 with the uniqueness it is stated to give on p. 337, invoked in the proof of Lemma 6.4-2, p. 345; Stakgold & Holst (2011), §10.6, Exercise 10.6.2 part 2, p. 808

Let n≥1n\ge1 and write ∥q∥∞=max⁡−1≤x≤1∣q(x)∣\lVert q\rVert_{\infty}=\max_{-1\le x\le1}\lvert q(x)\rvert. Among all monic real polynomials qq of degree nn,

min⁡q monic, deg⁡q=n ∥q∥∞  =  2 1−n,\min_{q\ \text{monic},\ \deg q=n}\ \lVert q\rVert_{\infty}\;=\;2^{\,1-n},

and the minimum is attained at q=21−nTnq=2^{1-n}T_{n} and at no other monic polynomial of degree nn.

Equivalently — and this is the form in which Kreyszig poses it — the best uniform approximation to xnx^{n} on [−1,1][-1,1] by a polynomial of degree at most n−1n-1 is

y(x)=xn−21−nTn(x),y(x)=x^{n}-2^{1-n}T_{n}(x),

with maximum error 21−n2^{1-n}; and if p(x)=βnxn+…p(x)=\beta_{n}x^{n}+\dots is any real polynomial of degree exactly nn, its best uniform approximation by a polynomial of degree at most n−1n-1 is p(x)−βn21−nTn(x)p(x)-\beta_{n}2^{1-n}T_{n}(x), with maximum error ∣βn∣ 21−n\lvert\beta_{n}\rvert\,2^{1-n}.

The minimality is Kreyszig's Theorem 6.4-3. Uniqueness is not part of that theorem's statement; it follows from the general result he records on p. 337, and invokes at the start of the proof of his Lemma 6.4-2 — that a finite-dimensional subspace of C[a,b]C[a,b] supplies exactly one closest element to each function precisely when the Haar condition holds in it — together with the observation that the polynomials of degree at most n−1n-1 form an nn-dimensional subspace satisfying the Haar condition, since a non-zero one of them has at most n−1n-1 zeros. Stakgold and Holst state the same minimality and attribute it to the equi-oscillation of TnT_{n}, which is exactly item 5 above.

Idea of the proofOriginal

Chebyshev polynomials are cosines wearing a polynomial costume, and almost every property is a trigonometric identity with the costume put back on.

That cos⁡nθ\cos n\theta is a polynomial in cos⁡θ\cos\theta is the one thing needing an argument, and it is the addition formula: cos⁡(n+1)θ+cos⁡(n−1)θ=2cos⁡θcos⁡nθ\cos(n+1)\theta+\cos(n-1)\theta=2\cos\theta\cos n\theta. Read as a statement about polynomials this is the three-term recurrence, and running it from 11 and xx both proves the polynomial character and doubles the leading coefficient at each step, which is where 2n−12^{n-1} comes from. Orthogonality is then the orthogonality of cosines on [0,π][0,\pi]: the weight (1−x2)−1/2(1-x^{2})^{-1/2} is nothing but the Jacobian dx=−sin⁡θ dθ\mathrm{d}x=-\sin\theta\,\mathrm{d}\theta in disguise, and a Chebyshev series is literally a Fourier cosine series of f(cos⁡θ)f(\cos\theta).

The minimax property is the one genuinely different idea. TnT_{n} oscillates between +1+1 and −1-1, touching both n+1n+1 times, and that is a lot of alternations for a polynomial of degree nn. Suppose some monic qq of degree nn stayed strictly closer to zero than 21−nTn2^{1-n}T_{n} does. Then the difference 21−nTn−q2^{1-n}T_{n}-q would inherit the sign of 21−nTn2^{1-n}T_{n} at each of the n+1n+1 alternation points, so it would change sign nn times and have nn roots. But the leading terms cancel, so the difference has degree at most n−1n-1 and cannot have nn roots unless it vanishes identically. The hypothesis is contradicted, and the only escape is that qq was 21−nTn2^{1-n}T_{n} all along.

ProofAdaptedAdapted from Kreyszig (1978), §6.4, the induction (4)–(5) and the recursion (10), pp. 347–349, and Theorem 6.4-3, p. 348; Riley, Hobson & Bence (2006), §18.4.1, the derivations of (18.69)–(18.70) from the cosine addition formulas and of the normalisation (18.62) by the substitution $x=\cos\theta$, pp. 600–602; Arfken, Weber & Harris (2013), §18.4, the trigonometric forms (18.104)–(18.110), p. 904

Throughout, x=cos⁡θx=\cos\theta with θ∈[0,π]\theta\in[0,\pi], so that θ=arccos⁡x\theta=\arccos x is a decreasing bijection of [−1,1][-1,1] onto [0,π][0,\pi] and sin⁡θ=1−x2≥0\sin\theta=\sqrt{1-x^{2}}\ge0.

Item 2, the recurrences. Adding the two cosine addition formulas cos⁡(n±1)θ=cos⁡nθcos⁡θ∓sin⁡nθsin⁡θ\cos(n\pm1)\theta=\cos n\theta\cos\theta\mp\sin n\theta\sin\theta cancels the sine terms:

cos⁡(n+1)θ+cos⁡(n−1)θ=2cos⁡θ cos⁡nθ.\cos(n+1)\theta+\cos(n-1)\theta=2\cos\theta\,\cos n\theta .

Written in xx, this is Tn+1+Tn−1=2xTnT_{n+1}+T_{n-1}=2xT_{n}. Adding the two sine addition formulas sin⁡(n±1)θ=sin⁡nθcos⁡θ±cos⁡nθsin⁡θ\sin(n\pm1)\theta=\sin n\theta\cos\theta\pm\cos n\theta\sin\theta likewise gives sin⁡(n+1)θ+sin⁡(n−1)θ=2cos⁡θsin⁡nθ\sin(n+1)\theta+\sin(n-1)\theta=2\cos\theta\sin n\theta; dividing by sin⁡θ\sin\theta (non-zero for 0<θ<π0<\theta<\pi) and shifting the index by one gives Un+1+Un−1=2xUnU_{n+1}+U_{n-1}=2xU_{n}, first on (−1,1)(-1,1) and then at x=±1x=\pm1 by continuity, both sides being continuous.

Item 1, from item 2. T0=1T_{0}=1 and T1=xT_{1}=x are polynomials of degrees 00 and 11 with leading coefficients 11 and 1=201=2^{0}. If Tn−1T_{n-1} and TnT_{n} are polynomials of degrees n−1n-1 and nn with TnT_{n} having leading coefficient 2n−12^{n-1}, then Tn+1=2xTn−Tn−1T_{n+1}=2xT_{n}-T_{n-1} has degree n+1n+1 with leading coefficient 2⋅2n−1=2n2\cdot2^{n-1}=2^{n}, the subtracted term being of lower degree. Induction gives the claim for all n≥1n\ge1. (Kreyszig proves the same statement in the equivalent trigonometric form, as the existence of an expansion cos⁡nθ=2n−1cos⁡nθ+∑j<nβnjcos⁡jθ\cos n\theta=2^{n-1}\cos^{n}\theta+\sum_{j<n}\beta_{nj}\cos^{j}\theta.) The argument for UnU_{n} is identical, starting from U0=1U_{0}=1 and U1=2xU_{1}=2x and giving leading coefficient 2n2^{n}.

For the parity, note that x↦−xx\mapsto-x corresponds to θ↦π−θ\theta\mapsto\pi-\theta, under which cos⁡nθ↦cos⁡(nπ−nθ)=(−1)ncos⁡nθ\cos n\theta\mapsto\cos(n\pi-n\theta)=(-1)^{n}\cos n\theta; and sin⁡((n+1)(π−θ))/sin⁡(π−θ)=(−1)nsin⁡((n+1)θ)/sin⁡θ\sin((n+1)(\pi-\theta))/\sin(\pi-\theta)=(-1)^{n}\sin((n+1)\theta)/\sin\theta by the same computation. Alternatively the recurrence shows by induction that TnT_{n} contains only powers of the parity of nn.

Item 3. Put θ=0\theta=0, θ=π\theta=\pi and θ=π/2\theta=\pi/2 in Tn=cos⁡nθT_{n}=\cos n\theta: these are x=1x=1, x=−1x=-1, x=0x=0, giving 11, cos⁡nπ=(−1)n\cos n\pi=(-1)^{n} and cos⁡(nπ/2)\cos(n\pi/2), which is (−1)m(-1)^{m} for n=2mn=2m and 00 for nn odd. For Un=sin⁡((n+1)θ)/sin⁡θU_{n}=\sin((n+1)\theta)/\sin\theta the values at θ=0,π\theta=0,\pi need a limit: as θ→0\theta\to0, sin⁡((n+1)θ)/sin⁡θ→n+1\sin((n+1)\theta)/\sin\theta\to n+1 by L'Hopital's rule or by comparing the leading terms of the two sines, and as θ→π\theta\to\pi one gets (−1)n(n+1)(-1)^{n}(n+1) the same way. At θ=π/2\theta=\pi/2, sin⁡((n+1)π/2)/1\sin((n+1)\pi/2)/1 is (−1)m(-1)^{m} for n=2mn=2m and 00 for nn odd.

Item 4. Differentiating Tn=cos⁡nθT_{n}=\cos n\theta with d/dx=(−1/sin⁡θ) d/dθ\mathrm{d}/\mathrm{d}x=(-1/\sin\theta)\,\mathrm{d}/\mathrm{d}\theta reverses the substitution made in the definition: y=cos⁡nθy=\cos n\theta satisfies yθθ+n2y=0y_{\theta\theta}+n^{2}y=0, and the chain rule converts that back to (1−x2)y′′−xy′+n2y=0(1-x^{2})y''-xy'+n^{2}y=0. For UnU_{n}, note that Vn+1=1−x2 UnV_{n+1}=\sqrt{1-x^{2}}\,U_{n} solves Chebyshev's equation with nn replaced by n+1n+1, since Vn+1=sin⁡((n+1)θ)V_{n+1}=\sin((n+1)\theta) is the other solution of yθθ+(n+1)2y=0y_{\theta\theta}+(n+1)^{2}y=0. Substituting Vn+1=(1−x2)1/2UnV_{n+1}=(1-x^{2})^{1/2}U_{n} and its derivatives

Vn+1′=(1−x2)1/2Un′−x(1−x2)−1/2Un,V_{n+1}'=(1-x^{2})^{1/2}U_{n}'-x(1-x^{2})^{-1/2}U_{n},Vn+1′′=(1−x2)1/2Un′′−2x(1−x2)−1/2Un′−(1−x2)−1/2Un−x2(1−x2)−3/2UnV_{n+1}''=(1-x^{2})^{1/2}U_{n}''-2x(1-x^{2})^{-1/2}U_{n}'-(1-x^{2})^{-1/2}U_{n}-x^{2}(1-x^{2})^{-3/2}U_{n}

into (1−x2)Vn+1′′−xVn+1′+(n+1)2Vn+1=0(1-x^{2})V_{n+1}''-xV_{n+1}'+(n+1)^{2}V_{n+1}=0 and dividing by (1−x2)1/2(1-x^{2})^{1/2} collects to

(1−x2)Un′′−3xUn′−Un+(n+1)2Un=0,(1-x^{2})U_{n}''-3xU_{n}'-U_{n}+(n+1)^{2}U_{n}=0 ,

and (n+1)2−1=n(n+2)(n+1)^{2}-1=n(n+2).

Item 5. Tn(x)=0T_{n}(x)=0 means cos⁡nθ=0\cos n\theta=0, that is nθ=(2j−1)π/2n\theta=(2j-1)\pi/2 for an integer jj; the values of θ\theta in (0,π)(0,\pi) are θj=(2j−1)π/(2n)\theta_{j}=(2j-1)\pi/(2n) for j=1,…,nj=1,\dots,n, and these give nn distinct values xj=cos⁡θjx_{j}=\cos\theta_{j} in (−1,1)(-1,1) because cos⁡\cos is injective on [0,π][0,\pi]. A polynomial of degree nn has at most nn roots, so these are all of them and each is simple. Similarly ∣Tn∣=∣cos⁡nθ∣\lvert T_{n}\rvert=\lvert\cos n\theta\rvert attains the value 11 exactly when nθn\theta is a multiple of π\pi, that is at θ=jπ/n\theta=j\pi/n for j=0,…,nj=0,\dots,n, where Tn=cos⁡jπ=(−1)jT_{n}=\cos j\pi=(-1)^{j}; and ∣cos⁡nθ∣≤1\lvert\cos n\theta\rvert\le1 everywhere.

Item 6. Substitute x=cos⁡θx=\cos\theta, so dx=−sin⁡θ dθ\mathrm{d}x=-\sin\theta\,\mathrm{d}\theta and 1−x2=sin⁡θ\sqrt{1-x^{2}}=\sin\theta, which cancels:

∫−11TmTn1−x2 dx=∫0πcos⁡mθ cos⁡nθ dθ.\int_{-1}^{1}\frac{T_{m}T_{n}}{\sqrt{1-x^{2}}}\,\mathrm{d}x=\int_{0}^{\pi}\cos m\theta\,\cos n\theta\,\mathrm{d}\theta .

This is the standard cosine orthogonality on [0,π][0,\pi] (Fourier series), equal to 00 for m≠nm\neq n, to π/2\pi/2 for m=n≥1m=n\ge1, and to π\pi for m=n=0m=n=0 where the integrand is 11. For the second kind, UmUn1−x2 dx=−sin⁡((m+1)θ)sin⁡((n+1)θ) dθU_{m}U_{n}\sqrt{1-x^{2}}\,\mathrm{d}x=-\sin((m+1)\theta)\sin((n+1)\theta)\,\mathrm{d}\theta after the same substitution, and ∫0πsin⁡((m+1)θ)sin⁡((n+1)θ) dθ=π2δmn\int_{0}^{\pi}\sin((m+1)\theta)\sin((n+1)\theta)\,\mathrm{d}\theta=\tfrac{\pi}{2}\delta_{mn}, with no exceptional case because m+1≥1m+1\ge1 always.

Item 7. Sum the geometric-like series directly. For ∣h∣<1|h|<1 and x=cos⁡θx=\cos\theta, ∑n≥0cos⁡(nθ)hn=Re⁡∑n≥0(heiθ)n=Re⁡11−heiθ\sum_{n\ge0}\cos(n\theta)h^{n}=\operatorname{Re}\sum_{n\ge0}(h\mathrm{e}^{\mathrm{i}\theta})^{n}=\operatorname{Re}\dfrac{1}{1-h\mathrm{e}^{\mathrm{i}\theta}}, and multiplying numerator and denominator by the conjugate 1−he−iθ1-h\mathrm{e}^{-\mathrm{i}\theta} gives

1−hcos⁡θ1−2hcos⁡θ+h2=1−xh1−2xh+h2.\frac{1-h\cos\theta}{1-2h\cos\theta+h^{2}}=\frac{1-xh}{1-2xh+h^{2}} .

Taking the imaginary part instead gives ∑n≥1sin⁡(nθ)hn=hsin⁡θ/(1−2xh+h2)\sum_{n\ge1}\sin(n\theta)h^{n}=h\sin\theta/(1-2xh+h^{2}); dividing by sin⁡θ\sin\theta and shifting the index gives the generating function for UnU_{n}.

Item 8. Tn′(x)=(−1/sin⁡θ) d(cos⁡nθ)/dθ=nsin⁡nθ/sin⁡θ=nUn−1(x)T_{n}'(x)=(-1/\sin\theta)\,\mathrm{d}(\cos n\theta)/\mathrm{d}\theta=n\sin n\theta/\sin\theta=nU_{n-1}(x). The second relation is the sine addition formula sin⁡((n+1)θ)−cos⁡θsin⁡(nθ)=cos⁡(nθ)sin⁡θ\sin((n+1)\theta)-\cos\theta\sin(n\theta)=\cos(n\theta)\sin\theta divided by sin⁡θ\sin\theta. The third comes from the cosine addition formula in the form cos⁡θcos⁡(nθ)−cos⁡((n+1)θ)=sin⁡θ sin⁡(nθ)\cos\theta\cos(n\theta)-\cos\bigl((n+1)\theta\bigr)=\sin\theta\,\sin(n\theta): the left side is xTn−Tn+1xT_{n}-T_{n+1}, and the right side is sin⁡θ⋅sin⁡θ Un−1=(1−x2)Un−1\sin\theta\cdot\sin\theta\,U_{n-1}=(1-x^{2})U_{n-1}. □\square

Proof(of the minimax property)AdaptedAdapted from Kreyszig (1978), §6.4, Theorem 6.4-3 and Lemma 6.4-2 with the alternating-set idea, pp. 345–348, and §6.3, the Haar condition and the uniqueness it gives, p. 337

Write T~n=21−nTn\widetilde{T}_{n}=2^{1-n}T_{n}. By item 1 of the first statement, T~n\widetilde{T}_{n} is monic of degree nn, and by item 5 it satisfies ∥T~n∥∞=21−n\lVert\widetilde{T}_{n}\rVert_{\infty}=2^{1-n}, attained with alternating signs at the n+1n+1 points yj=cos⁡(jπ/n)y_{j}=\cos(j\pi/n):

T~n(yj)=(−1)j21−n,j=0,1,…,n.\widetilde{T}_{n}(y_{j})=(-1)^{j}2^{1-n},\qquad j=0,1,\dots,n .

Note 1=y0>y1>⋯>yn=−11=y_{0}>y_{1}>\dots>y_{n}=-1, since cos⁡\cos is decreasing on [0,π][0,\pi].

The minimum is not smaller. Let qq be monic of degree nn and suppose, for contradiction, that ∥q∥∞<21−n\lVert q\rVert_{\infty}<2^{1-n}. Put

d=T~n−q.d=\widetilde{T}_{n}-q .

Both are monic of degree nn, so the xnx^{n} terms cancel and deg⁡d≤n−1\deg d\le n-1. At each yjy_{j},

(−1)jd(yj)=21−n−(−1)jq(yj)  ≥  21−n−∥q∥∞  >  0,(-1)^{j}d(y_{j})=2^{1-n}-(-1)^{j}q(y_{j})\;\ge\;2^{1-n}-\lVert q\rVert_{\infty}\;>\;0 ,

so d(yj)d(y_{j}) is non-zero and has the sign of (−1)j(-1)^{j}. Since dd is continuous and changes sign on each of the nn intervals (yj+1,yj)(y_{j+1},y_{j}), j=0,…,n−1j=0,\dots,n-1, the intermediate value theorem gives a root of dd strictly inside each, hence nn distinct roots. A polynomial of degree at most n−1n-1 with nn distinct roots is identically zero, so d≡0d\equiv0 and q=T~nq=\widetilde{T}_{n}; but then ∥q∥∞=21−n\lVert q\rVert_{\infty}=2^{1-n}, contradicting the assumed strict inequality. Therefore ∥q∥∞≥21−n\lVert q\rVert_{\infty}\ge2^{1-n} for every monic qq of degree nn, and T~n\widetilde{T}_{n} attains the bound.

Uniqueness. Minimising ∥q∥∞\lVert q\rVert_{\infty} over monic qq of degree nn is the same problem as minimising ∥xn−y∥∞\lVert x^{n}-y\rVert_{\infty} over yy in the subspace YY of polynomials of degree at most n−1n-1, by the correspondence q=xn−yq=x^{n}-y. Now dim⁡Y=n\dim Y=n, and a non-zero element of YY has at most n−1n-1 zeros in [−1,1][-1,1], so YY satisfies the Haar condition (Kreyszig's Definition 6.3-2). Kreyszig's theory then gives a unique best uniform approximation to xnx^{n} out of YY — this is the fact he states on p. 337 and uses at the start of the proof of his Lemma 6.4-2 — and hence a unique minimising monic qq, which by the previous paragraph is T~n\widetilde{T}_{n}.

The two restatements. The first is just the correspondence above read backwards: the best approximation is y=xn−T~ny=x^{n}-\widetilde{T}_{n}, whose degree really is at most n−1n-1 because the leading terms cancel, and the error is ∥T~n∥∞=21−n\lVert\widetilde{T}_{n}\rVert_{\infty}=2^{1-n}. For the second, write p=βnxn+(lower)p=\beta_{n}x^{n}+(\text{lower}) with βn≠0\beta_{n}\ne0. Then p−yp-y ranges, as yy runs over YY, over all polynomials of degree exactly nn with leading coefficient βn\beta_{n}, that is over βnq\beta_{n}q with qq monic; and ∥βnq∥∞=∣βn∣ ∥q∥∞\lVert\beta_{n}q\rVert_{\infty}=\lvert\beta_{n}\rvert\,\lVert q\rVert_{\infty} is minimised exactly when ∥q∥∞\lVert q\rVert_{\infty} is. □\square

Remark(Where the hypotheses are used)Original

The interval must be [−1,1][-1,1], and rescaling costs a factor. Everything above is tied to [−1,1][-1,1] through x=cos⁡θx=\cos\theta. On a general interval [a,b][a,b], substitute

x=2t−a−bb−a,x=\frac{2t-a-b}{b-a},

which maps [a,b][a,b] onto [−1,1][-1,1]. The monic polynomial of degree nn in tt of smallest sup-norm on [a,b][a,b] is then (b−a2)nT~n(x)\bigl(\tfrac{b-a}{2}\bigr)^{n}\widetilde{T}_{n}(x), and its sup-norm is

2(b−a4)n.2\left(\frac{b-a}{4}\right)^{n} .

The factor shrinks geometrically as the interval shrinks, which is the reason an interpolation error estimate on a short interval is so much better than on a long one, and the reason adaptive subdivision works.

Monic is essential; without it the problem is trivial. Drop the constraint on the leading coefficient and the infimum of ∥q∥∞\lVert q\rVert_{\infty} over polynomials of degree nn is 00, attained by q≡0q\equiv0. The minimax theorem is a statement about normalising the top coefficient, and every application reads it that way: in approximation one normalises the leading coefficient of the error polynomial, in iterative methods one normalises the residual polynomial by P(0)=1P(0)=1.

The norm must be the maximum norm. In the L2L^{2} norm with weight (1−x2)−1/2(1-x^{2})^{-1/2} the minimising monic polynomial is also T~n\widetilde{T}_{n}, because orthogonal polynomials always minimise the weighted L2L^{2} norm among monic polynomials of their degree (General properties of orthogonal polynomials); but in the unweighted L2L^{2} norm on [−1,1][-1,1] it is the monic Legendre polynomial instead (Legendre polynomials), and in the L1L^{1} norm it is 2−nUn2^{-n}U_{n}. Different norms, different extremal polynomials. What is special about TnT_{n} is the max-norm statement, and that is why it governs worst-case error rather than mean-square error.

The equi-oscillation is doing the work, not the orthogonality. The proof of the minimax property never used item 6. It used only that T~n\widetilde{T}_{n} is monic and that it attains ±∥T~n∥∞\pm\lVert\widetilde{T}_{n}\rVert_{\infty} alternately at n+1n+1 points. This is why Kreyszig develops the alternating set as the central concept, and why Stakgold and Holst describe the result as being due to an equi-oscillation property. A family with the same alternation count would serve equally well; it happens that the cosines supply it.

Counting the alternations exactly. The argument would fail with only nn alternation points, because then dd would be forced to have only n−1n-1 roots, which a polynomial of degree n−1n-1 can have. The margin is exactly one: TnT_{n} has n+1n+1 extreme points on a closed interval, including both endpoints, and dropping either endpoint from [−1,1][-1,1] destroys the proof. This is also why the minimax polynomial for a general function equi-oscillates at n+2n+2 points rather than n+1n+1 when the approximating space has dimension n+1n+1.

Remark(Conventions: which family is "the second kind")AdaptedAdapted from Kreyszig (1978), §6.4, footnote 1 to Theorem 6.4-3, p. 348; Riley, Hobson & Bence (2006), §18.4, eqs. (18.55) and (18.58), pp. 596–597; Arfken, Weber & Harris (2013), §18.4, eqs. (18.106)–(18.107), p. 904, and footnote 1 on the spelling, p. 871; Hassani (2013), Problem 8.37, p. 263

Our sources do not agree, and the disagreement is a real one rather than a matter of scale.

  • Riley, Hobson and Bence, Arfken, Weber and Harris, and Hassani all set Un(cos⁡θ)=sin⁡((n+1)θ)sin⁡θ,U_{n}(\cos\theta)=\frac{\sin\bigl((n+1)\theta\bigr)}{\sin\theta}, a polynomial of degree nn with Un(1)=n+1U_{n}(1)=n+1. Riley, Hobson and Bence and Arfken, Weber and Harris give the non-polynomial sin⁡(nθ)\sin(n\theta) the separate name VnV_{n}.
  • Kreyszig, in footnote 1 on p. 348, attaches the label second kind to sin⁡nθ\sin n\theta instead, writing it Un(t)U_{n}(t) as well. That is the function the other three call VnV_{n}, and it is not a polynomial.

This unit follows the first group, so UnU_{n} here always means sin⁡((n+1)θ)/sin⁡θ\sin((n+1)\theta)/\sin\theta, the degree-nn polynomial tabulated in item 1, orthogonal for the weight (1−x2)+1/2(1-x^{2})^{+1/2}. Nothing in this unit uses Kreyszig's UnU_{n}; his contribution here is the minimax theorem, which concerns TnT_{n} only and is unaffected. The quick test on a formula from an unfamiliar source: evaluate at x=1x=1. Our Un(1)=n+1U_{n}(1)=n+1; Kreyszig's vanishes identically there, since sin⁡nθ→0\sin n\theta\to0 as θ→0\theta\to0.

Two smaller conventions are worth recording. First, the spelling: Arfken, Weber and Harris note in a footnote that they take "Chebyshev" from the tables of Abramowitz and Stegun, and that forms such as "Tschebyscheff" are also found; the symbol TT comes from one of those transliterations, as Kreyszig remarks. Second, the normalisation of the orthogonality integral: all our sources use the TnT_{n} with Tn(1)=1T_{n}(1)=1 and therefore all get π\pi at n=0n=0 and π/2\pi/2 for n≥1n\ge1, but an orthonormal convention rescaling TnT_{n} by 2/π\sqrt{2/\pi} is common in the numerical-analysis literature and changes every coefficient formula below by that factor.

Method(Chebyshev approximation and economisation)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.4.1, the expansion $f=\tfrac12a_0+\sum a_nT_n$ with its coefficient formula, p. 600, and the use of the recurrences for practical computation, p. 602; Kreyszig (1978), §6.4, eqs. (8)–(9), the subtraction of a scaled $T_n$ from a polynomial, p. 349; Arfken, Weber & Harris (2013), §18.4, the discussion of the truncation error and Example 18.4.1, pp. 905–906

To expand a function in Chebyshev polynomials. For ff continuous on [−1,1][-1,1]:

  1. Write the Chebyshev series f(x)=12a0+∑n≥1anTn(x)f(x)=\tfrac12a_{0}+\sum_{n\ge1}a_{n}T_{n}(x) with an=2π∫−11f(x)Tn(x)1−x2 dx  =  2π∫0πf(cos⁡θ)cos⁡(nθ) dθ.a_{n}=\frac{2}{\pi}\int_{-1}^{1}\frac{f(x)T_{n}(x)}{\sqrt{1-x^{2}}}\,\mathrm{d}x \;=\;\frac{2}{\pi}\int_{0}^{\pi}f(\cos\theta)\cos(n\theta)\,\mathrm{d}\theta . The halved zeroth term is exactly what the exceptional normalisation in item 6 forces. The second form shows what a Chebyshev series is: the Fourier cosine series of f(cos⁡θ)f(\cos\theta) (Fourier series). Everything known about the convergence of Fourier series transfers, with one bonus — f(cos⁡θ)f(\cos\theta) is automatically even and 2π2\pi-periodic, so there is no jump at the ends of the interval and no Gibbs phenomenon there, however badly ff behaves at ±1\pm1 in the xx variable.
  2. In practice evaluate the ana_{n} by the discrete cosine transform: sample ff at the N+1N+1 points yj=cos⁡(jπ/N)y_{j}=\cos(j\pi/N) and transform. This costs O(Nlog⁡N)O(N\log N) with a fast algorithm (The discrete Fourier transform and the FFT) and returns the coefficients of the polynomial interpolating ff at those points, which for smooth ff differs from the truncated series only in the tail.
  3. Truncate after TNT_{N}. If the coefficients are decaying fast, the error is dominated by the first omitted term, so it is close to aN+1TN+1(x)a_{N+1}T_{N+1}(x): of size ∣aN+1∣\lvert a_{N+1}\rvert, nearly equal in magnitude at N+2N+2 places and alternating in sign. Arfken, Weber and Harris make exactly this point — the error maxima come out comparable, which approximately minimises the largest one, whereas a Taylor polynomial is excellent at one point and poor at the ends.
  4. Evaluate the truncated series by the recurrence of item 2, never by expanding into powers of xx: the expansion into monomials involves large cancelling coefficients and loses precision.

To economise a polynomial. Given pp of degree nn with leading coefficient βn\beta_{n}, and a tolerance ε\varepsilon:

  1. Compute ∣βn∣ 21−n\lvert\beta_{n}\rvert\,2^{1-n}. If it exceeds ε\varepsilon, stop: the degree cannot be reduced within tolerance.
  2. Otherwise replace pp by p−βn21−nTnp-\beta_{n}2^{1-n}T_{n}, which has degree at most n−1n-1 and differs from pp by at most ∣βn∣21−n\lvert\beta_{n}\rvert2^{1-n} anywhere on [−1,1][-1,1]. This is Kreyszig's formula (9).
  3. Repeat on the new polynomial, accumulating the error bounds, until step 1 stops you.

This is the classical way to shorten a truncated Taylor series, and it is strictly better than simply dropping the top terms: dropping βnxn\beta_{n}x^{n} costs ∣βn∣\lvert\beta_{n}\rvert, while economising it costs only ∣βn∣21−n\lvert\beta_{n}\rvert2^{1-n} — a saving of a factor 2n−12^{n-1}.

Example(A cubic for ex\mathrm{e}^{x} on [−1,1][-1,1])Original

Approximate f(x)=exf(x)=\mathrm{e}^{x} on [−1,1][-1,1] by a cubic, three ways, and compare the worst-case errors.

Step 1: the Chebyshev coefficients. By the second form of the coefficient formula,

ak=2π∫0πecos⁡θcos⁡(kθ) dθ=2 Ik(1),a_{k}=\frac{2}{\pi}\int_{0}^{\pi}\mathrm{e}^{\cos\theta}\cos(k\theta)\,\mathrm{d}\theta=2\,I_{k}(1),

where IkI_{k} is the modified Bessel function of the first kind, whose standard integral representation is Ik(z)=1π∫0πezcos⁡θcos⁡(kθ) dθI_{k}(z)=\frac1\pi\int_{0}^{\pi}\mathrm{e}^{z\cos\theta}\cos(k\theta)\,\mathrm{d}\theta (Modified Bessel functions). So the coefficients are known in closed form:

12a0=I0(1)=1.2660659,a1=2I1(1)=1.1303182,\tfrac12a_{0}=I_{0}(1)=1.2660659,\quad a_{1}=2I_{1}(1)=1.1303182,a2=2I2(1)=0.2714953,a3=2I3(1)=0.0443368,a4=2I4(1)=0.0054742.a_{2}=2I_{2}(1)=0.2714953,\quad a_{3}=2I_{3}(1)=0.0443368,\quad a_{4}=2I_{4}(1)=0.0054742 .

(The quadrature and the Bessel values were computed independently and agree to 25 digits.) Note how fast they fall: each is roughly a fifth to a tenth of the last.

Step 2: the truncated series. The cubic truncation is

p3(x)=12a0+a1T1+a2T2+a3T3,p_{3}(x)=\tfrac12a_{0}+a_{1}T_{1}+a_{2}T_{2}+a_{3}T_{3},

and substituting T2=2x2−1T_{2}=2x^{2}-1, T3=4x3−3xT_{3}=4x^{3}-3x and collecting powers,

p3(x)=0.9945705+0.9973077 x+0.5429907 x2+0.1773474 x3.p_{3}(x)=0.9945705+0.9973077\,x+0.5429907\,x^{2}+0.1773474\,x^{3} .

Its largest error on [−1,1][-1,1], found by sampling on a fine grid, is

max⁡[−1,1]∣ex−p3(x)∣=0.0060656,\max_{[-1,1]}\lvert \mathrm{e}^{x}-p_{3}(x)\rvert=0.0060656 ,

attained at x=1x=1; at x=−1x=-1 the error is 0.00497330.0049733, and the interior extrema are of comparable size with alternating signs. The first omitted coefficient predicted this: ∣a4∣=0.0054742\lvert a_{4}\rvert=0.0054742, within ten per cent of the realised maximum, exactly as step 3 of the method says.

Step 3: the alternatives. The cubic Taylor polynomial 1+x+x22+x361+x+\tfrac{x^{2}}{2}+\tfrac{x^{3}}{6} has

max⁡[−1,1]∣ex−(1+x+x22+x36)∣=e−83=0.0516152,\max_{[-1,1]}\bigl\lvert \mathrm{e}^{x}-\bigl(1+x+\tfrac{x^{2}}{2}+\tfrac{x^{3}}{6}\bigr)\bigr\rvert=\mathrm{e}-\tfrac{8}{3}=0.0516152 ,

attained at x=1x=1 and tiny near x=0x=0 — superb at one point and eight and a half times worse than the Chebyshev truncation where it matters. The true minimax cubic, computed by the exchange algorithm, has error

E3=0.0055284.E_{3}=0.0055284 .

Step 4: the comparison.

0.00606560.0055284=1.097,0.05161520.0055284=9.336.\frac{0.0060656}{0.0055284}=1.097, \qquad \frac{0.0516152}{0.0055284}=9.336 .

The Chebyshev truncation is within 10%10\% of the unimprovable cubic, and it took one integral per coefficient; the minimax cubic took an iteration. That ratio — a few per cent, not a few per cent of a factor — is why truncated Chebyshev series are used in place of true minimax approximations almost everywhere in practice.

Step 5: a pure economisation, for contrast. Take p(x)=x4p(x)=x^{4}, with n=4n=4 and β4=1\beta_{4}=1. Step 1 of the economisation recipe gives β421−4=18\beta_{4}2^{1-4}=\tfrac18, and step 2 replaces x4x^{4} by

x4−18T4(x)=x4−18(8x4−8x2+1)=x2−18.x^{4}-\tfrac18T_{4}(x)=x^{4}-\tfrac18\bigl(8x^{4}-8x^{2}+1\bigr)=x^{2}-\tfrac18 .

So the best uniform cubic (indeed quadratic) approximation to x4x^{4} on [−1,1][-1,1] is x2−18x^{2}-\tfrac18, with error exactly 18=0.125\tfrac18=0.125 everywhere the error is extreme — at the five points ±1,±1/2,0\pm1,\pm1/\sqrt2,0, with alternating signs, as item 5 requires. Dropping the x4x^{4} term instead would have cost 11: economisation is eight times better here, the predicted factor 2n−1=232^{n-1}=2^{3}.

ApplicationOriginal

Spectral methods on a bounded interval. A Chebyshev spectral method represents the unknown by its values at the extrema yj=cos⁡(jπ/N)y_{j}=\cos(j\pi/N) and differentiates by transforming to coefficients, differentiating the series with Tn′=nUn−1T_{n}'=nU_{n-1}, and transforming back (Spectral methods). For a smooth solution the error falls faster than any power of NN, the same spectral accuracy a Fourier method gives on a periodic domain — and the reason is the one noted in the method: a Chebyshev series is a Fourier cosine series in disguise, and the awkward boundaries of the interval have been mapped to the smooth turning points of the cosine. The clustering of the yjy_{j} near ±1\pm1, with spacing O(N−2)O(N^{-2}) there against O(N−1)O(N^{-1}) in the middle, is not a defect but the mechanism: it resolves boundary layers and it is what defeats the Runge phenomenon that wrecks equispaced polynomial interpolation (Polynomial and spline interpolation).

Clenshaw-Curtis quadrature. Integrating the interpolant through the same points yjy_{j} gives a quadrature rule whose weights are computable by one cosine transform and whose accuracy, for smooth integrands, is in practice comparable to Gauss-Legendre of the same order despite being exact only to degree NN rather than 2N−12N-1 (Numerical quadrature). Its practical advantages are that the nodes are nested — doubling NN reuses every old evaluation — and that the weights come from a transform rather than from a root-finding problem.

Chebyshev acceleration of iterative solvers. This is the minimax theorem doing real work. In a polynomial iterative method for Ax=bAx=b, the error after kk steps is ek=Pk(A)e0e^{k}=P_{k}(A)e^{0} for some polynomial PkP_{k} with Pk(0)=1P_{k}(0)=1, so the worst-case error reduction is governed by max⁡λ∈σ(A)∣Pk(λ)∣\max_{\lambda\in\sigma(A)}\lvert P_{k}(\lambda)\rvert. Stakgold and Holst set this up as a mini-max problem over the spectral interval [λmin⁡,λmax⁡][\lambda_{\min},\lambda_{\max}] and identify the optimum as the Chebyshev polynomial shifted and scaled onto that interval and divided by its value at the origin,

Pk(λ)=Tk ⁣(λmax⁡+λmin⁡−2λλmax⁡−λmin⁡)Tk ⁣(λmax⁡+λmin⁡λmax⁡−λmin⁡).P_{k}(\lambda)=\frac{T_{k}\!\left(\dfrac{\lambda_{\max}+\lambda_{\min}-2\lambda}{\lambda_{\max}-\lambda_{\min}}\right)}{T_{k}\!\left(\dfrac{\lambda_{\max}+\lambda_{\min}}{\lambda_{\max}-\lambda_{\min}}\right)} .

The numerator is bounded by 11 on the spectrum by item 5; the denominator is TkT_{k} evaluated outside [−1,1][-1,1], where the cosh⁡\cosh continuation makes it grow geometrically. The resulting bound is

∥ek∥A≤2(κ−1κ+1)k∥e0∥A,\lVert e^{k}\rVert_{A}\le2\left(\frac{\sqrt{\kappa}-1}{\sqrt{\kappa}+1}\right)^{k}\lVert e^{0}\rVert_{A},

with κ\kappa the condition number — the square root is the whole point, and it is the same bound that governs the conjugate gradient method, which achieves it without needing to know λmin⁡\lambda_{\min} and λmax⁡\lambda_{\max} in advance (Solving linear systems numerically). Stakgold and Holst also record the degenerate case k=1k=1, where T1(x)=xT_{1}(x)=x reproduces the optimal relaxation parameter αopt=2/(λmin⁡+λmax⁡)\alpha_{\mathrm{opt}}=2/(\lambda_{\min}+\lambda_{\max}) of simple Richardson iteration (Relaxation methods for elliptic problems).

Equiripple filters. A Chebyshev type I analogue filter has magnitude response ∣H(iω)∣2=1/(1+ϵ2Tn2(ω/ω0))\lvert H(\mathrm{i}\omega)\rvert^{2}=1/\bigl(1+\epsilon^{2}T_{n}^{2}(\omega/\omega_{0})\bigr). Inside the passband TnT_{n} oscillates between ±1\pm1, so the response ripples between 11 and 1/(1+ϵ2)1/(1+\epsilon^{2}) — equal ripple, which is the minimax property transcribed into engineering — and outside, where TnT_{n} grows like cosh⁡\cosh, the response falls off faster than that of any other filter of the same order with the same passband tolerance. Trading ripple for roll-off is trading ϵ\epsilon against nn, and the Chebyshev polynomial is what makes the trade optimal.

Economised series in libraries and firmware. Before a special function is evaluated in a numerical library, its defining series is folded onto a short interval and economised exactly as in step 5 of the example. A degree-15 Taylor polynomial for sin⁡x\sin x on a reduced range typically economises to degree 7 or 9 at full double precision, which is a factor of two in the cost of every call. The saving per removed term is the factor 2n−12^{n-1} computed above, and it is why the coefficients printed in function-approximation handbooks are never the Taylor coefficients.

Definition · advanced

Legendre functions of the second kind

Adapted

The second solution Q_n of Legendre's equation for integer degree: its series definition and normalisation, the closed form P_n(x) artanh(x) minus a polynomial, the logarithmic blow-up at x = ±1 with the harmonic number H_n as its constant, and the Wronskian (1-x^2)^{-1} with P_n.

Prerequisites: Legendre polynomials, Power-series solutions at ordinary points, Reduction of order, Sturm-Liouville eigenvalue problems, Structure of linear equations and the Wronskian

Definition(The second solution for integer degree)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.1 and §18.1.1, eqs. (18.2)–(18.7), pp. 578–579; Arfken, Weber & Harris (2013), §15.6, eqs. (15.153)–(15.156), pp. 766–767

Throughout, nn is a non-negative integer, PnP_n is the Legendre polynomial of Legendre polynomials, and Legendre's equation is

(1−x2)y′′−2xy′+ℓ(ℓ+1)y=0,equivalently[(1−x2)y′]′+ℓ(ℓ+1)y=0.(L)(1-x^{2})y''-2xy'+\ell(\ell+1)y=0, \qquad\text{equivalently}\qquad \bigl[(1-x^{2})y'\bigr]'+\ell(\ell+1)y=0 . \tag{L}

Write

L(x):=12ln⁡1+x1−x=artanh⁡x(−1<x<1),Hn:=∑m=1n1m(H0=0),L(x):=\tfrac12\ln\frac{1+x}{1-x}=\operatorname{artanh}x \quad(-1<x<1), \qquad H_{n}:=\sum_{m=1}^{n}\frac1m \quad (H_{0}=0),

so that L′(x)=(1−x2)−1L'(x)=(1-x^{2})^{-1}.

In (L) the origin is an ordinary point whose nearest singular neighbours are x=±1x=\pm1, so there are two solutions in powers of xx, each with radius of convergence at least 11 — exactly 11 for general ℓ\ell, which is the case Riley, Hobson and Bence record. Substituting y=∑k≥0akxky=\sum_{k\ge0}a_{k}x^{k} gives the two-step recurrence

ak+2=k(k+1)−ℓ(ℓ+1)(k+1)(k+2) ak,a_{k+2}=\frac{k(k+1)-\ell(\ell+1)}{(k+1)(k+2)}\,a_{k},

and the choices (a0,a1)=(1,0)(a_{0},a_{1})=(1,0) and (0,1)(0,1) produce an even solution yey_{\mathrm{e}} and an odd solution yoy_{\mathrm{o}} respectively. When ℓ=n\ell=n the numerator vanishes at k=nk=n, so exactly one of the two series terminates — yey_{\mathrm{e}} if nn is even, yoy_{\mathrm{o}} if nn is odd — and that one is the Legendre polynomial up to a factor. The other series does not terminate, and converges only on (−1,1)(-1,1).

Riley, Hobson and Bence take the non-terminating series as the second solution and normalise it: the Legendre function of the second kind of degree nn is

Qn:=αn yo  (n even),Qn:=βn ye  (n odd),Q_{n}:=\alpha_{n}\,y_{\mathrm{o}}\ \ (n\text{ even}), \qquad Q_{n}:=\beta_{n}\,y_{\mathrm{e}}\ \ (n\text{ odd}),αn=(−1)n/2 2n[(n/2)!]2n!,βn=(−1)(n+1)/2 2n−1{[(n−1)/2]!}2n!.\alpha_{n}=\frac{(-1)^{n/2}\,2^{n}\bigl[(n/2)!\bigr]^{2}}{n!}, \qquad \beta_{n}=\frac{(-1)^{(n+1)/2}\,2^{n-1}\bigl\{[(n-1)/2]!\bigr\}^{2}}{n!} .

Their stated reason for these particular constants is that the resulting QnQ_{n} then satisfy, verbatim, the recurrences already established for PnP_{n}. Part 6 of the statement below identifies αn\alpha_{n} and βn\beta_{n} in closed form and the proof shows what forces them.

Arfken, Weber and Harris define the same functions by reduction of order (Reduction of order) instead. In the self-adjoint form of (L) the leading coefficient is p=1−x2p=1-x^{2}, so a second solution is

Qn(x)=Pn(x)∫xdt(1−t2) Pn(t)2,(R)Q_{n}(x)=P_{n}(x)\int^{x}\frac{\mathrm{d}t}{(1-t^{2})\,P_{n}(t)^{2}} , \tag{R}

and they remark that (R) fixes both the scale and the functional form of QnQ_{n}. By construction QnQ_{n} is of the opposite parity to PnP_{n}, since it is the series that PnP_{n} is not.

Outside the interval. Formulas for QnQ_{n} on (−1,1)(-1,1) contain ln⁡1+x1−x\ln\frac{1+x}{1-x}. For real x>1x>1 this is complex, and Arfken, Weber and Harris replace it by ln⁡x+1x−1\ln\frac{x+1}{x-1}, discarding the resulting constant multiple ±iπPn\pm\mathrm{i}\pi P_{n} of the first solution. In the complex plane the logarithm forces a branch cut, traditionally [−1,1][-1,1] (Branch points and branch cuts); the functions on (−1,1)(-1,1) are the average of the two sides of that cut. Hassani arrives at QλQ_{\lambda} for arbitrary complex degree by specialising the second-kind Jacobi functions Qλ(α,β)Q^{(\alpha,\beta)}_{\lambda} to α=β=0\alpha=\beta=0,

Qλ(z)=2λ Γ2(λ+1)Γ(2λ+2) (z−1)λ+1  F(λ+1,λ+1;2λ+2;21−z),Q_{\lambda}(z)=\frac{2^{\lambda}\,\Gamma^{2}(\lambda+1)}{\Gamma(2\lambda+2)\,(z-1)^{\lambda+1}}\; F\Bigl(\lambda+1,\lambda+1;2\lambda+2;\frac{2}{1-z}\Bigr),

with FF the hypergeometric function of The hypergeometric function; at λ=n\lambda=n this agrees with the x>1x>1 convention above.

Definition(Closed form, singularities and Wronskian)AdaptedAdapted from Riley, Hobson & Bence (2006), §18.1.1, eqs. (18.5)–(18.8), pp. 579–580, and §18.1.2, eqs. (18.23)–(18.27), p. 586; Arfken, Weber & Harris (2013), §15.6, eqs. (15.155)–(15.161), pp. 766–769, and §7.6, Exercise 7.6.9, p. 371

Let n≥0n\ge0 be an integer. Parts 1–6 concern the interval −1<x<1-1<x<1; part 7 steps outside it, to x∈(1,∞)x\in(1,\infty). Then:

  1. Closed form. There is a polynomial RnR_{n} of degree exactly n−1n-1 (with R0≡0R_{0}\equiv0) such that Qn(x)=Pn(x) L(x)−Rn(x),Rn(x)=∑m=1n1m Pm−1(x) Pn−m(x).Q_{n}(x)=P_{n}(x)\,L(x)-R_{n}(x), \qquad R_{n}(x)=\sum_{m=1}^{n}\frac1m\,P_{m-1}(x)\,P_{n-m}(x) . In particular Q0=L,Q1=xL−1,Q2=P2L−32x,Q3=P3L−52x2+23.Q_{0}=L,\qquad Q_{1}=xL-1,\qquad Q_{2}=P_{2}L-\tfrac32x,\qquad Q_{3}=P_{3}L-\tfrac52x^{2}+\tfrac23 . Arfken, Weber and Harris write the polynomial part as Rn=2n−11⋅nPn−1+2n−53(n−1)Pn−3+⋯R_{n}=\frac{2n-1}{1\cdot n}P_{n-1}+\frac{2n-5}{3(n-1)}P_{n-3}+\cdots, which is the same polynomial resolved in the Legendre basis.
  2. Logarithmic singularity. Rn(1)=HnR_{n}(1)=H_{n}, and as x→1−x\to1^{-} Qn(x)=12ln⁡21−x−Hn+O((1−x)ln⁡(1−x)).Q_{n}(x)=\tfrac12\ln\frac{2}{1-x}-H_{n}+O\bigl((1-x)\ln(1-x)\bigr) . By the parity in part 4 the behaviour at the other end is Qn(x)=(−1)n+1[12ln⁡21+x−Hn]+O((1+x)ln⁡(1+x))Q_{n}(x)=(-1)^{n+1}\bigl[\tfrac12\ln\frac{2}{1+x}-H_{n}\bigr]+O\bigl((1+x)\ln(1+x)\bigr) as x→−1+x\to-1^{+}. So QnQ_{n} is unbounded at both ends for every nn — Arfken, Weber and Harris record this as Qn(1)=∞Q_{n}(1)=\infty — while remaining square integrable on (−1,1)(-1,1), the singularity being only logarithmic.
  3. Recurrences. The QnQ_{n} satisfy the same relations as the PnP_{n}: (n+1)Qn+1=(2n+1)x Qn−n Qn−1,(2n+1)Qn=Qn+1′−Qn−1′,(n+1)Q_{n+1}=(2n+1)x\,Q_{n}-n\,Q_{n-1}, \qquad (2n+1)Q_{n}=Q_{n+1}'-Q_{n-1}' , together with (1−x2)Qn′=n(Qn−1−xQn)(1-x^{2})Q_{n}'=n(Q_{n-1}-xQ_{n}) and the rest of the family in Legendre polynomials.
  4. Parity and values at the origin. Qn(−x)=(−1)n+1Qn(x)Q_{n}(-x)=(-1)^{n+1}Q_{n}(x): the QnQ_{n} of even degree are odd functions and conversely. Hence Q2p(0)=0Q_{2p}(0)=0, while Q2p+1(0)=(−1)p+1(2p)!!(2p+1)!!.Q_{2p+1}(0)=(-1)^{p+1}\frac{(2p)!!}{(2p+1)!!} . There is no singularity at x=0x=0, which is an ordinary point of (L).
  5. Wronskian. On (−1,1)(-1,1), W[Pn,Qn](x)=Pn(x)Qn′(x)−Pn′(x)Qn(x)=11−x2.W[P_{n},Q_{n}](x)=P_{n}(x)Q_{n}'(x)-P_{n}'(x)Q_{n}(x)=\frac{1}{1-x^{2}} . Arfken, Weber and Harris state this in the form An/(1−x2)A_{n}/(1-x^{2}) with AnA_{n} independent of xx; with the normalisation above An=1A_{n}=1 for every nn. The Wronskian never vanishes, so {Pn,Qn}\{P_{n},Q_{n}\} is a basis of the solution space of (L) at ℓ=n\ell=n: every solution on (−1,1)(-1,1) is c1Pn+c2Qnc_{1}P_{n}+c_{2}Q_{n}. The companion contiguous relation of Arfken, Weber and Harris is n[PnQn−1−Pn−1Qn]=1n\bigl[P_{n}Q_{n-1}-P_{n-1}Q_{n}\bigr]=1.
  6. The normalising constants. αn=1/Pn(0)\alpha_{n}=1/P_{n}(0) for even nn and βn=−1/Pn′(0)\beta_{n}=-1/P_{n}'(0) for odd nn; these are precisely the values that make An=1A_{n}=1 in part 5.
  7. Outside the interval. With the convention L↦12ln⁡x+1x−1L\mapsto\tfrac12\ln\frac{x+1}{x-1} for x>1x>1, each QnQ_{n} is finite and positive on (1,∞)(1,\infty), decreasing to 00; Arfken, Weber and Harris obtain the expansion in inverse powers of xx whose leading term is Qn(x)∼n!(2n+1)!! x−n−1(x→∞).Q_{n}(x)\sim\frac{n!}{(2n+1)!!}\,x^{-n-1} \qquad (x\to\infty).
Idea of the proofOriginal

Everything follows from one generating function. Put G(x,h)=(1−2xh+h2)−1/2=∑nPnhnG(x,h)=(1-2xh+h^{2})^{-1/2}=\sum_{n}P_{n}h^{n} and look for a second family packed the same way. Two observations do it.

First, PnLP_{n}L almost solves Legendre's equation: because (1−x2)L′=1(1-x^{2})L'=1, substituting PnLP_{n}L into (L) leaves the clean residue 2Pn′2P_{n}' and nothing else. So the second solution must be PnLP_{n}L corrected by whatever cancels 2Pn′2P_{n}', and because 2Pn′2P_{n}' is a polynomial, the correction should be one too.

Second, the correction can be read off in one stroke rather than degree by degree. The generating function of the corrections turns out to be GG multiplied by A(x,h)=∫0hG(x,s) dsA(x,h)=\int_{0}^{h}G(x,s)\,\mathrm{d}s, an integral that is elementary — complete the square under the square root — so the whole family ∑Qnhn=G⋅[L−A]\sum Q_{n}h^{n}=G\cdot[L-A] is available in closed form. Differentiating that product in xx and in hh gives the two recurrences exactly as for the PnP_{n}; reading off the coefficient of hnh^{n} gives the closed form with Rn=∑m1mPm−1Pn−mR_{n}=\sum_{m}\frac1mP_{m-1}P_{n-m}; and setting x=1x=1 in RnR_{n} gives ∑m1m=Hn\sum_{m}\frac1m=H_{n} because every Pj(1)=1P_{j}(1)=1. That last line is the whole content of the logarithmic singularity.

The Wronskian is separate and cheaper: (1−x2)W(1-x^{2})W has zero derivative for any two solutions of (L), so it is a constant, and the constant is best evaluated by letting x→1−x\to1^{-}, where the polynomial pieces are killed by the factor 1−x21-x^{2} and only Pn(1)2=1P_{n}(1)^{2}=1 survives.

Proof(The generating function of the second solutions)Original

Write G=(1−2xh+h2)−1/2G=(1-2xh+h^{2})^{-1/2}, and let

M[Ψ]:=∂x[(1−x2)Ψx]+∂h[h2Ψh].\mathcal{M}[\Psi]:=\partial_{x}\bigl[(1-x^{2})\Psi_{x}\bigr]+\partial_{h}\bigl[h^{2}\Psi_{h}\bigr] .

If Ψ=∑n≥0ψn(x)hn\Psi=\sum_{n\ge0}\psi_{n}(x)h^{n} then ∂h[h2Ψh]=∑nn(n+1)ψnhn\partial_{h}[h^{2}\Psi_{h}]=\sum_{n}n(n+1)\psi_{n}h^{n}, so

M[Ψ]=∑n≥0{[(1−x2)ψn′]′+n(n+1)ψn}hn:(1)\mathcal{M}[\Psi]=\sum_{n\ge0}\Bigl\{\bigl[(1-x^{2})\psi_{n}'\bigr]'+n(n+1)\psi_{n}\Bigr\}h^{n} : \tag{1}

M[Ψ]=0\mathcal{M}[\Psi]=0 says precisely that every coefficient ψn\psi_{n} solves (L) with ℓ=n\ell=n. For GG itself this is the statement proved in Legendre polynomials, so we may use M[G]=0\mathcal{M}[G]=0. We shall also use the two elementary derivatives

Gx=hG3,Gh=(x−h)G3,(2)G_{x}=hG^{3}, \qquad G_{h}=(x-h)G^{3}, \tag{2}

each obtained by differentiating (1−2xh+h2)−1/2(1-2xh+h^{2})^{-1/2} once.

Step 1: the candidate. Set

A(x,h):=∫0hds1−2xs+s2,Λ:=L−A,Φ:=G Λ.A(x,h):=\int_{0}^{h}\frac{\mathrm{d}s}{\sqrt{1-2xs+s^{2}}}, \qquad \Lambda:=L-A, \qquad \Phi:=G\,\Lambda .

Completing the square, 1−2xs+s2=(s−x)2+(1−x2)1-2xs+s^{2}=(s-x)^{2}+(1-x^{2}), so the integral is a standard inverse hyperbolic sine and

A(x,h)=[ln⁡(s−x+1−2xs+s2)]s=0s=h=ln⁡h−x+1−2xh+h21−x,A(x,h)=\Bigl[\ln\bigl(s-x+\sqrt{1-2xs+s^{2}}\bigr)\Bigr]_{s=0}^{s=h} =\ln\frac{h-x+\sqrt{1-2xh+h^{2}}}{1-x},

whence, using L=ln⁡(1+x)/(1−x)L=\ln\sqrt{(1+x)/(1-x)},

Λ(x,h)=ln⁡1−x2 h−x+1−2xh+h2 .(3)\Lambda(x,h)=\ln\frac{\sqrt{1-x^{2}}}{\,h-x+\sqrt{1-2xh+h^{2}}\,} . \tag{3}

Step 2: Φ\Phi is analytic in hh on the unit disc. For ∣x∣<1\lvert x\rvert<1 the roots of 1−2xh+h21-2xh+h^{2} in hh are x±i1−x2x\pm\mathrm{i}\sqrt{1-x^{2}}, both of modulus 11, so 1−2xh+h2\sqrt{1-2xh+h^{2}} has an analytic branch on ∣h∣<1\lvert h\rvert<1 equal to 11 at h=0h=0. The denominator in (3) cannot vanish there: h−x+1−2xh+h2=0h-x+\sqrt{1-2xh+h^{2}}=0 would force 1−2xh+h2=(x−h)21-2xh+h^{2}=(x-h)^{2}, that is x2=1x^{2}=1. Being non-vanishing on a disc and equal to 1−x>01-x>0 at h=0h=0, it admits a single-valued analytic logarithm. Hence Φ=∑n≥0qn(x)hn\Phi=\sum_{n\ge0}q_{n}(x)h^{n} with radius of convergence at least 11.

Step 3: the two derivatives of Λ\Lambda. Differentiating (3) in hh recovers the integrand, Λh=−Ah=−G\Lambda_{h}=-A_{h}=-G. For the xx-derivative, write u=1−2xh+h2u=1-2xh+h^{2} and differentiate the logarithm:

Λx=−x1−x2+u+hu (h−x+u).\Lambda_{x}=\frac{-x}{1-x^{2}}+\frac{\sqrt{u}+h}{\sqrt{u}\,\bigl(h-x+\sqrt{u}\bigr)} .

Multiply numerator and denominator of the second term by u−h+x\sqrt{u}-h+x. The denominator becomes u [u−(h−x)2]=u (1−x2)\sqrt{u}\,[u-(h-x)^{2}]=\sqrt{u}\,(1-x^{2}), and the numerator becomes u+xu−h2+hx=1−xh+xuu+x\sqrt{u}-h^{2}+hx=1-xh+x\sqrt{u}. The two terms in xx cancel and

Λh=−G,Λx=(1−xh) G1−x2.(4)\Lambda_{h}=-G, \qquad \Lambda_{x}=\frac{(1-xh)\,G}{1-x^{2}} . \tag{4}

Step 4: M[Φ]=0\mathcal{M}[\Phi]=0. For any product GΛG\Lambda, expanding each derivative by the Leibniz rule and collecting,

M[GΛ]=Λ M[G]+2{(1−x2)GxΛx+h2GhΛh}+G M[Λ].\mathcal{M}[G\Lambda] =\Lambda\,\mathcal{M}[G] +2\bigl\{(1-x^{2})G_{x}\Lambda_{x}+h^{2}G_{h}\Lambda_{h}\bigr\} +G\,\mathcal{M}[\Lambda] .

The first term vanishes. For the second, (2) and (4) give

(1−x2)GxΛx=h(1−xh)G4,h2GhΛh=−h2(x−h)G4,(1-x^{2})G_{x}\Lambda_{x}=h(1-xh)G^{4}, \qquad h^{2}G_{h}\Lambda_{h}=-h^{2}(x-h)G^{4},

and since (1−xh)−h(x−h)=1−2xh+h2=G−2(1-xh)-h(x-h)=1-2xh+h^{2}=G^{-2}, their sum is hG2hG^{2}; the second term is therefore 2hG22hG^{2}. For the third, (4) gives (1−x2)Λx=(1−xh)G(1-x^{2})\Lambda_{x}=(1-xh)G and h2Λh=−h2Gh^{2}\Lambda_{h}=-h^{2}G, so

M[Λ]=∂x[(1−xh)G]−∂h[h2G]=−hG+(1−xh)hG3−2hG−h2(x−h)G3=−3hG+hG3G−2=−2hG.\mathcal{M}[\Lambda]=\partial_{x}\bigl[(1-xh)G\bigr]-\partial_{h}\bigl[h^{2}G\bigr] =-hG+(1-xh)hG^{3}-2hG-h^{2}(x-h)G^{3} =-3hG+hG^{3}G^{-2}=-2hG .

Hence G M[Λ]=−2hG2G\,\mathcal{M}[\Lambda]=-2hG^{2}, the two surviving terms cancel, and M[Φ]=0\mathcal{M}[\Phi]=0. By (1), every coefficient qnq_{n} solves (L) with ℓ=n\ell=n. □\square

Proof(Closed form, recurrences, parity and the singularity)Original

We now read the properties off Φ=GΛ=LG−GA\Phi=G\Lambda=LG-GA.

Closed form (part 1). Integrating G=∑k≥0PkhkG=\sum_{k\ge0}P_{k}h^{k} term by term from 00 to hh,

A(x,h)=∑k≥0Pk(x)k+1hk+1=∑m≥1Pm−1(x)mhm.A(x,h)=\sum_{k\ge0}\frac{P_{k}(x)}{k+1}h^{k+1}=\sum_{m\ge1}\frac{P_{m-1}(x)}{m}h^{m}.

Multiplying this by GG and taking the coefficient of hnh^{n} in Φ=LG−GA\Phi=LG-GA,

qn=PnL−Rn,Rn=∑m=1n1mPm−1Pn−m(R0=0).q_{n}=P_{n}L-R_{n}, \qquad R_{n}=\sum_{m=1}^{n}\frac1m P_{m-1}P_{n-m} \quad (R_{0}=0).

Each summand has degree (m−1)+(n−m)=n−1(m-1)+(n-m)=n-1, and the leading coefficients of Legendre polynomials are positive, so the leading coefficients cannot cancel: RnR_{n} is a polynomial of degree exactly n−1n-1 for n≥1n\ge1. Taking n=0,1,2,3n=0,1,2,3 gives R0=0R_{0}=0, R1=1R_{1}=1, R2=P0P1+12P1P0=32xR_{2}=P_{0}P_{1}+\tfrac12P_{1}P_{0}=\tfrac32x and R3=P2+12x2+13P2=52x2−23R_{3}=P_{2}+\tfrac12x^{2}+\tfrac13P_{2}=\tfrac52x^{2}-\tfrac23, which are the four displayed cases.

The three-term recurrence (part 3). From (2) and (4),

Φh=GhΛ+GΛh=(x−h)G3Λ−G2,\Phi_{h}=G_{h}\Lambda+G\Lambda_{h}=(x-h)G^{3}\Lambda-G^{2},

and multiplying by 1−2xh+h2=G−21-2xh+h^{2}=G^{-2},

(1−2xh+h2) Φh=(x−h)Φ−1.(5)(1-2xh+h^{2})\,\Phi_{h}=(x-h)\Phi-1 . \tag{5}

This is the identity that produces the recurrence for the PnP_{n}, except for the extra constant −1-1 on the right, which only affects the coefficient of h0h^{0}. Comparing coefficients of hnh^{n} with n≥1n\ge1,

(n+1)qn+1−2xn qn+(n−1)qn−1=x qn−qn−1,(n+1)q_{n+1}-2xn\,q_{n}+(n-1)q_{n-1}=x\,q_{n}-q_{n-1},

which rearranges to (n+1)qn+1=(2n+1)xqn−nqn−1(n+1)q_{n+1}=(2n+1)xq_{n}-nq_{n-1}. At n=0n=0 the same comparison gives q1=xq0−1q_{1}=xq_{0}-1.

The derivative recurrence (part 3). In the same way, Φx=GxΛ+GΛx=hG3Λ+(1−xh)G2/(1−x2)\Phi_{x}=G_{x}\Lambda+G\Lambda_{x}=hG^{3}\Lambda+(1-xh)G^{2}/(1-x^{2}), so

(1−2xh+h2) Φx=h Φ+1−xh1−x2.(6)(1-2xh+h^{2})\,\Phi_{x}=h\,\Phi+\frac{1-xh}{1-x^{2}} . \tag{6}

For n≥2n\ge2 the coefficient of hnh^{n} reads qn′−2xqn−1′+qn−2′=qn−1q_{n}'-2xq_{n-1}'+q_{n-2}'=q_{n-1}, which is the relation the PnP_{n} satisfy; eliminating between it and the three-term recurrence exactly as in Legendre polynomials yields (2n+1)qn=qn+1′−qn−1′(2n+1)q_{n}=q_{n+1}'-q_{n-1}' and (1−x2)qn′=n(qn−1−xqn)(1-x^{2})q_{n}'=n(q_{n-1}-xq_{n}).

Parity and the origin (part 4). LL is odd and Pj(−x)=(−1)jPj(x)P_{j}(-x)=(-1)^{j}P_{j}(x), so every summand of RnR_{n} has parity (−1)n−1(-1)^{n-1} and

qn(−x)=(−1)nPn(x)(−L(x))−(−1)n−1Rn(x)=(−1)n+1qn(x).q_{n}(-x)=(-1)^{n}P_{n}(x)\bigl(-L(x)\bigr)-(-1)^{n-1}R_{n}(x)=(-1)^{n+1}q_{n}(x) .

For even nn this makes qnq_{n} odd, so qn(0)=0q_{n}(0)=0. The odd indices are linked to one another by taking nn even in the three-term recurrence: with n=2pn=2p and x=0x=0 the middle term drops out and (2p+1) q2p+1(0)=−2p q2p−1(0)(2p+1)\,q_{2p+1}(0)=-2p\,q_{2p-1}(0), so from q1(0)=−1q_{1}(0)=-1,

q2p+1(0)=(−1)p+12⋅4⋯(2p)3⋅5⋯(2p+1)=(−1)p+1(2p)!!(2p+1)!!,q_{2p+1}(0)=(-1)^{p+1}\frac{2\cdot4\cdots(2p)}{3\cdot5\cdots(2p+1)} =(-1)^{p+1}\frac{(2p)!!}{(2p+1)!!},

which is Arfken, Weber and Harris's value. (For p=1p=1 and p=2p=2 this is 23\tfrac23 and −815-\tfrac8{15}.)

The logarithmic singularity (part 2). Since Pj(1)=1P_{j}(1)=1 for every jj,

Rn(1)=∑m=1n1m⋅1⋅1=Hn,R_{n}(1)=\sum_{m=1}^{n}\frac1m\cdot1\cdot1=H_{n} ,

with no computation at all. Now split LL as L(x)=12ln⁡21−x+12ln⁡1+x2L(x)=\tfrac12\ln\frac{2}{1-x}+\tfrac12\ln\frac{1+x}{2}. As x→1−x\to1^{-} the second piece is O(1−x)O(1-x), and Pn(x)−1=O(1−x)P_{n}(x)-1=O(1-x) because PnP_{n} is a polynomial with Pn(1)=1P_{n}(1)=1; hence

Pn(x)L(x)=12ln⁡21−x+O((1−x)ln⁡(1−x)).P_{n}(x)L(x)=\tfrac12\ln\frac{2}{1-x}+O\bigl((1-x)\ln(1-x)\bigr).

Likewise Rn(x)=Hn+O(1−x)R_{n}(x)=H_{n}+O(1-x). Subtracting gives part 2, and the statement at x=−1x=-1 follows from the parity. In particular ∣qn(x)∣→∞\lvert q_{n}(x)\rvert\to\infty at both ends, while ∫−11qn2\int_{-1}^{1}q_{n}^{2} converges because ln⁡2\ln^{2} is integrable. □\square

Proof(The Wronskian, and the identification with the normalised series)AdaptedAdapted from Arfken, Weber & Harris (2013), §7.6, Exercises 7.6.9 and 7.6.11(a), p. 371; Riley, Hobson & Bence (2006), §18.1.1, eqs. (18.5)–(18.6), p. 579

Step 1: (1−x2)W(1-x^{2})W is constant. Let u,vu,v be any two solutions of (L) at ℓ=n\ell=n and W=uv′−u′vW=uv'-u'v. Then W′=uv′′−u′′vW'=uv''-u''v and, using (L) for each of uu and vv to replace (1−x2)v′′(1-x^{2})v'' and (1−x2)u′′(1-x^{2})u'',

(1−x2)W′=u[2xv′−n(n+1)v]−[2xu′−n(n+1)u]v=2xW.(1-x^{2})W'=u\bigl[2xv'-n(n+1)v\bigr]-\bigl[2xu'-n(n+1)u\bigr]v=2xW .

Hence ddx[(1−x2)W]=(1−x2)W′−2xW=0\frac{\mathrm{d}}{\mathrm{d}x}\bigl[(1-x^{2})W\bigr]=(1-x^{2})W'-2xW=0. This is Arfken, Weber and Harris's observation that pWpW is constant for a self-adjoint equation, here with p=1−x2p=1-x^{2}; their Exercise 7.6.9 states the Legendre case as W=An/(1−x2)W=A_{n}/(1-x^{2}).

Step 2: the constant is 11. Take u=Pnu=P_{n}, v=qn=PnL−Rnv=q_{n}=P_{n}L-R_{n}. Using (1−x2)L′=1(1-x^{2})L'=1,

(1−x2)W=(1−x2)[Pn(Pn′L+PnL′−Rn′)−Pn′(PnL−Rn)]=Pn2+(1−x2)[Pn′Rn−PnRn′],(1-x^{2})W=(1-x^{2})\bigl[P_{n}\bigl(P_{n}'L+P_{n}L'-R_{n}'\bigr)-P_{n}'\bigl(P_{n}L-R_{n}\bigr)\bigr] =P_{n}^{2}+(1-x^{2})\bigl[P_{n}'R_{n}-P_{n}R_{n}'\bigr],

the terms carrying LL having cancelled. The right-hand side is a polynomial, so we may evaluate it at x=1x=1: the bracket is finite there and is multiplied by 1−x2=01-x^{2}=0, while Pn(1)2=1P_{n}(1)^{2}=1. Hence An=1A_{n}=1 for every nn and

Pnqn′−Pn′qn=11−x2.P_{n}q_{n}'-P_{n}'q_{n}=\frac{1}{1-x^{2}} .

Since this never vanishes on (−1,1)(-1,1), PnP_{n} and qnq_{n} are linearly independent there and span the solution space (Structure of linear equations and the Wronskian), giving part 5. The contiguous relation n[PnQn−1−Pn−1Qn]=1n[P_{n}Q_{n-1}-P_{n-1}Q_{n}]=1 of Arfken, Weber and Harris is the same statement rewritten with the recurrences of part 3.

Step 3: qn=Qnq_{n}=Q_{n}. Both qnq_{n} and QnQ_{n} solve (L) at ℓ=n\ell=n, and both have parity (−1)n+1(-1)^{n+1} — for QnQ_{n} by its definition as the non-terminating series, for qnq_{n} by part 4. Since PnP_{n} has the opposite parity, a solution of definite parity (−1)n+1(-1)^{n+1} is unique up to scale, so qnq_{n} and QnQ_{n} are proportional and it is enough to match one value at x=0x=0.

Let nn be even. Then Qn=αnyoQ_{n}=\alpha_{n}y_{\mathrm{o}} with yo(0)=0y_{\mathrm{o}}(0)=0, yo′(0)=1y_{\mathrm{o}}'(0)=1, so Qn(0)=0Q_{n}(0)=0 and Qn′(0)=αnQ_{n}'(0)=\alpha_{n}; also Pn′(0)=0P_{n}'(0)=0. Putting x=0x=0 in Step 2 for the pair Pn,QnP_{n},Q_{n} — whose Wronskian constant we may call AnA_{n} — gives An=Pn(0)αnA_{n}=P_{n}(0)\alpha_{n}. Now Legendre polynomials gives P2m(0)=(−1)m(2m)!/[22m(m!)2]P_{2m}(0)=(-1)^{m}(2m)!/\bigl[2^{2m}(m!)^{2}\bigr], that is Pn(0)=(−1)n/2n!/(2n[(n/2)!]2)P_{n}(0)=(-1)^{n/2}n!/\bigl(2^{n}[(n/2)!]^{2}\bigr), which is exactly 1/αn1/\alpha_{n}. So An=1A_{n}=1: Riley, Hobson and Bence's constant αn\alpha_{n} is the one that normalises the Wronskian, αn=1/Pn(0)\alpha_{n}=1/P_{n}(0), and qnq_{n} and QnQ_{n} — proportional, with the same Wronskian against PnP_{n} — coincide.

Let nn be odd. Then Qn=βnyeQ_{n}=\beta_{n}y_{\mathrm{e}} with ye(0)=1y_{\mathrm{e}}(0)=1, ye′(0)=0y_{\mathrm{e}}'(0)=0, so Qn(0)=βnQ_{n}(0)=\beta_{n}, Qn′(0)=0Q_{n}'(0)=0, and Pn(0)=0P_{n}(0)=0; the Wronskian at the origin is −Pn′(0)βn-P_{n}'(0)\beta_{n}. Setting x=0x=0 in the derivative recurrence (1−x2)Pn′=n(Pn−1−xPn)(1-x^{2})P_{n}'=n(P_{n-1}-xP_{n}) gives Pn′(0)=nPn−1(0)P_{n}'(0)=nP_{n-1}(0), so with n=2p+1n=2p+1,

Pn′(0)=(2p+1)(−1)p(2p)!22p(p!)2=(−1)p(2p+1)!22p(p!)2=−1βn,P_{n}'(0)=(2p+1)\frac{(-1)^{p}(2p)!}{2^{2p}(p!)^{2}}=\frac{(-1)^{p}(2p+1)!}{2^{2p}(p!)^{2}}=-\frac{1}{\beta_{n}} ,

the last equality being the definition of βn\beta_{n} with (n−1)/2=p(n-1)/2=p and (−1)(n+1)/2=(−1)p+1(-1)^{(n+1)/2}=(-1)^{p+1}. Hence the Wronskian constant is again 11 and qn=Qnq_{n}=Q_{n}. This proves part 6 and completes the identification. □\square

Remark(Where the hypotheses are used)Original

nn an integer. Only then does one of the two series terminate, so that "the second solution" is unambiguous and the first is a polynomial. For non-integer ℓ\ell both series are infinite and both blow up at x=±1x=\pm1; the functions PλP_{\lambda}, QλQ_{\lambda} of Hassani's hypergeometric formulas still exist, but nothing in this unit's proofs survives, because RnR_{n} was a polynomial precisely because GG generates polynomials.

∣x∣<1\lvert x\rvert<1. Used three times. The series defining QnQ_{n} converges only there. The generating function argument needs ∣x∣≤1\lvert x\rvert\le1 so that the singularities of GG in hh sit on ∣h∣=1\lvert h\rvert=1, and needs ∣x∣<1\lvert x\rvert<1 strictly so that h−x+uh-x+\sqrt{u} does not vanish — at x=±1x=\pm1 it does, which is the analytic shadow of the singularity of QnQ_{n} at the ends. And L=artanh⁡xL=\operatorname{artanh}x is real only there; beyond x=1x=1 one must switch to 12ln⁡x+1x−1\tfrac12\ln\frac{x+1}{x-1}, which differs by 12ln⁡(−1)\tfrac12\ln(-1) and so changes the solution by a multiple of PnP_{n} — a real change of function, not a change of notation.

Regularity at the ends is what selects PnP_{n}. Part 2 is the general-nn version of the n=0n=0 calculation in Legendre polynomials: the obstruction is not a failure of square integrability, since ∫−11Qn2\int_{-1}^{1}Q_{n}^{2} converges for every nn, but the survival of the boundary form p (uv′−u′v)p\,(uv'-u'v), with p=1−x2p=1-x^{2}, at the endpoints x=±1x=\pm1: the Sturm-Liouville argument needs that form to vanish at each end for the operator to be symmetric. Part 5 makes that explicit and quantitative: for u=Pnu=P_{n}, v=Qnv=Q_{n} the boundary form equals 11 at each endpoint separately, as far from zero as it could be. That single number is why a problem on the full sphere must set c2=0c_{2}=0 in y=c1Pn+c2Qny=c_{1}P_{n}+c_{2}Q_{n}, and hence why the separation constant is quantised.

Where the normalisation enters. Nothing in parts 1–4 depends on the scale of QnQ_{n} except through the starting values Q0=LQ_{0}=L, Q1=xL−1Q_{1}=xL-1. The content of the constants αn,βn\alpha_{n},\beta_{n} is exactly part 6: they are the unique choice making the Wronskian constant equal 11 for every degree, and therefore the unique choice under which PnP_{n} and QnQ_{n} obey one and the same recurrence. A different scaling for each nn — for instance normalising QnQ_{n} to have unit value at x=0x=0 — would destroy part 3 and with it every computational use of these functions.

Two silent points in the reduction-of-order formula (R). Its lower limit is left open, and a different lower limit changes the answer by a constant multiple of PnP_{n}; what pins that multiple down is the parity, since PnP_{n} and QnQ_{n} have opposite parities and so no multiple of PnP_{n} can be added without destroying it. Also, PnP_{n} has nn simple zeros inside (−1,1)(-1,1), so the integrand of (R) is not locally integrable at them; QnQ_{n} itself nevertheless has a finite limit at each, because the double pole of Pn−2P_{n}^{-2} integrates to a simple pole in the quotient Qn/PnQ_{n}/P_{n} — which is what (R) computes — and the prefactor PnP_{n} cancels exactly that pole. It is the quotient, not QnQ_{n}, that blows up there. We have taken the generating-function route instead of (R) partly for this reason: it never divides by PnP_{n} and so never meets either difficulty. Arfken, Weber and Harris state that (R) fixes the scale and the form of QnQ_{n} without comment on either point.

Rn(1)=HnR_{n}(1)=H_{n} needs Pj(1)=1P_{j}(1)=1. This is the standardisation of the Legendre polynomials, not a coincidence. With the monic standardisation of General properties of orthogonal polynomials the constant in part 2 would be a different and far less memorable number.

Degree exactly n−1n-1. The claim that the leading coefficients in Rn=∑m1mPm−1Pn−mR_{n}=\sum_{m}\frac1mP_{m-1}P_{n-m} cannot cancel uses that all of them are positive, which is read off Rodrigues' formula. Without that, RnR_{n} could in principle drop degree, and the asymptotic statements in part 7 would need separate proof.

Method(Computing with QnQ_{n})Original
  1. On (−1,1)(-1,1): set L=artanh⁡xL=\operatorname{artanh}x, take Q0=LQ_{0}=L and Q1=xL−1Q_{1}=xL-1, and iterate upward, Qn+1=(2n+1)x Qn−n Qn−1n+1.Q_{n+1}=\frac{(2n+1)x\,Q_{n}-n\,Q_{n-1}}{n+1} . This is stable, and for the same reason as for PnP_{n}: on (−1,1)(-1,1) the two solutions of the recurrence are of comparable size (both decay only like n−1/2n^{-1/2} with oscillating sign), so no step has to recover a small quantity from large ones. Near the ends, cancel the singularity by hand — compute LL and the polynomial RnR_{n} separately and subtract — rather than letting LL and PnLP_{n}L both overflow.
  2. On (1,∞)(1,\infty): do not iterate upward. There PnP_{n} grows like (x+x2−1)n(x+\sqrt{x^{2}-1})^{n} while QnQ_{n} decays like its reciprocal, so an upward sweep for QnQ_{n} is contaminated by the dominant solution and loses everything. At x=5x=5 in double precision the upward recurrence already gives Q10Q_{10} wrong by four orders of magnitude and Q20Q_{20} wrong by twenty-three. Instead either use the closed form of part 1 with LL replaced by 12ln⁡x+1x−1\tfrac12\ln\frac{x+1}{x-1}, or run the recurrence downward from a high starting index with arbitrary seed values and rescale at the end using the known Q0Q_{0}: downward iteration is stable exactly when the wanted solution is the decaying one.
  3. To obtain the polynomial part RnR_{n} in closed form: use Rn=∑m=1n1mPm−1Pn−mR_{n}=\sum_{m=1}^{n}\frac1mP_{m-1}P_{n-m} directly, or iterate Rn+1=[(2n+1)xRn−nRn−1]/(n+1)R_{n+1}=\bigl[(2n+1)xR_{n}-nR_{n-1}\bigr]/(n+1) from R0=0R_{0}=0, R1=1R_{1}=1 — the same recurrence, since PnLP_{n}L satisfies it on its own. Check the result with Rn(1)=HnR_{n}(1)=H_{n}, which catches almost every arithmetic slip.
  4. To solve a boundary-value problem for (L) on [a,b]⊂(−1,1)[a,b]\subset(-1,1): write the general solution c1Pn+c2Qnc_{1}P_{n}+c_{2}Q_{n} and impose the two conditions. Because part 5 gives pW=1pW=1 exactly, no normalising determinant has to be computed: the solution of the two-by-two system can be written down directly, and and the Green's function of Green's function for a boundary-value problem built from u=aPn+bQnu=aP_{n}+bQ_{n} and v=cPn+dQnv=cP_{n}+dQ_{n} carries the denominator p W[u,v]=(ad−bc) p W[Pn,Qn]=ad−bcp\,W[u,v]=(ad-bc)\,p\,W[P_{n},Q_{n}]=ad-bc, a number fixed once by the two boundary conditions rather than a determinant to be evaluated at every point.
  5. For the associated equation with m≠0m\neq0, use Qnm=(1−x2)m/2dmQn/dxmQ_{n}^{m}=(1-x^{2})^{m/2}\mathrm{d}^{m}Q_{n}/\mathrm{d}x^{m}, the second-kind counterpart of the construction in Associated Legendre functions; these are again singular at x=±1x=\pm1 and again discarded on the full sphere.
Example(Building Q3Q_{3}, and four checks)Original

Step 1: the polynomial part. From R0=0R_{0}=0, R1=1R_{1}=1 and Rn+1=[(2n+1)xRn−nRn−1]/(n+1)R_{n+1}=\bigl[(2n+1)xR_{n}-nR_{n-1}\bigr]/(n+1),

R2=3x⋅1−1⋅02=3x2,R3=5x⋅3x2−2⋅13=5x22−23.R_{2}=\frac{3x\cdot1-1\cdot0}{2}=\frac{3x}{2}, \qquad R_{3}=\frac{5x\cdot\frac{3x}{2}-2\cdot1}{3}=\frac{5x^{2}}{2}-\frac{2}{3}.

The sum formula agrees: R3=P0P2+12P1P1+13P2P0=43⋅3x2−12+12x2=52x2−23R_{3}=P_{0}P_{2}+\tfrac12P_{1}P_{1}+\tfrac13P_{2}P_{0}=\tfrac43\cdot\tfrac{3x^{2}-1}{2}+\tfrac12x^{2}=\tfrac52x^{2}-\tfrac23. Hence, with L=artanh⁡xL=\operatorname{artanh}x,

Q3(x)=12(5x3−3x) L(x)−52x2+23.Q_{3}(x)=\tfrac12(5x^{3}-3x)\,L(x)-\tfrac52x^{2}+\tfrac23 .

Check 1 (the harmonic number). R3(1)=52−23=15−46=116R_{3}(1)=\tfrac52-\tfrac23=\tfrac{15-4}{6}=\tfrac{11}{6}, and H3=1+12+13=116H_{3}=1+\tfrac12+\tfrac13=\tfrac{11}{6}. Agreed.

Step 2: a value. Take x=12x=\tfrac12, where L=12ln⁡3L=\tfrac12\ln3 and

P2(12)=−18,P3(12)=12(58−32)=−716,R3(12)=58−23=−124.P_{2}(\tfrac12)=-\tfrac18, \qquad P_{3}(\tfrac12)=\tfrac12\bigl(\tfrac58-\tfrac32\bigr)=-\tfrac7{16}, \qquad R_{3}(\tfrac12)=\tfrac58-\tfrac23=-\tfrac1{24}.

So

Q3(12)=−716⋅12ln⁡3+124=124−732ln⁡3=0.0416666…−0.2403214…=−0.1986547715.Q_{3}(\tfrac12)=-\tfrac7{16}\cdot\tfrac12\ln3+\tfrac1{24} =\frac{1}{24}-\frac{7}{32}\ln3 =0.0416666\ldots-0.2403214\ldots=-0.1986547715 .

Check 2 (the recurrence, independently). Q1(12)=14ln⁡3−1=−0.7253469278Q_{1}(\tfrac12)=\tfrac14\ln3-1=-0.7253469278 and Q2(12)=−18⋅12ln⁡3−34=−0.8186632680Q_{2}(\tfrac12)=-\tfrac18\cdot\tfrac12\ln3-\tfrac34=-0.8186632680. Then

Q3(12)=5⋅12⋅(−0.8186632680)−2⋅(−0.7253469278)3=−2.0466581701+1.45069385573=−0.1986547715,Q_{3}(\tfrac12)=\frac{5\cdot\tfrac12\cdot(-0.8186632680)-2\cdot(-0.7253469278)}{3} =\frac{-2.0466581701+1.4506938557}{3}=-0.1986547715 ,

matching the closed form to ten digits.

Check 3 (the Wronskian). P3′=12(15x2−3)P_{3}'=\tfrac12(15x^{2}-3), so P3′(12)=38P_{3}'(\tfrac12)=\tfrac38; and Q3′=P3′L+P3L′−R3′Q_{3}'=P_{3}'L+P_{3}L'-R_{3}' with L′(12)=43L'(\tfrac12)=\tfrac43 and R3′(12)=52R_{3}'(\tfrac12)=\tfrac52, so

Q3′(12)=38⋅12ln⁡3−716⋅43−52=316ln⁡3−3712.Q_{3}'(\tfrac12)=\tfrac38\cdot\tfrac12\ln3-\tfrac7{16}\cdot\tfrac43-\tfrac52=\tfrac3{16}\ln3-\tfrac{37}{12}.

Therefore

P3Q3′−P3′Q3=−716(316ln⁡3−3712)−38(−732ln⁡3+124)=(−21256+21256)ln⁡3+259192−3192=256192=43.P_{3}Q_{3}'-P_{3}'Q_{3} =-\tfrac7{16}\Bigl(\tfrac3{16}\ln3-\tfrac{37}{12}\Bigr)-\tfrac38\Bigl(-\tfrac7{32}\ln3+\tfrac1{24}\Bigr) =\Bigl(-\tfrac{21}{256}+\tfrac{21}{256}\Bigr)\ln3+\tfrac{259}{192}-\tfrac{3}{192} =\tfrac{256}{192}=\tfrac43 .

The logarithms cancel exactly, and 43=1/(1−14)\tfrac43=1/(1-\tfrac14) as part 5 requires.

Check 4 (the singularity). Part 2 predicts Q3(x)≈12ln⁡21−x−116Q_{3}(x)\approx\tfrac12\ln\frac{2}{1-x}-\tfrac{11}{6} near x=1x=1. At x=1−10−5x=1-10^{-5} that is 12ln⁡(2×105)−1.8333333=6.1030363−1.8333333=4.2697030\tfrac12\ln(2\times10^{5})-1.8333333=6.1030363-1.8333333=4.2697030, while the closed form gives 4.26938434.2693843. The discrepancy is 3.2×10−43.2\times10^{-4}, of the predicted size O((1−x)ln⁡(1−x))=1.2×10−4O\bigl((1-x)\ln(1-x)\bigr)=1.2\times10^{-4} up to a constant of order one. Every number above was verified symbolically and in extended precision.

ApplicationAdaptedAdapted from Riley, Hobson & Bence (2006), §18.1.1, eq. (18.7) and the discussion of regularity, pp. 579–581; Arfken, Weber & Harris (2013), §15.6, the convention for x > 1 and Properties 2 and 4, pp. 767–769

The solution one throws away — and why that is a theorem. In any problem posed on a full sphere the polar angle runs over 0≤θ≤π0\le\theta\le\pi, so x=cos⁡θx=\cos\theta reaches both ±1\pm1, and the general separated solution c1Pn+c2Qnc_{1}P_{n}+c_{2}Q_{n} must have c2=0c_{2}=0 because QnQ_{n} is infinite there. Riley, Hobson and Bence take this step at the start of their treatment of PnP_{n} (Separation of variables in standard coordinates). Part 5 is what makes it rigorous rather than cosmetic: it is not that QnQ_{n} is "badly behaved", it is that the bilinear boundary form p (uv′−u′v)p\,(uv'-u'v) for the pair Pn,QnP_{n},Q_{n} equals exactly 11 at each pole instead of vanishing there, so the Legendre operator is self-adjoint on the regular solutions and on nothing larger.

Regions that do not contain the poles. Exclude one or both poles — a conical hole, the gap between two coaxial cones, a sector of a spherical shell — and the boundary condition that killed QnQ_{n} is gone, so both solutions survive and the degree is no longer forced to be an integer. Physically the QQ-part is the piece of the field that remembers the axis has been cut out.

Prolate spheroidal problems and the charged rod. Here is the cleanest place where QnQ_{n} is the answer rather than the discard. Prolate spheroidal coordinates based on foci at z=±cz=\pm c use ξ=(r1+r2)/2c≥1\xi=(r_{1}+r_{2})/2c\ge1 and η=(r1−r2)/2c∈[−1,1]\eta=(r_{1}-r_{2})/2c\in[-1,1], where r1,r2r_{1},r_{2} are the distances to the two foci. Laplace's equation separates with Pn(η)P_{n}(\eta) in the angular variable and, in ξ\xi, the same Legendre equation again — now on (1,∞)(1,\infty), where PnP_{n} grows and QnQ_{n} decays. An exterior field therefore expands as ∑nanPn(η)Qn(ξ)\sum_{n}a_{n}P_{n}(\eta)Q_{n}(\xi), Arfken, Weber and Harris's property Qn(∞)=0Q_{n}(\infty)=0 being exactly the radiation-free condition at infinity. The simplest case is a uniformly charged rod of half-length cc and total charge qq along the axis. Integrating 1/∣r−r′∣1/\lvert\mathbf{r}-\mathbf{r}'\rvert over the rod gives the elementary closed form ln⁡r1+r2+2cr1+r2−2c\ln\frac{r_{1}+r_{2}+2c}{r_{1}+r_{2}-2c}, and in these coordinates r1+r2=2cξr_{1}+r_{2}=2c\xi, so

V=q4πε0 c⋅12ln⁡ξ+1ξ−1=q4πε0 c Q0(ξ).V=\frac{q}{4\pi\varepsilon_{0}\,c}\cdot\frac12\ln\frac{\xi+1}{\xi-1} =\frac{q}{4\pi\varepsilon_{0}\,c}\,Q_{0}(\xi) .

The potential of a charged needle is the Legendre function of the second kind of degree zero. Two checks confirm it: part 7 gives Q0(ξ)∼1/ξQ_{0}(\xi)\sim1/\xi, and ξ→r/c\xi\to r/c far away, so V→q/(4πε0r)V\to q/(4\pi\varepsilon_{0}r) as it must; and the equipotentials ξ=const\xi=\text{const} are confocal prolate spheroids, which is why the same formula is the potential of any charged conducting spheroid with those foci. The higher Qn(ξ)Q_{n}(\xi) do the same job for the multipole terms of a spheroidal conductor, the spheroidal analogue of the r−n−1r^{-n-1} of Multipole expansions.

Green's functions. Part 5 says p W[Pn,Qn]=1p\,W[P_{n},Q_{n}]=1 identically, which is unusually convenient: the one-dimensional Green's function of Green's function for a boundary-value problem for the Legendre operator on a sub-interval of (−1,1)(-1,1) is

G(x,s)=u(x<) v(x>)p W[u,v]G(x,s)=\frac{u(x_{<})\,v(x_{>})}{p\,W[u,v]}

with u,vu,v the combinations of Pn,QnP_{n},Q_{n} satisfying the left and right conditions; the prefactor 1/(pW)1/(pW) is a constant, by Step 1 of the Wronskian proof, so it can be computed once from the two combinations rather than at every evaluation. The same structure, with PnP_{n} regular at the inner end and QnQ_{n} decaying at the outer, is what makes Pn(x<)Qn(x>)P_{n}(x_{<})Q_{n}(x_{>}) the standard kernel in exterior and annular problems.