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.
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 be an interval, finite or infinite. A weight on (a,b) is a measurable function w with
w(x)>0 for every x∈(a,b), and
finite momentsμk=∫abxkw(x)dx for every integer k≥0.
Hassani's Lw2(a,b) is the space of measurable f with ∫ab∣f∣2w<∞, carrying
(f,g)w=∫abf(x)g(x)w(x)dx,∥f∥w=(f,f)w1/2,
and it is complete. His standing requirement on w is that it be strictly positive and integrable on (a,b); the moment condition above is the extra hypothesis that makes every monomial a member of Lw2, since ∥xk∥w2=μ2k. Both are needed here and both are stated, rather than left to be read off from the examples.
A sequence (pn)n≥0 of real polynomials is an orthogonal family for w if degpn=n for every n and (pn,pm)w=0 whenever n=m. We write
hn=∥pn∥w2
for the squared norms, and kn(n), kn(n−1) for the coefficients of xn and of xn−1 in pn, following Hassani's notation. The family is monic if kn(n)=1 for all n. A standardisation is any rule fixing the one free scalar in each pn: Legendre polynomials are standardised by pn(1)=1, Chebyshev polynomials of the first kind by a leading coefficient 2n−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 w be a weight on (a,b) in the sense above. Then:
Existence. Applying the Gram-Schmidt process of Orthonormal bases and Gram-Schmidt to 1,x,x2,… in Lw2(a,b) produces an orthogonal family (pn) for w, with degpn=n exactly.
Uniqueness up to scale. If (pn) and (p~n) are both orthogonal families for w, then p~n=cnpn with cn=0 for every n. In particular the monic family is unique, and any standardisation determines the family completely.
Orthogonality to all lower degrees. For every n,
(pn,q)w=0for every polynomial q with degq<n.
Equivalently, pn is orthogonal to 1,x,…,xn−1.
Hassani records part 3 as the "obvious property" that any polynomial of degree below n is a combination of p0,…,pn−1.
Theorem(Three-term recurrence)AdaptedAdapted from Hassani (2013), §7.2.1, eqs. (7.9)–(7.13), pp. 223–224
Let (pn) be an orthogonal family for w. Then for every n≥1,
valid for n≥1, together with p0=1 and p1(x)=x−B0, B0=μ1/μ0.
Theorem(The zeros are real, simple and interior)Original
Let (pn) be an orthogonal family for a weight w on (a,b), and let n≥1. Then pn has exactly n zeros, they are all real and simple, and they all lie in the open interval (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 s is a polynomial of degree at most 2 with only real roots, w is strictly positive and integrable on (a,b) with
w(a)s(a)=0=w(b)s(b),
and the function F1 defined below is a polynomial of degree 1. Set
Fn(x)=Knw(x)1dxndn[w(x)s(x)n]
with normalising constants Kn: this is the generalised Rodrigues formula. Then each Fn is a polynomial of degree exactly n and ∫abFnpw=0 for every polynomial p of degree <n, so (Fn) is an orthogonal family for w. Moreover each Fn solves the Sturm-Liouville equation
dxd(wsdxdFn)=wλnFn,λn=K1k1(1)n+σ2n(n−1),
where σ2 is the coefficient of x2 in s.
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,
multiply by the factor w=p−1exp∫(q/p) that makes it self-adjoint (so that (wp)′=wq), and then
yn(x)=w(x)1(dxd)n[w(x)p(x)n]
is a polynomial of degree n solving the equation with
λn=−n(n−1)α−nμ.
The two routes agree with s=p: Arfken's self-adjointness condition (wp)′=wq is Hassani's requirement that (1/w)d(ws)/dx be a polynomial of degree 1.
Classification. Hassani's Proposition 8.2.1 shows that the degree of s decides everything. Up to an affine change of variable there are only three cases:
degs
s(x)
w(x)
(a,b)
family
0
1
e−x2
(−∞,∞)
Hermite Hn
1
x
xνe−x, ν>−1
(0,∞)
Laguerre Lnν
2
1−x2
(1+x)μ(1−x)ν, μ,ν>−1
(−1,1)
Jacobi Pnμ,ν
Legendre (μ=ν=0), Gegenbauer (μ=ν=λ−21) and the two kinds of Chebyshev (μ=ν=∓21) 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)=x on (0,1) in the example below, which is Jacobi only after the affine map to (−1,1), or w(x)=e−x4 on 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 — pn is orthogonal to everything of lower degree — used three times.
Existence and uniqueness. Gram-Schmidt cannot fail, because w>0 makes ∥q∥w>0 for every nonzero polynomial q, so no vector in the process is ever killed; and taking n+1 monomials in and getting n+1 polynomials out forces the degrees to be 0,1,…,n. Uniqueness holds because pn is determined up to scale as the one direction in the (n+1)-dimensional space of polynomials of degree ≤n perpendicular to the n-dimensional subspace of degree <n.
Recurrence. Subtract from pn+1 the multiple of xpn that cancels its leading term. What is left has degree ≤n, so it expands in p0,…,pn. Test that expansion against pm: the pn+1 term dies for m≤n by orthogonality, and the xpn term dies as soon as xpm has degree below n+1, that is for m≤n−2. So all coefficients below pn−1 vanish and only three terms survive.
Zeros. If pn changed sign at fewer than n interior points, we could manufacture a low-degree polynomial q whose sign changes are exactly those of pn. Then pnq never changes sign, so ∫pnqw cannot vanish — yet it must, because degq<n.
Rodrigues. Different in kind, and the only place where w has to be differentiable. The whole statement rests on one bookkeeping lemma: differentiating wsn once lowers the power of s by one and raises the degree of the polynomial factor by one, because the hypothesis on F1 says that (ws)′ is w times a polynomial of degree ≤1. Iterating gives dm(wsn)/dxm=wsn−m×(polynomial), which is a polynomial times w when m=n — so Fn is a polynomial — and which vanishes at both ends when m<n, because a factor ws survives. Those vanishing boundary values are exactly what is needed to integrate by parts k+1 times against a polynomial of degree k<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 Lw2 and are independent.∥xk∥w2=μ2k<∞ by the moment hypothesis, so each monomial is in the space. A nontrivial linear combination of monomials is a nonzero polynomial q, and
∥q∥w2=∫ab∣q∣2w>0
because ∣q∣2w is continuous, non-negative, strictly positive off the finitely many zeros of q — this is where w>0 enters. So q=0 in Lw2 and the monomials are linearly independent.
Gram-Schmidt. Run the process of Orthonormal bases and Gram-Schmidt on 1,x,x2,…. At step n it subtracts from xn 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,…,xn, hence a polynomial of degree ≤n; and the coefficient of xn is the (nonzero) reciprocal of that norm times 1, because the subtracted projection involves only 1,…,xn−1. So degpn=n exactly. This is the construction Kreyszig carries out on [−1,1] with w≡1 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, the polynomials of degree ≤n−1. Since pn⊥pm for m<n, pn⊥Pn−1. This is part 3.
Uniqueness. Let p~n be any polynomial of degree n orthogonal to Pn−1. Write p~n=∑j=0ncjpj, possible because degp~n≤n. For m<n, orthogonality to pm∈Pn−1 gives 0=(p~n,pm)w=cmhm, and hm>0, so cm=0. Hence p~n=cnpn, with cn=0 since degp~n=n. Any second orthogonal family satisfies the same hypothesis at each n, which proves part 2. □
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)−αnxpn(x),αn=kn(n)kn+1(n+1),
has no xn+1: the coefficient of xn+1 in xpn is kn(n), and αn was chosen to match. So degr≤n and r=∑j=0najpj for some scalars aj.
Step 2: test against pm for m≤n−2. Take (⋅,pm)w of r=pn+1−αnxpn:
The first term vanishes by orthogonality since m≤n−2<n+1. In the second we moved the real factor x across the inner product, which is legitimate because x is real and the inner product has xpm=xpm. Now deg(xpm)=m+1≤n−1<n, so (pn,xpm)w=0 by part 3 of the previous statement. Hence (r,pm)w=0, while also (r,pm)w=amhm from the expansion. Since hm>0,
am=0for m=0,1,…,n−2,
which is exactly where the hypothesis "orthogonal to all lower degrees" does the work. Only an and an−1 survive, giving (∗) with βn=an and γn=an−1.
Step 3: identify βn. Compare coefficients of xn in pn+1=αnxpn+βnpn+γnpn−1. The last term contributes nothing (degpn−1=n−1), so
To evaluate (pn,xpn−1)w, apply (∗) at index n−1: xpn−1=αn−1−1(pn−βn−1pn−1−γn−1pn−2), and only the pn term survives against pn:
(pn,xpn−1)w=αn−1hn.
Therefore γn=−αnhn/(αn−1hn−1), as claimed. Since hn,hn−1>0, the sign of γn is opposite to that of αn/αn−1.
Step 5: the monic form. Put αn=1 for all n, write Bn=−βn and Cn=−γn=hn/hn−1>0; then (∗) becomes (∗∗). For Bn the coefficient formula is awkward to use directly, so take (⋅,pn)w of (∗∗) instead:
giving Bn=(xpn,pn)w/hn. For n=0 the same computation with p0=1 gives B0=(x,1)w/(1,1)w=μ1/μ0. □
Proof(The zeros)Original
pn is a real polynomial of degree n, so it has n zeros in C counted with multiplicity. We show n of them are distinct points of (a,b); that exhausts the count and leaves none complex and none repeated.
Step 1: count the interior sign changes. Let x1<⋯<xr be the points of (a,b) at which pn changes sign, that is the real zeros of pn in (a,b) of odd multiplicity. Suppose, for contradiction, that r<n. (If pn has no interior sign change, take r=0 and q≡1 below.)
Step 2: build a test polynomial. Put
q(x)=j=1∏r(x−xj),
a real polynomial of degree r<n.
Step 3: the product has one sign. The function pnq is a real polynomial that changes sign at no point of (a,b): at each xj both factors change sign, so the product does not, and at every other point of (a,b) neither factor does. Hence pnq≥0 throughout (a,b) or pnq≤0 throughout (a,b). It is not the zero polynomial, since pn=0 and q=0.
Step 4: the integral cannot vanish. Replacing q by −q if necessary, assume pnq≥0 on (a,b). The product vanishes only at its finitely many zeros, and w>0 everywhere on (a,b), so pnqw is non-negative and strictly positive except at finitely many points. Therefore
∫abpn(x)q(x)w(x)dx>0.
(The integral is finite: pnq is a polynomial, and all moments of w are finite.) This is the step that needs w>0 on all of (a,b).
Step 5: but orthogonality forbids it.q is real and degq=r<n, so by part 3 of the existence statement ∫abpnqw=(pn,q)w=0. This contradicts Step 4.
Hence r≥n. A degree-n polynomial has at most n zeros, so r=n: there are exactly n distinct points of (a,b) where pn changes sign. Each is a zero, and together they account for all n zeros of pn; so each has multiplicity one and there are no others, real or complex. □
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≤k for an unspecified polynomial of degree at most k, as Hassani does, and let Fn=(Knw)−1dn(wsn)/dxn. The hypotheses are: degs≤2 with real roots only; w>0 and integrable on (a,b) with w(a)s(a)=0=w(b)s(b); and F1 a polynomial of degree 1. The last one says precisely that
dxd(ws)=wt,t=K1F1a polynomial withdegt≤1.(H)
Lemma A (Hassani's 8.1.2). For 1≤m≤n,
dxmdm[wsnp≤k]=wsn−mp≤k+m.
Proof, by induction on m. For m=1, expand and use w′=w(t−s′)/s, which is (H) rearranged (valid where s=0, and the conclusion is an identity between continuous functions, so it extends to the whole interval):
In the bracket, t and s′ have degree ≤1, so the first product has degree ≤k+1; and sp≤k′ has degree ≤2+(k−1)=k+1. The bracket is therefore p≤k+1, which is the claim for m=1. If the claim holds for m<n, apply the case m=1 once more with n−m in place of n and k+m in place of k; the power of s is still at least 1, so the step is legitimate, and we get wsn−m−1p≤k+m+1. □
Lemma B (Hassani's 8.1.3). For 0≤m<n, dm(wsn)/dxm vanishes at x=a and at x=b.
Proof. Take k=0, p≤0=1 in Lemma A: the derivative equals wsn−mp≤m. Since m<n we have n−m≥1, so this is (ws)sn−m−1p≤m, and ws vanishes at both endpoints by hypothesis. □
Fn is a polynomial of degree at most n. Lemma A with m=n, k=0 gives dn(wsn)/dxn=ws0p≤n=wp≤n, so Fn=p≤n/Kn is a polynomial of degree at most n.
Orthogonality to lower degrees. Let q be a polynomial with degq=k<n. Then
Kn∫abqFnwdx=∫abq(x)dxndn(wsn)dx.
Integrate by parts, moving one derivative at a time onto q. After j steps the boundary term is
(−1)j−1[q(j−1)dxn−jdn−j(wsn)]ab,
which vanishes by Lemma B because n−j<n for j≥1. Carrying this out k times leaves
where C=q(k) is the constant kth derivative of a degree-k polynomial, and where n−k−1≥0 because k<n. The last integral is [dn−k−1(wsn)/dxn−k−1]ab=0, again by Lemma B. So ∫abqFnw=0.
The degree is exactly n. Write Fn=p≤n−1+kn(n)xn, multiply by wFn and integrate:
∫abFn2w==0by the previous step∫abp≤n−1Fnw+kn(n)∫abxnFnw.
The left side is strictly positive, because w>0 and Fn is a nonzero polynomial (it is nonzero: if Fn≡0 then dn(wsn)/dxn≡0, making wsn a polynomial of degree <n, which contradicts Lemma A at m=n−1 unless w vanishes somewhere). Hence the right side is nonzero, so kn(n)=0 and degFn=n.
Combining the last two steps: (Fn) has degFn=n and is orthogonal to all lower degrees, so by the uniqueness statement proved above it is the orthogonal family for w, standardised by the choice of Kn. 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. □
Remark(Where the hypotheses are used)Original
w>0 on all of (a,b). Used twice: to make ∥q∥w>0 for every nonzero polynomial, so that Gram-Schmidt never divides by zero and hn>0 (needed in every step of the recurrence proof); and in Step 4 of the zeros proof. Weakening it to w≥0 breaks the location of the zeros. Take (a,b)=(−1,1) with w=1 on (−1,0) and w=0 on [0,1). The resulting "orthogonal polynomials" are the Legendre polynomials transplanted to (−1,0); all of their zeros lie in (−1,0), and the half-interval (0,1) contains none, so "every zero lies in (a,b)" is no longer the sharp statement it was. If instead w is replaced by a measure with only N atoms, then PN has dimension N+1 but the space has dimension N: the Gram-Schmidt step at index N divides by zero and the family stops at pN−1.
Finite moments. Used to put the monomials in Lw2 at all. For w(x)=(1+x2)−1 on R one has μ0=π<∞, so w is integrable and strictly positive, but μ2=∫x2w=∞: the function x is not in Lw2, and the construction produces only p0. 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 pm). 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)-term: the degree of xpm must stay below n.
w 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 x across the inner product. With a complex-valued weight neither survives, and the zeros need not be real.
Nothing about smoothness of w. No derivative of w is used in any of the three theorems. Smoothness enters only in the Rodrigues construction, where hypothesis (H) makes (ws)′ exist and equal w 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 w on (a,b) and a target degree N:
Compute the momentsμ0,…,μ2N. Stop here if any is infinite: the weight is not admissible.
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)=1) and rescale at the end.
Start the recurrence.p0=1, h0=μ0, B0=μ1/μ0, p1(x)=x−B0, h1=μ2−μ12/μ0.
Iterate (∗∗). For n=1,…,N−1: form (xpn,pn)w and hn as integrals against w (or as combinations of the moments), set Bn=(xpn,pn)w/hn and Cn=hn/hn−1, and put pn+1=(x−Bn)pn−Cnpn−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.
Check as you go.Cn>0 always; and pn must have n simple zeros inside (a,b). Either failing is a sign of arithmetic error.
For the nodes and weights of the n-point Gauss rule, take the zeros x1,…,xn of pn from step 5 and solve the n×n linear system ∑iWixik=μk, k=0,…,n−1, for the quadrature weights Wi.
Example(The family for w(x)=x on (0,1), and its two-point Gauss rule)Original
Take (a,b)=(0,1) and w(x)=x: strictly positive on the open interval, with moments μk=∫01xk+1dx=1/(k+2), all finite. We build the monic family.
since 4516−181=9032−905=9027=103: the recurrence reproduces p2 exactly, as (∗∗) requires. One more turn, with h2=∫01p22xdx=6001, gives B2=3518, C2=h2/h1=503 and
p3(x)=x3−712x2+76x−354.
Step 4: the zeros. From 10x2−12x+3=0,
x1,2=106∓6,x1=0.355051…,x2=0.844949…,
two simple zeros, both inside (0,1), as the zeros theorem demands. (p3 likewise has three: 0.212341, 0.590533, 0.911412.)
Step 5: the two-point Gauss rule. Take the two zeros as nodes and solve W1+W2=μ0=21, W1x1+W2x2=μ1=31. Subtracting x2 times the first from the second,
Step 6: the checkable number. The rule ∫01f(x)xdx≈W1f(x1)+W2f(x2) should be exact for f of degree ≤3 and fail at degree 4. For f(x)=x3 the true value is ∫01x4dx=51, and
exactly — in decimals 0.1819586×0.0447582+0.3180414×0.6032418=0.0081441+0.1918559=0.2000000. At degree 4 the rule gives 20033=0.165 against the true ∫01x5dx=61=0.16, so degree 3=2⋅2−1 is indeed the last exact degree. Every quantity in this example was checked symbolically.
Gaussian quadrature. The zeros theorem is the reason Gauss rules exist. Riley, Hobson and Bence describe the scheme: to integrate over (a,b) against a weight w, take the n nodes to be the zeros of the nth orthogonal polynomial for w — which the theorem guarantees are n 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 n-point rule is exact for every polynomial of degree up to 2n−1, twice what n equally spaced points buy. Their Gauss-Legendre case (w≡1 on [−1,1], nodes the zeros of Pn, weights 2/[(1−xi2)∣Pn′(xi)∣2]) is the workhorse of finite-element and spectral codes; their three-point rule applied to ∫01(1+x2)−1dx returns 0.78527 against the exact π/4=0.78540, where the trapezium and Simpson rules on the same three points give 0.7750 and 0.7833. The weight is chosen to absorb whatever is awkward about the integrand: e−x on (0,∞) for Gauss-Laguerre, e−x2 on R for Gauss-Hermite, (1−x2)±1/2 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 (∗) upward from p0 and p1.
Least squares and spectral methods. Because pn⊥Pn−1, the best degree-N approximation to f in ∥⋅∥w is the truncation ∑n≤N(f,pn)wpn/hn of a single expansion: raising N 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.
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).
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
so p=1−x2, q=0, weight ρ=1, eigenvalue λ=ℓ(ℓ+1), on the natural interval [−1,1]. The equation has regular singular points at x=−1,1,∞ and an ordinary point at x=0. Riley, Hobson and Bence call any solution of (L) a Legendre function; in applications x=cosθ is the cosine of a polar angle, so −1≤x≤1.
For ℓ=n a non-negative integer, (L) has a polynomial solution of degree n, unique up to scale. The Legendre polynomialPn is that solution standardised by
Pn(1)=1.
This standardisation forces Pn(−1)=(−1)n. Explicitly,
The second solution of (L) for integer ℓ is an infinite series, singular at x=±1; 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) is the orthogonal family on (−1,1) for the weight w≡1, standardised at x=1 instead of monically. Kreyszig reaches it from the other end: he applies Gram-Schmidt to 1,x,x2,… in L2[−1,1] and identifies the output, up to the factor (2n+1)/2 that makes it a unit vector, as Pn. Hassani reaches it as the Jacobi case μ=ν=0 of the general Rodrigues construction, with standardisation constant Kn=(−1)n2nn!.
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,k denote non-negative integers. The following hold.
Rodrigues' formula.Pl(x)=2ll!1dxldl(x2−1)l.
(Hassani writes the same thing as (−1)l(2ll!)−1dl(1−x2)l/dxl.)
Generating function. For ∣x∣≤1 and ∣h∣<1,
G(x,h)=(1−2xh+h2)−1/2=l=0∑∞Pl(x)hl.
The closed endpoints x=±1 are included on the authority of Riley, Hobson and Bence, who evaluate G(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 (and then also sets z=1 in it).
Three-term recurrence.(l+1)Pl+1(x)=(2l+1)xPl(x)−lPl−1(x),P0=1,P1=x.
Among the derivative recurrences, the ones most used are
(1−x2)Pl′=l(Pl−1−xPl),(2l+1)Pl=Pl+1′−Pl−1′.
Orthogonality and normalisation.∫−11Pl(x)Pk(x)dx=2l+12δlk.
Consequently ((2l+1)/2Pl)l≥0 is orthonormal in L2[−1,1], and (Kreyszig) it is total there: every f∈L2[−1,1] has
f=l=0∑∞alPl,al=22l+1∫−11f(x)Pl(x)dx,
the series converging in L2. Riley, Hobson and Bence state the expansion for any f satisfying the Dirichlet conditions of a Fourier series.
Parity and zeros.Pl(−x)=(−1)lPl(x). Inside (−1,1), Pl has exactly l zeros, all simple; and (Strauss) for 1≤k≤l the kth derivative dkPl/dxk has exactly l−k zeros in (−1,1), none of these derivatives vanishing at x=±1.
Values at the ends and the centre.Pl(1)=1, Pl(−1)=(−1)l, and (Hassani)
P2m(0)=22m(m!)2(−1)m(2m)!,P2m+1(0)=0.
Idea of the proofOriginal
The four descriptions are four ways to pin down "the degree-l polynomial solution of (L) with value 1 at x=1", and each proof is the same two moves: show the candidate solves (L), then check it at x=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)l satisfies the first-order relation (x2−1)u′=2lxu, and differentiating that l+1 times turns it into (L) for u(l). For the generating function, differentiating G in x and in h produces two identities between ∑Plhl and ∑Pl′hl; 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 integrated by parts, where the boundary terms die because p=1−x2 vanishes at both ends. The value2/(2l+1) is a separate computation, and there are two cheap routes: integrate Rodrigues' formula by parts l times (Riley-Hobson-Bence, Kreyszig), or play the recurrence against itself (Strauss) and never integrate anything.
Step 1: a first-order relation. Differentiating, u′=2lx(x2−1)l−1, hence
(x2−1)u′−2lxu=0.
Step 2: differentiate it l+1 times. Apply Leibniz's rule to each product. In (x2−1)u′ only three terms survive, because x2−1 has vanishing third derivative; in xu only two:
and multiplying by −1 this is precisely (L) with y=u(l) and ℓ=l.
Step 3: u(l) is a polynomial of degree l, regular at ±1.u is a polynomial of degree 2l, so u(l) is a polynomial of degree l; being a polynomial it is regular everywhere, in particular at the singular points x=±1 of (L). The space of such solutions is one-dimensional, so u(l)=clPl for some constant cl.
Step 4: evaluate at x=1 to get cl. Write u=(x−1)l(x+1)l and apply Leibniz:
u(l)=r=0∑l(rl)dxrdr(x−1)ldxl−rdl−r(x+1)l.
At x=1 every term with r<l carries a surviving positive power of x−1 and so vanishes; the term r=l gives l!⋅(x+1)l, that is l!2l. Hence u(l)(1)=2ll!, and since Pl(1)=1 we get cl=2ll!, which is Rodrigues' formula. □
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)−1/2. For ∣x∣≤1 the quadratic 1−2xh+h2 has no zero with ∣h∣<1 (its roots are x±i1−x2, of modulus 1), so G is analytic in h on the unit disc and has a Taylor expansion G=∑l≥0Ql(x)hl there, with coefficients Ql to be identified.
Step 1: two differentiated forms. Differentiating G gives
Step 2: a relation with no fractional powers. From the first display, (1−2xh+h2)∑lQl′hl=h∑lQlhl. Matching the coefficient of hl+1,
Ql=Ql+1′−2xQl′+Ql−1′.(a)
Step 3: a second relation. Dividing the two displays of Step 1 eliminates the −3/2 power: (x−h)∑lQl′hl=h∑llQlhl−1. The coefficient of hl gives
xQl′−Ql−1′=lQl.(b)
Step 4: eliminate. Read (b) as Ql−1′=xQl′−lQl and substitute it into (a); the term Ql−1′ disappears and the two multiples of xQl′ combine to leave
(l+1)Ql=Ql+1′−xQl′.(c)
Now take (c) with l replaced by l−1, that is lQl−1=Ql′−xQl−1′, and subtract from it x times (b), that is lxQl=x2Ql′−xQl−1′. The terms xQl−1′ cancel and
l(Ql−1−xQl)=Ql′−x2Ql′=(1−x2)Ql′.(d)
Step 5: (L). Differentiate (d) and use (b) once more:
where the middle equality is (b) solved for Ql−1′−xQl′=−lQl. So each Ql satisfies (L) with ℓ=l.
Step 6: identify Ql=Pl. Each Ql is a polynomial of degree l (expand G in powers of 2xh−h2 by the binomial series and collect; the coefficient of hl is a polynomial in x of degree l), so it is regular at x=±1 and therefore a multiple of Pl. Setting x=1,
G(1,h)=[(1−h)2]−1/2=1−h1=l≥0∑hl,
so Ql(1)=1 for every l, matching the standardisation. Hence Ql=Pl, which proves part 2, and (b), (c), (d) become the derivative recurrences of part 3.
(Strauss obtains Step 5 in one stroke instead: G satisfies the partial differential equation [(1−x2)Gx]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 l-fold derivative into Cauchy's integral — the Schlaefli integral — summing the resulting geometric series in h and evaluating one residue.)
Step 7: the three-term recurrence. Substitute ∑Plhl=(1−2xh+h2)−1/2 into ∂G/∂h from Step 1, so that
(x−h)l∑Plhl=(1−2xh+h2)l∑lPlhl−1.
The coefficient of hl on the left is xPl−Pl−1; on the right it is (l+1)Pl+1−2xlPl+(l−1)Pl−1. Equating and rearranging,
(l+1)Pl+1=(2l+1)xPl−lPl−1.□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=k. Both Pl and Pk satisfy (L), written in Sturm-Liouville form; multiply the equation for Pl by Pk and integrate over [−1,1]:
∫−11Pk[(1−x2)Pl′]′dx+l(l+1)∫−11PkPldx=0.
Integrate the first term by parts. The boundary term is [(1−x2)PkPl′]−11=0, because the factor 1−x2 vanishes at both ends while Pk and Pl′ are finite there — this is precisely where the regularity of Pl at x=±1 is used. So
−∫−11(1−x2)Pk′Pl′dx+l(l+1)∫−11PkPl=0.
Now swap the roles of l and k and subtract: the first integral is symmetric and cancels, leaving
[l(l+1)−k(k+1)]∫−11PkPldx=0.
Since l=k are non-negative integers, l(l+1)=k(k+1), so the integral vanishes. (Taking k=0 gives the frequently used ∫−11Pl=0 for l≥1.) Alternatively this is immediate from part 3 of General properties of orthogonal polynomials once Pl is known to be the orthogonal family for w≡1; Kreyszig proves it directly by showing ∫−11xmPldx=0 for m<l through m integrations by parts of Rodrigues' formula.
The norm, by the recurrence (Strauss's route). Write (f,g)=∫−11fg. Take the inner product of the three-term recurrence (l+1)Pl+1=(2l+1)xPl−lPl−1 with Pl−1. Orthogonality kills (Pl+1,Pl−1), leaving
0=(2l+1)(xPl,Pl−1)−l(Pl−1,Pl−1).
Next write the recurrence with l replaced by l−1, namely lPl=(2l−1)xPl−1−(l−1)Pl−2, and take its inner product with Pl; now (Pl−2,Pl)=0 and
l(Pl,Pl)=(2l−1)(xPl−1,Pl).
The two mixed terms are the same number, since (xPl,Pl−1)=(xPl−1,Pl). Eliminating it,
(2l+1)(Pl,Pl)=(2l−1)(Pl−1,Pl−1),
valid for l≥1. So (2l+1)(Pl,Pl) does not depend on l; at l=0 it equals 1⋅∫−111dx=2. Hence (Pl,Pl)=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 l times; every boundary term carries an undifferentiated factor x2−1 and dies, so
using d2l(x2−1)l/dx2l=(2l)!. The remaining integral Kl=∫−11(1−x2)l satisfies (2l+1)Kl=2lKl−1 (integrate by parts against the constant 1 and write 2x2=2−2(1−x2)), whence Kl=22l+1(l!)2/(2l+1)! and the product is 2/(2l+1) again.
Completeness. Kreyszig's argument: polynomials are dense in C[−1,1] in the supremum norm by the Weierstrass approximation theorem, C[−1,1] is dense in L2[−1,1], and span{P0,…,Pm} is the set of polynomials of degree ≤m. So the orthonormal system (2l+1)/2Pl 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 Pk term by term. □
Remark(Where the hypotheses are used)Original
ℓ an integer. Nothing else makes the Frobenius series terminate. Riley, Hobson and Bence solve (L) at the ordinary point x=0 and find the two-step recurrence
an+2=(n+1)(n+2)n(n+1)−ℓ(ℓ+1)an,
whose numerator vanishes at n=ℓ exactly when ℓ is a non-negative integer. Both series otherwise continue for ever, with radius of convergence 1 — the distance to the nearest singular point — and both diverge at x=±1. So for non-integer ℓ there is no solution of (L) bounded on the closed interval [−1,1], which is why in a physical problem on a full sphere the separation constant is forced to be ℓ(ℓ+1) with ℓ integral.
Regularity at x=±1. Used three times: in Step 3 of the Rodrigues proof and Step 6 of the generating-function proof, to identify a candidate with Pl rather than with some mixture of Pl 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,
Q0(x)=21ln1−x1+x=artanhx.
Its singularity at x=±1 is only logarithmic, so Q0 is perfectly square integrable: ∫−11Q02dx=π2/6=1.644934…, a convergent integral. What goes wrong is the one step the proof cannot do without. Since Q0′=(1−x2)−1, the product
(1−x2)Q0′=1
identically, so the factor 1−x2 no longer quenches anything and the boundary term survives. Running the Sturm-Liouville argument on the pair P1=x and Q0, with eigenvalues λ1=2 and λ0=0,
the second piece vanishing at both ends because the linear factor beats the logarithm: (1−x2)artanhx→0 as x→±1. Hence ∫−11P1Q0dx=1, not 0: two solutions of (L) with different eigenvalues, both square integrable, and not orthogonal. Regularity at x=±1, not integrability, is the hypothesis doing the work.
p=1−x2 vanishing at both ends. This is what makes [−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] with c<1 the boundary term at c does not vanish and the Pl are not orthogonal there; the correct family for [a,b] is Kreyszig's affine transplant Pl(1+2b−ax−b).
∣h∣<1 in the generating function. The series is a Taylor series about h=0 and the nearest singularity of G in h sits at modulus 1 for ∣x∣≤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 L2 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 f produces the same overshoot near the jump as in a Fourier series.
Method(Computing with Legendre polynomials)Original
To tabulate or evaluate Pl(x): never differentiate Rodrigues' formula and never sum the explicit coefficient series. Start from P0=1, P1=x and iterate
Pl+1=l+1(2l+1)xPl−lPl−1.
Each step costs one multiplication and one subtraction. Upward iteration is numerically safe on ∣x∣≤1, though not because the coefficients are small or of one sign — the Pl−1 term enters with a minus sign, and (2l+1)x/(l+1) is negative throughout x<0. The reason is that on [−1,1] the two solutions of this recurrence, Pl and the second-kind solution Ql, are of the same order of magnitude: each decays only like l−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 l stays a rounding error. (Outside [−1,1], where Pl grows and Ql decays, the same recurrence is stable upward for Pl and useless upward for Ql.) Note also ∣Pl(x)∣≤1 on [−1,1], so nothing overflows.
To expand a given f on [−1,1]: compute al=22l+1∫−11fPl. Use the parity rule first: if f is even, every odd al vanishes, and conversely, which halves the work.
To expand a polynomial exactly: match from the top down. The coefficient of xN in PN is (2N)!/[2N(N!)2]; subtract the right multiple of PN, and repeat on the remainder, which now has degree ≤N−1. This is exact arithmetic and avoids every integral.
To differentiate or integrate an expansion: use (2l+1)Pl=Pl+1′−Pl−1′, which integrates to ∫Pl=(Pl+1−Pl−1)/(2l+1) up to a constant, and (1−x2)Pl′=l(Pl−1−xPl) for the derivative.
To solve a boundary-value problem on a sphere with azimuthal symmetry: expand the boundary data in Pl(cosθ) and match coefficient by coefficient against the general separated solution ∑l(Alrl+Blr−l−1)Pl(cosθ); see Separation of variables in standard coordinates.
Example(Expanding x4, and five values of Pl at x=0.3)Original
Part 1: the expansion of x4. The function x4 is even and of degree 4, so only a0,a2,a4 can be nonzero. Work from the top, as in step 3 of the method. The leading coefficient of P4=81(35x4−30x2+3) is 35/8, so
x4−358P4=x4−(x4−3530x2+353)=76x2−353.
The leading coefficient of P2=21(3x2−1) is 3/2, so subtract 76⋅32=74 times P2:
Compare with the closed forms: P4(0.3)=81(35(0.0081)−30(0.09)+3)=81(0.2835−2.7+3)=81(0.5835)=0.0729375 and P5(0.3)=81(63(0.00243)−70(0.027)+15(0.3))=81(0.15309−1.89+4.5)=81(2.76309)=0.34538625. The recurrence and the explicit polynomials agree to the last digit, and as rationals P5(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′/r and x=cosθ in the generating function, where θ is the angle between two position vectors of lengths r′<r. Since ∣r−r′∣2=r2+r′2−2rr′cosθ,
Riley, Hobson and Bence read off at once the potential of a point charge q at r′,
V(r)=4πε0rql≥0∑(rr′)lPl(cosθ),
whose l=0 term is the familiar q/(4πε0r). Integrating over a charge distribution turns the lth term into the lth 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 r and r′ exchanged is needed when the field point is the nearer one — a consequence of the ∣h∣<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 J2 (the equatorial bulge) and its successors are exactly the coefficients of this series.
Angular momentum and selection rules. With x=cosθ, the Legendre polynomials are the m=0 angular wavefunctions of any central-force problem: they are the θ-dependence of the s, pz, dz2,… orbitals, and the orthogonality relation of part 4 is what makes states of different l non-interacting. The three-term recurrence, rewritten as a statement about cosθPl, is the origin of the dipole selection rule Δl=±1: multiplying by cosθ can only move a state to its two neighbours in l. Extending this to m=0 is the business of Associated Legendre functions and Spherical harmonics.
Approximation, filtering and quadrature. Truncating the Legendre expansion is the optimal degree-N polynomial approximation to f in the L2[−1,1] norm, which makes it the natural basis for spectral-element discretisations of a one-dimensional domain. The zeros of Pn are the nodes of the n-point Gauss-Legendre rule (Numerical quadrature), and the fact that they are n distinct interior points is part 5 above. Hassani's expansion of the delta function, δ(x)=∑l22l+1Pl(0)Pl(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).
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.
Definition(The associated Legendre equation and Plm)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)−1−x2m2]y=0,(AL)
or in Sturm-Liouville form
dxd[(1−x2)dxdy]−1−x2m2y+ℓ(ℓ+1)y=0,
with p=1−x2, q=−m2/(1−x2), weight ρ=1, eigenvalue λ=ℓ(ℓ+1), natural interval [−1,1]. It reduces to Legendre's equation of Legendre polynomials when m=0, and it has the same three regular singular points x=−1,1,∞. Only m2 appears, so m and −m give the same equation. In applications m is an integer, x=cosθ, and −1≤x≤1. Throughout this unit ℓ is the parameter of the equation, so far unrestricted, while l is reserved for a non-negative integer; part 3 below is the statement that a solution bounded on the whole of [−1,1] forces ℓ to be such an integer.
For integers l≥0 and ∣m∣≤l, the associated Legendre function of the first kind is
for m≥0, the second form extending the definition verbatim to −l≤m<0. The superscript is an index, not a power. Since Pl has degree l, we have Plm≡0 for ∣m∣>l, and Pl0=Pl. The second, linearly independent solution of (AL) for integer l and m is the associated Legendre function of the second kind Qlm, built the same way from Ql; like Ql it is singular at x=±1 and is therefore discarded in every problem on a full sphere. The first few are
or, with x=cosθ so that (1−x2)1/2=sinθ: P11=−sinθ, P22=3sin2θ, P33=−15sin3θ.
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 Plm 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/dxm. Arfken, Weber and Harris insert the factor (−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 m among the positive-mspherical 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/dxm. 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 times ours, so where we have P11=−sinθ he has P11=+sinθ. Two of the three then put the (−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 Ylm agrees with Arfken, Weber and Harris's. Strauss does not: his spherical harmonic is Ylm(θ,φ)=Pl∣m∣(cosθ)eimφ (§10.3, p. 275), carrying neither the phase nor any normalising constant, so for odd m>0 his Ylm differs from the Condon-Shortley one in sign as well as in scale. See Spherical harmonics.
So the two conventions differ by (−1)m in Plm and, Strauss aside, agree in Ylm. Everything that is quadratic in Plm — both orthogonality relations and both normalisation constants — is the same in either convention, and so is the equation (AL).
For anything linear in the Plm there is a one-line rule. Write P~lm for the phase-free function, so that Plm=(−1)mP~lm — for negative m as well, since (−1)−m=(−1)m and both conventions use the same negative-m relation. Substituting this into an identity multiplies the term carrying upper index m by (−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 Plm 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 m and not for even m.
The step-in-l recurrence of part 8 carries the one upper index m 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-m relation of part 4 involves only −m and m, 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-m recurrence of part 8 mixes the upper indices m+1, m and m−1, so it changes for every m=0: 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 2mx term. (At m=0 that term is absent and the two remaining indices ±1 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−m is defined to equal Plm, which Riley, Hobson and Bence flag; with that choice m must be replaced by ∣m∣ throughout, as Strauss does in his definition of the spherical harmonics. We do not use it.
Definition(Properties of Plm)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≥0 and m be integers with ∣m∣≤l.
It solves the equation. If u solves Legendre's equation with parameter ℓ, then (1−x2)∣m∣/2d∣m∣u/dx∣m∣ solves (AL) with the same ℓ and with that m. In particular Plm solves (AL) with ℓ=l, and it is regular at x=±1.
Shape.Plm(x)=(1−x2)m/2Plm(x) for m≥0, where Plm is a polynomial of degree l−m. Hence Plm is a polynomial of degree l when m is even, and carries a fractional power of 1−x2 when m is odd; either way it is regular on [−1,1], and it vanishes at x=±1 unless m=0. Written in terms of these factors, the orthogonality of part 6 reads ∫−11PlmPkm(1−x2)mdx=0 for l=k, so at fixed m the Plm, l=m,m+1,…, are up to scale the orthogonal polynomials of the weight (1−x2)m on (−1,1); by General properties of orthogonal polynomials each of them therefore has exactly l−m zeros, all real, all simple and all interior to (−1,1).
Integrality is forced. Seek the factor P of part 2 as a Frobenius series about x=0, P(x)=∑j≥0ajxj+k with indicial exponent k and a0=0. Substituting it into the equation that P satisfies, namely (1−x2)P′′−2(m+1)xP′+[λ−m(m+1)]P=0, which is equation (D) of the proof of part 1 because (ℓ−m)(ℓ+m+1)=λ−m(m+1), gives the recurrence
aj+2=(j+1)(j+2)j2+(2m+1)j−λ+m(m+1)aj
for its coefficients, this display being the branch belonging to the indicial root k=0; the series diverges at x=±1 unless the numerator vanishes, terminating it. Arfken, Weber and Harris keep the two Frobenius branches apart. For k=0 the numerator must vanish at a non-negative evenj, and it does so at j=l−m when λ=l(l+1), which requires l to be an integer, at least m, and of the same parity asm. The second indicial root, k=1, supplies by the same argument the admissible l of the opposite parity to m. Taken together the two branches give: for each m the admissible l run from m upwards, one for every integer; equivalently, for each l the admissible m run from −l to l.
Negative m.Pl−m(x)=(−1)m(l+m)!(l−m)!Plm(x).
So Pl−m carries no new information; it is kept because the recurrences in m produce it.
Parity and special values.Plm(−x)=(−1)l+mPlm(x),Plm(±1)=0(m=0),
and at the centre
Plm(0)=⎩⎨⎧(l−m)!!(−1)(l+m)/2(l+m−1)!!,0,l+m even,l+m odd.
Also Pmm(x)=(−1)m(2m−1)!!(1−x2)m/2 and Pm+1m(x)=(−1)m(2m+1)!!x(1−x2)m/2.
Orthogonality in l at fixed m, with unit weight.∫−11Plm(x)Pkm(x)dx=2l+12(l−m)!(l+m)!δlk,
or with x=cosθ,
∫0πPlm(cosθ)Pkm(cosθ)sinθdθ=2l+12(l−m)!(l+m)!δlk.
Consequently, at fixed m≥0, the normalised functions 22l+1(l+m)!(l−m)!Plm with l≥m are orthonormal in L2(−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) has
f=l≥m∑alPlm,al=22l+1(l+m)!(l−m)!∫−11f(x)Plm(x)dx,
the series converging in L2. The sum starts at l=m because Plm=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.
Orthogonality in m at fixed l, with weight (1−x2)−1.∫−11Plm(x)Plk(x)1−x2dx=m(l−m)!(l+m)!δmk(m,k≥1).
Both books that state this relation add that it is of no practical use, because in spherical coordinates the φ boundary condition already separates the different m.
Generating function and recurrences. For m≥0,
2mm!(1−2hx+h2)m+1/2(2m)!(1−x2)m/2=n=0∑∞P~n+mm(x)hn,
where P~ denotes the phase-free functions of Riley, Hobson and Bence (so P~lm=(−1)mPlm); 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 of part 2 rather than the functions themselves: their left-hand side carries (−1)m(2m−1)!! in place of our (2m)!(1−x2)m/2/2mm!, and (2m−1)!!=(2m)!/2mm!, so the two agree once the common factor (1−x2)m/2 and the phase are accounted for. The two most useful recurrences are
(2l+1)xPlm=(l+m)Pl−1m+(l−m+1)Pl+1m,Plm+1+(1−x2)1/22mxPlm+(l+m)(l−m+1)Plm−1=0,
the first stepping in l at fixed m ((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 m at fixed l ((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 m with the definition adopted here.
Idea of the proofOriginal
Everything comes from the single observation that differentiating Legendre's equation m times almost gives the associated equation. Differentiating (1−x2)u′′−2xu′+ℓ(ℓ+1)u=0m times with Leibniz produces
(1−x2)v′′−2x(m+1)v′+(ℓ−m)(ℓ+m+1)v=0,v=dxmdmu,
which differs from (AL) only in the first-derivative coefficient and the eigenvalue. The factor (1−x2)m/2 is exactly the change of dependent variable that repairs both: it trades the extra −2mxv′ for the term −m2y/(1−x2). 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 be an integer at least ∣m∣.
The orthogonality in l (part 6) is again the Sturm-Liouville argument: p=1−x2 vanishes at both ends, so the boundary terms die, and the q term is the same for both functions and cancels on subtraction. The orthogonality in m (part 7) is the same equation read as an eigenvalue problem in −m2 with weight (1−x2)−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+m times so that all boundary terms vanish, and then notice that of all the terms Leibniz produces in differentiating (1−x2)mdl+m(x2−1)l/dxl+m a further l+m times, exactly one has the degrees right to survive. That term carries the factor (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≥0; the case m<0 then follows because (AL) contains only m2.
Step 1: differentiate Legendre's equation m times. Let u solve (1−x2)u′′−2xu′+ℓ(ℓ+1)u=0 and set v=dmu/dxm. Apply dm/dxm term by term with Leibniz's rule. The factor 1−x2 contributes three terms, −2x contributes two, and the constant one:
Step 3: substitute into (D). Write σ=1−x2 for brevity. Putting the two displays into σv′′−2x(m+1)v′+(ℓ−m)(ℓ+m+1)v=0 and cancelling the common factor σ−m/2, which is nonzero for ∣x∣<1,
The coefficient of y′ collapses to −2x. In the coefficient of y, the two fractions combine with m(m+2)−2m(m+1)=−m2 into −m2x2/σ, and since σm2x2=σm2−m2,
m−σm2x2=m+m2−σm2=m(m+1)−σm2.
Adding the last piece and using (ℓ−m)(ℓ+m+1)=ℓ(ℓ+1)−m(m+1), the coefficient of y is ℓ(ℓ+1)−m2/σ. Altogether
(1−x2)y′′−2xy′+[ℓ(ℓ+1)−1−x2m2]y=0,
which is (AL).
Step 4: specialise and check regularity. Take u=Pl, ℓ=l. Then y=(1−x2)m/2dmPl/dxm, and multiplying by the constant (−1)m — which (AL) permits, being linear — gives Plm. For part 2, dmPl/dxm is a polynomial of degree l−m (zero if m>l), so Plm=(−1)m(1−x2)m/2P with degP=l−m. The factor (1−x2)m/2 is continuous on [−1,1] and vanishes at ±1 for m≥1, so Plm is regular on the closed interval and vanishes at the ends unless m=0. For part 5, each differentiation flips parity and (1−x2)m/2 is even, so Plm has the parity of l+m. □
Proof(Parts 4, 6 and 7: negative m, 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 m. Take m≥0 and work on (−1,1), where 1−x>0 and 1+x>0 so that every fractional power below is unambiguous. Split (x2−1)l=(x−1)l(x+1)l and apply Leibniz to the (l+m)-fold derivative in the definition. The term with r derivatives on (x−1)l is
(rl+m)(l−r)!l!(x−1)l−r⋅(r−m)!l!(x+1)r−m,
nonzero only when r≤l (first factor) and l+m−r≤l, that is r≥m (second). Multiply by (1−x2)m/2=(1−x)m/2(1+x)m/2 and write (x−1)l−r=(−1)l−r(1−x)l−r:
Now repeat with −m in place of m, so that the definition reads Pl−m=2ll!(−1)−m(1−x2)−m/2dl−m(x2−1)l/dxl−m. Leibniz now gives terms indexed by s, with s derivatives on (x−1)l:
Multiplying by (1−x)−m/2(1+x)−m/2 and again replacing (x−1)l−s by (−1)l−s(1−x)l−s, the term carries (1−x)l−s−m/2(1+x)s+m/2. Substitute s=r−m, so that r again runs from m to l: the powers become
identical to those in (†), and the sign becomes (−1)l−s=(−1)l−r+m. The factorials become
(r−m)!(l−r)!(l−m)!⋅(l−r+m)!r!(l!)2,
which is the same set {r!,(l+m−r)!,(l−r)!,(r−m)!} as in (†) with (l+m)! replaced by (l−m)!. Collecting the prefactors, (−1)−m(−1)l−r+m=(−1)l−r, whereas (†) carries (−1)m(−1)l−r. Therefore every term of Pl−m equals the corresponding term of Plm times the single constant
(l+m)!(l−m)!⋅(−1)m1=(−1)m(l+m)!(l−m)!,
and summing,
Pl−m=(−1)m(l+m)!(l−m)!Plm.
Hassani obtains the same relation from the ladder-operator construction, by building Yl,−m from the bottom state with L+ instead of from the top state with L− and comparing.
Part 6, orthogonality in l. Fix m and let l=k. Both Plm and Pkm satisfy the Sturm-Liouville form of (AL) with the samep=1−x2 and the sameq=−m2/(1−x2), and different eigenvalues l(l+1), k(k+1). Multiply the equation for Plm by Pkm, integrate over [−1,1], and integrate the p-term by parts. The boundary contribution is
[(1−x2)PkmdxdPlm]−11=0,
because 1−x2 vanishes at both ends while Pkm and dPlm/dx stay bounded there — for m≥1 the derivative of (1−x2)m/2 can blow up like (1−x2)m/2−1, but 1−x2 times it still tends to 0, and Pkm itself vanishes at the ends. Subtracting the same identity with l and k exchanged, the symmetric integral ∫(1−x2)(Pkm)′(Plm)′ and the q-integral ∫qPlmPkm both cancel, leaving
[l(l+1)−k(k+1)]∫−11PlmPkmdx=0,
hence the integral vanishes. Note that m must be the same in both factors for the q-terms to cancel.
The normalisation constant. Take l=k and write R=x2−1. By the definition,
the two factors (−1)m having squared away. Integrate by parts l+m times, differentiating (1−x2)mdl+mRl/dxl+m and integrating the other factor. The boundary term produced by the (i+1)st of these integrations, i=0,…,l+m−1, is
and it vanishes for either of two reasons according to the size of i. For i≤m−1: expanding the first factor by Leibniz, every term differentiates (1−x2)m=(−R)m at most i≤m−1 times and so still carries a positive power of 1−x2. For i≥m: the second factor differentiates Rl=(x−1)l(x+1)l only l+m−1−i≤l−1 times, which leaves at least one undifferentiated factor x−1 and one x+1. Either way the product is 0 at x=±1. Thus
Expand the inner (l+m)-fold derivative by Leibniz. A term differentiates (1−x2)mr times and dl+mRl/dxl+m a further l+m−r times. The first is nonzero only for r≤2m; the second needs 2l+2m−r≤2l, that is r≥2m. So r=2m is the only surviving term, and it contributes
At m=0 this is the Legendre value 2/(2l+1), as it must be. The θ-form follows from x=cosθ, dx=−sinθdθ.
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 Plm span. That step is done separately below.
Part 7, orthogonality in m. Read (AL) the other way: for fixed l, the family {Plm}m consists of eigenfunctions of
dxd[(1−x2)dxdy]+l(l+1)y=+m21−x21y,
a Sturm-Liouville problem with p=1−x2, q=l(l+1), eigenvalue −m2 and weight ρ=(1−x2)−1; the natural interval is again [−1,1] and the Plm are regular there. Distinct eigenvalues force orthogonality with respect to that weight. Since the eigenvalue is −m2, "distinct" means ∣m∣=∣k∣, and that is the hypothesis under which Riley, Hobson and Bence state the vanishing, their eq. (18.39); m=−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≥1 in part 7 belongs to the constant, not to the vanishing. The value (l+m)!/[m(l−m)!], which Arfken, Weber and Harris and Riley, Hobson and Bence both record, blows up as m→0, and rightly so: at m=0 the integral itself diverges, because Pl0(±1)=(±1)l=0 leaves the integrand Pl2/(1−x2) behaving like [2(1−x)]−1 at the upper end. For m≥1 the factor (1−x2)m in (Plm)2 more than cancels the weight and the integral converges. Both books add that the relation is not used in practice. □
Lemma(Part 6: totality of {Plm}l≥m at fixed m)Original
Fix an integer m≥0 and suppose f∈L2(−1,1) satisfies ∫−11fPlmdx=0 for every l≥m. We show f=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/2Plm with degPlm=l−m, and the leading coefficient is nonzero because dmPl/dxm has degree exactly l−m. As l runs over m,m+1,m+2,… the degrees l−m run over 0,1,2,…, one of each, so span{Plm:l≥m} is the whole space of polynomials. Put g=(1−x2)m/2f. The hypothesis says ∫−11gPlmdx=0 for every l≥m, hence ∫−11gqdx=0 for every polynomial q.
Step 2: use the Legendre case. The factor (1−x2)m/2 is bounded by 1 on (−1,1), so g∈L2(−1,1) with ∥g∥≤∥f∥. Taking q=Pk for each k≥0, g is orthogonal to every Legendre polynomial, and that system is total in L2[−1,1] (Legendre polynomials, part 4, where it is proved from the Weierstrass approximation theorem). Therefore g=0 almost everywhere.
Step 3: return to f.(1−x2)m/2>0 at every point of the open interval (−1,1), so g=0 a.e. forces f=0 a.e. there. □
Remark(Where the hypotheses are used)Original
∣m∣≤l and l an integer. These are not conventions but the only way to get a solution bounded on [−1,1], by part 3. Arfken, Weber and Harris note in a footnote the natural hope that the prefactor (1−x2)m/2, which vanishes at the ends, might tame a divergent series P; they state that it does not. For ∣m∣>l the function Plm is not a small solution but identically zero, which is why the expansion in part 6 starts at l=m.
The same m in both factors of part 6. The q-term −m2/(1−x2) cancels on subtraction only if it is identical in the two equations. With different m's there is a leftover (k2−m2)∫PlmPl′k(1−x2)−1 and the relation fails; what survives instead is part 7, with its different weight. Riley, Hobson and Bence are explicit that m must match.
Vanishing of p=1−x2 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≥1: dPlm/dx is unbounded at x=±1 when m=1 (it behaves like (1−x2)−1/2), and what vanishes is the product (1−x2)Pkm(Plm)′, in which Pkm supplies another factor (1−x2)m/2. On a spherical cap θ∈[0,θ0] with θ0<π the boundary term at θ0 does not vanish and the Plm are not orthogonal; the correct eigenfunctions there solve (AL) with non-integer ℓ. Strauss sets the diffusion equation in the solid cone {r<a,θ<α} 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(cosθ) satisfies the associated Legendre equation, expand p in powers of s−1, and write down the equations that determine the eigenvalues. It stops there and does not say what those eigenvalues are; the condition p(cosα)=0 is what pushes ℓ off the integers for a general half-angle α.
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=−sinθ with Riley, Hobson and Bence's P11=+sinθ: both books are internally consistent, and both produce the same Y11.
Reality. All of Plm, (AL) and both weights are real, so nothing above needs complex conjugation. Conjugation appears only once the azimuthal factor eimφ is attached, in Spherical harmonics.
Method(Getting Plm 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
For a single value, small l and m: differentiate Plm times, multiply by (1−x2)m/2, and attach (−1)m.
For a whole table, or for large l: climb in l at fixed m, 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,
the second being the l=m case of the recurrence with the Pm−1m=0 term dropped, and then iterate
Pl+1m=l−m+1(2l+1)xPlm−(l+m)Pl−1m.
Remember that any Pl′m with l′<m is zero, so terms containing one are simply omitted.
To raise m instead: use the second recurrence of part 8, solved for Plm+1.
For negative m: do not differentiate a negative number of times; use part 4.
For any integral ∫−11PlmPkmdx: do not integrate. It is 0 unless l=k, and then it is 2l+12(l−m)!(l+m)!.
To normalise for quantum mechanics: the combination that is a unit vector in L2(0,π;sinθdθ) is
Θlm(θ)=22l+1(l+m)!(l−m)!Plm(cosθ),
which is the θ half of a spherical harmonic.
Example(P32 from three routes, and its norm)Original
Step 1: from the definition.P3=21(5x3−3x), so P3′=21(15x2−3) and P3′′=15x. With m=2 the phase is (−1)2=+1 and
P32(x)=(1−x2)1⋅15x=15x(1−x2).
Step 2: from the Rodrigues form. Here l+m=5 and (x2−1)3=x6−3x4+3x2−1. Only the leading term survives five differentiations: d5x6/dx5=6!x/1!=720x, while x4, x2 and the constant are annihilated. Then
Step 3: from the recurrence. Start at the top: P22=(−1)2(3)!!(1−x2)=3(1−x2), since (2⋅2−1)!!=3!!=3. Then the l=m=2 case of the step-in-l recurrence, with P12=0 dropped, gives
P32=2−2+1(2⋅2+1)xP22−0=5x⋅3(1−x2)=15x(1−x2).
Three routes, one answer.
Step 4: the normalisation integral, by hand. Part 6 predicts
Over 105: 32=10570, 54=10584, 72=10530, so the bracket is 10570−84+30=10516 and the integral is 225⋅10516=1053600=7240=34.2857… — exactly the predicted value.
Step 5: two more checks. The m=1 case: P31=−23(5x2−1)(1−x2)1/2, and
∫−11(P31)2dx=49∫−11(5x2−1)2(1−x2)dx=724,
against the formula 72⋅2!4!=72⋅12=724. And the second orthogonality of part 7 at l=3, m=1:
since 10−320+2=316; and this matches m(l−m)!(l+m)!=1⋅2!4!=12. Also ∫−11P31P32(1−x2)−1dx=0 by parity, since P31 is even and P32 is odd — consistent with part 5, (−1)3+1=+1 and (−1)3+2=−1. Every integral in this example was checked symbolically.
ApplicationOriginal
Anything separated in spherical coordinates. Whenever ∇2 is separated in spherical polars without assuming axial symmetry, the polar-angle factor is forced to be Plm(cosθ) and the azimuthal factor eimφ (or cosmφ, sinmφ). The integer m comes from single-valuedness in φ; the integer l≥∣m∣ comes from part 3; and the pair (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) commutes with the angular-momentum operators, and Plm(cosθ)eimφ is the simultaneous eigenfunction: l is the orbital quantum number, m the magnetic one, and the (2l+1)-fold degeneracy in m is the statement that l fixes the radial equation while m 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. Hassani's construction, which produces the angular eigenfunctions by lowering from the top state with L−, shows what the ladder route does and does not settle: it fixes the relative phases along a ladder of fixed l, and it delivers the phase-free Plm of his eq. (13.35), with the (−1)m then put into Ylm 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(cosθ){cosmφ,sinmφ} — the Gauss coefficients glm, hlm of the International Geomagnetic Reference Field, and the Stokes coefficients Clm, Slm of a gravity model. Here l measures the spatial scale and m 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, which is Plm times (2−δm0)(2l+1)(l−m)!/(l+m)!, because the raw (l+m)!/(l−m)! in part 6 overflows at the degree l∼2000 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(cosθ) 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∣ nodal circles of latitude from the zeros of the polynomial factor, and ∣m∣ nodal meridians from the azimuthal factor.
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.
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) be spherical polar angles on the unit sphere S2: θ∈[0,π] the polar angle from the z-axis, φ∈[0,2π) the azimuth. Write dΩ=sinθdθdφ=dφd(cosθ) for the element of solid angle, so that ∫S2dΩ=4π, and
(f,g)=∫S2f(θ,φ)g(θ,φ)dΩ
for the inner product of L2(S2).
For integers l≥0 and −l≤m≤l, the spherical harmonic is
Ylm(θ,φ)=4π2l+1(l+m)!(l−m)!Plm(cosθ)eimφ,
with Plm the associated Legendre function of Associated Legendre functions, including its Condon-Shortley phase (−1)m. There are 2l+1 harmonics of each order l. This is Arfken, Weber and Harris's eq. (15.137) verbatim; books whose Plm is phase-free display an extra (−1)m in front and arrive at the same function, as the remark below sets out. Equivalently Ylm=Θlm(θ)Φm(φ) with the two separately normalised factors of Arfken, Weber and Harris,
which is why they appear in every separation of ∇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)Ylm and LzYlm=mYlm, built by applying the lowering operator L− repeatedly to the top state Yll∝sinlθeilφ.
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 Ylm 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 P. Arfken, Weber and Harris put the (−1)m into Plm (their eq. (15.78)) and so write Ylm=4π2l+1(l+m)!(l−m)!Plm(cosθ)eimφ, eq. (15.137) — our definition.
The phase in Y. Riley, Hobson and Bence and Hassani keep Plm phase-free (eqs. (18.32) and (13.35) respectively; Hassani's P11=+sinθ, where ours is −sinθ) and display the factor in the harmonic instead: Ylm=(−1)m4π2l+1(l+m)!(l−m)!Plm(cosθ)eimφ, 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 sameYlm: the (−1)m sits in a different factor, not in the product. The tables agree, including Y11=−3/8πsinθeiφ. 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∣(cosθ)eimφ 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 sin2θcos2φ=(x2−y2)/r2.
A fourth variant, the real spherical harmonics, replaces eimφ by cosmφ and sinmφ. Arfken, Weber and Harris use these in the geodesy application, writing Ymle=Plm(cosθ)cosmφ and Ymlo=Plm(cosθ)sinmφ after Morse and Feshbach — note that those authors put the m 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-like and so phase-blind). What is convention-dependent: every entry with odd m in a table of Plm (not of Ylm, 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, with no (−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 Ylm; 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′≥0 and ∣m∣≤l, ∣m′∣≤l′ be integers, and write Ω=(θ,φ).
Conjugation and parity.Ylm(Ω)=(−1)mYl−m(Ω),Ylm(π−θ,φ+π)=(−1)lYlm(θ,φ).
Values on the axis.Ylm(0,φ)=4π2l+1δm0,Ylm(π,φ)=(−1)l4π2l+1δm0,Yl0=4π2l+1Pl(cosθ).
Completeness (the Laplace series). The family {Ylm} is a complete orthonormal system in L2(S2): every f∈L2(S2) has
f(Ω)=l=0∑∞m=−l∑lalmYlm(Ω),alm=∫S2Ylm(Ω)f(Ω)dΩ,
the series converging in L2(S2), with ∥f∥2=∑l,m∣alm∣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 f 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 L2 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 f and are not proved here.
Closure. As an identity of distributions on S2×S2,
l=0∑∞m=−l∑lYlm(Ω)Ylm(Ω′)=δ(Ω−Ω′),
where δ(Ω−Ω′) is the Dirac delta on the sphere: δ(Ω−Ω′)=0 for Ω=Ω′ and ∫S2δ(Ω)dΩ=1.
Addition theorem. Let Ω=(θ,φ) and Ω′=(θ′,φ′) be two directions separated by the angle γ, so that
cosγ=cosθcosθ′+sinθsinθ′cos(φ−φ′).
Then for every l≥0,
Pl(cosγ)=2l+14πm=−l∑lYlm(Ω)Ylm(Ω′).
Equivalently, in terms of associated Legendre functions,
Pl(cosγ)=m=−l∑l(l+m)!(l−m)!Plm(cosθ)Plm(cosθ′)eim(φ−φ′).
Solid harmonics.rlYlm(θ,φ) is a harmonic function on R3 and, in fact (Strauss), a homogeneous polynomial of degree l in x,y,z. The general solution of ∇2u=0 regular at the origin is ∑l,mAlmrlYlm, and the general solution regular at infinity is ∑l,mBlmr−l−1Ylm.
Idea of the proofOriginal
Parts 1 to 3 are bookkeeping: the sphere's measure factorises as dφd(cosθ), so the double integral splits into the φ-integral ∫02πei(m−m′)φdφ=2πδmm′ and, once m=m′, the associated Legendre orthogonality already proved. The constant under the square root is exactly what makes the product come out as 1.
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′) 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, because Pl(cosγ) is a complete system for functions of one angle. Computing those coefficients needs only Pl(1)=1, and gives (2l+1)/4π. So the two expansions are
l∑[m∑Ylm(Ω)Ylm(Ω′)]=l∑4π2l+1Pl(cosγ),
two sums over l with equal totals.
Second step. Equal totals are not equal terms. What makes the sums match term by term is that each l-block is rotation invariant: the 2l+1 functions Ylm of a fixed l span a space that a rotation of the axes maps into itself and never mixes with another l. Choose the new polar axis along Ω′; then Pl(cosγ) is a multiple of Yl0 in the new frame, hence a combination of the Ylm of the old frame with the same l. So the lth term on the right lives in the lth block, matching the lth 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=dφd(cosθ) and put x=cosθ. Then
with Nlm the square root in the definition, which is real so no conjugation touches it; the Plm are real as well. The φ-integral is 2πδmm′ — this is where the azimuthal orthogonality lives, and it is the reason m=m′ may be assumed in the x-integral. With m=m′ the x-integral is part 6 of Associated Legendre functions, namely 2l+12(l−m)!(l+m)!δll′. Hence
The product of the two m-dependent factors is (−1)m4π2l+1(l+m)!(l−m)!=(−1)mNlm, so Yl−m=(−1)mYlm, which rearranges to the stated relation because (−1)m=(−1)−m. Hassani records exactly this as his eq. (13.37). For the parity statement, the antipodal map is θ↦π−θ, φ↦φ+π, that is x↦−x and eimφ↦(−1)meimφ; combining with Plm(−x)=(−1)l+mPlm(x) gives the factor (−1)l+m(−1)m=(−1)l, with no m left.
Part 3. At θ=0 the azimuth is meaningless, so any harmonic with a genuine φ-dependence must vanish there; concretely Plm(±1)=0 for m=0, leaving only m=0. For m=0, Nl0=(2l+1)/4π and Pl(1)=1, Pl(−1)=(−1)l, which gives the two displayed values. The identity Yl0=(2l+1)/4πPl(cosθ) is the same statement without evaluating θ; 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:
The bracket therefore reproduces f at Ω when integrated against f, for every f; that is the defining property of δ(Ω−Ω′). (Read as an identity of distributions: the double series does not converge pointwise, exactly as ∑neinθ/2π does not. Riley, Hobson and Bence derive the relation in this way and use it immediately in the next proof.) □
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′ and let γ be the polar angle of Ω measured from Ω′.
Step 1: expand the delta function in Legendre polynomials of γ. The function δ(Ω−Ω′) can only depend on Ω and Ω′ through their relative position, which on a sphere means through γ alone. Expanding a function of the single variable cosγ∈[−1,1] in Legendre polynomials (part 4 of Legendre polynomials),
Step 2: evaluate bl. The integrand does not depend on the azimuth ψ about the axis Ω′, so multiplying by 2π/2π and inserting ∫02πdψ=2π turns the one-dimensional integral into an integral over the whole sphere in the coordinates (γ,ψ):
Both sides are sums over l. It remains to show that the lth summands are separately equal; that is the whole content of the theorem, and it does not follow from (♯) alone.
Step 4: each l-block is closed under rotations. Rotate the coordinate axes rigidly and call the new angular coordinates Ωˉ (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), as a function of direction, is again an eigenfunction with eigenvalue −l(l+1). The eigenspace for that eigenvalue is spanned by the 2l+1 functions Ylm′(Ω) of the samel and no other, because distinct l give distinct eigenvalues −l(l+1). Therefore
Ylm(Ωˉ)=m′=−l∑lDm′mlYlm′(Ω)
for coefficients Dm′ml 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(cosγ) inside the l-block. Choose the rotation that sends the polar axis to Ω′. In the new frame the polar angle of Ω is exactly γ, so by part 3,
So, as a function of Ω for fixed Ω′, the lth term on the right of (♯) lies in the span of {Ylm′(Ω)}m′; and the lth term on the left obviously does too. Since the spans for distinct l are mutually orthogonal in L2(S2) by part 1, the two sums in (♯) can be matched term by term: project both sides onto the l-block. This gives
m=−l∑lYlm(Ω)Ylm(Ω′)=4π2l+1Pl(cosγ)
for each l separately, which is the addition theorem.
Step 6: the P-form. Insert the definition and use eimφ′=e−imφ′:
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. □
Remark(Where the hypotheses are used)Original
m integral, and ∣m∣≤l. Single-valuedness of the solution in φ forces eim(φ+2π)=eimφ, hence m∈Z; and regularity on the polar axis forces l∈Z≥0 with l≥∣m∣ (part 3 of Associated Legendre functions). Hassani is careful about the first point: in classical physics the function must agree at φ and φ+2π, which gives integer m; in quantum mechanics only the absolute value of the function is measurable, so on that criterion alone m could also be a half-integer. He goes no further with it, and neither do we: his Box 13.4.1 assumes m 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 Ylm still excludes them for orbital angular momentum.)
Both factors normalised, and Plm real. Part 1 needs the constant Nlm to be real so that it is unaffected by the conjugation, and the x-integral to be the unit-weight orthogonality relation. The second orthogonality of the associated Legendre functions, the one with weight (1−x2)−1, plays no part here: the φ-integral has already separated the different m, 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 (♯), which is far weaker: it would be consistent with the l-terms being unequal and compensating. Arfken, Weber and Harris make the same point from the other direction, observing that the three Y1m are proportional to z/r and (x±iy)/r, so that a rotation of the axes can only produce combinations of them — never a Y2m.
Ω and Ω′ enter only through γ (Step 1). This is what licenses a one-variable expansion of a two-variable object. It holds because δ(Ω−Ω′) 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.
L2 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 f, for which they refer to Hobson rather than proving it; Strauss states the theorem for square-integrable f, which is the version we take. A discontinuous f — say a hemisphere at one potential and the other at another, Strauss's Exercise 7 — gives a series that converges in L2 and oscillates near the equator, exactly as a Fourier series does at a jump.
Finitely many terms per l. The inner sum in part 4 is finite (2l+1 terms) and this is used tacitly in Step 5, where "the l-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) on a sphere, or to solve a boundary-value problem with that data:
Reduce by symmetry first. If g is independent of φ, only m=0 survives and the expansion is in Pl(cosθ). If g depends on φ only through cosm0φ or sinm0φ, only m=±m0 survive. If g is a polynomial in x,y,z restricted to the sphere, the expansion terminates, by part 7.
Identify, do not integrate. Write g in terms of cosθ, sinθ and e±imφ and match it against the table of Ylm. This is exact and much faster than evaluating ∫YlmgdΩ.
Otherwise project.alm=∫02π∫0πYlm(θ,φ)g(θ,φ)sinθdθdφ. Do the φ-integral first: it is a Fourier coefficient and kills all but one m.
For the interior Dirichlet problem∇2u=0 in r<a with u=g on r=a: discard the r−l−1 solutions, which are singular at the origin, and write
u(r,θ,φ)=l,m∑alm(ar)lYlm(θ,φ),
which matches g at r=a term by term. For the exterior problem, keep r−l−1 instead and use (a/r)l+1.
For the Helmholtz or Schrödinger equation, replace rl by the appropriate radial solution — jl(kr) and yl(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.
To collapse a double sum over m that has the structure ∑mYlm(Ω)Ylm(Ω′), use the addition theorem backwards: it becomes 4π2l+1Pl(cosγ), 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 inside the ball r<a with
u(a,θ,φ)=sin2θcos2φ.
Step 1: expand the data. Write cos2φ=21(e2iφ+e−2iφ). From the table, Y2±2=15/32πsin2θe±2iφ, so
Only l=2, m=±2 occur: by step 1 of the method this had to happen, since g carries cos2φ and is a quadratic polynomial on the sphere.
Step 2: check the coefficient by projection. By part 4, a22=∫S2Y22gdΩ. Substituting the expansion and using orthonormality, a22=c, and likewise a2,−2=c. As an independent check, Parseval must give
the second term vanishing because cos(φ−φ′)=cos(−π/2)=0. Now evaluate both sides for l=2. The right-hand side of the theorem is a sum of five terms; computing them from the table gives
54πm=−2∑2Y2m(Ω)Y2m(Ω′)=−0.3125000,
while the left-hand side is
P2(cosγ)=23cos2γ−1=23⋅81−1=20.375−1=−0.3125
exactly, since cos2γ=2/16=1/8. The same test at l=1 gives 0.3535534 on both sides (P1(cosγ)=cosγ), and at l=3 gives −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),
with M the mass, R the equatorial radius. The coefficient C20 measures the equatorial bulge and C22, S22 the azimuthal asymmetry. Their measured values illustrate how the shapes differ: C20=1.083×10−3 for the Earth, (0.200±0.002)×10−3 for the Moon, (1.96±0.01)×10−3 for Mars. Every satellite geodesy mission, and every orbit propagator, is an exercise in estimating these numbers; the l=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
in which the two directions are separated: the source variables (r′,Ω′) and the field variables (r,Ω) no longer appear in a single entangled factor Pl(cosγ). 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 N-body potential evaluation from O(N2) to O(N) work.
Quantum mechanics.Ylm is the angular part of every central-field wavefunction: the atomic orbitals s,p,d,f are l=0,1,2,3, the (2l+1)-fold degeneracy in m 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 cosθ gives at once
which is the electric-dipole selection rule Δl=±1, Δm=0 for z-polarised light: the integral vanishes unless l 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=2l+11∑m∣alm∣2 of its Laplace series, and the rotation invariance of each l-block (Step 4 above) is precisely why Cl, and not the individual alm, 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 L2 error, and the truncation is rotation-covariant.
The polynomials generated by e−t2+2tx: their Rodrigues formula, three-term recurrence, Hermite equation and orthogonality with weight e−x2 on the whole line, the Hermite functions they build, and the two incompatible conventions (physicists' Hn, probabilists' Hen).
The Hermite polynomialsH0,H1,H2,… are the coefficients in the expansion of
g(x,t)=e−t2+2tx=n=0∑∞Hn(x)n!tn,
which converges for every x∈R and every t∈C, because t↦g(x,t) is entire. Equivalently, Hn(x)=∂tng(x,t)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)nex2dxndne−x2,
and the two agree: the statement below derives each from the other. In the language of General properties of orthogonal polynomials, Hn is the classical family belonging to the weight w(x)=e−x2 on (−∞,∞), with the scale fixed by the choice Hassani writes as Kn=(−1)n. That scale is a convention and nothing more; it makes the leading coefficient of Hn equal to 2n rather than 1.
The first few are
H0H3=1,=8x3−12x,H1H4=2x,=16x4−48x2+12,H2H5=4x2−2,=32x5−160x3+120x.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),
which is no longer a polynomial. Kreyszig and Reed and Simon use the normalised version, which we write
ψn(x)=(2nn!π)−1/2e−x2/2Hn(x),n=0,1,2,…
Kreyszig obtains exactly these functions by running Gram-Schmidt (Orthonormal bases and Gram-Schmidt) on the sequence tne−t2/2 in the real space L2(−∞,∞); Reed and Simon obtain them instead by repeated application of A†=2−1/2(x−d/dx) to the Gaussian ψ0=π−1/4e−x2/2, as ψn=(n!)−1/2(A†)nψ0. 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 Hn be defined by the generating function above and let ψn be the normalised Hermite functions. Write H−1≡0.
Explicit form. For n≥0,
Hn(x)=s=0∑⌊n/2⌋(n−2s)!s!(−1)sn!(2x)n−2s.
In particular Hn is a polynomial of degree exactly n with leading coefficient 2n, the coefficient of xn−1 vanishes, and
Hn(−x)=(−1)nHn(x),H2m(0)=(−1)mm!(2m)!,H2m+1(0)=0.
Rodrigues formula.Hn(x)=(−1)nex2dxndne−x2.
Recurrences. For n≥0,
Hn+1(x)=2xHn(x)−2nHn−1(x),Hn′(x)=2nHn−1(x).
Hermite's equation.Hn solves
y′′−2xy′+2ny=0.
Riley, Hobson and Bence state this equation with a general real parameter ν in place of n; it is exactly when ν is a non-negative integer that it has a polynomial solution.
Orthogonality and norm. For all m,n≥0,
∫−∞∞e−x2Hm(x)Hn(x)dx=2nn!πδmn.
Equivalently the ψn are orthonormal in L2(R) with unit weight.
The eigenvalue problem solved. Each ψn satisfies
−ψn′′(x)+x2ψn(x)=(2n+1)ψn(x),
which is the self-adjoint form Arfken, Weber and Harris write as φn′′+(2n+1−x2)φn=0 and Reed and Simon as (−d2/dx2+x2)ϕn=(2n+1)ϕn.
Completeness.{ψn}n≥0 is an orthonormal basis of L2(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) 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) has the mean-square convergent expansion f=∑n(f,ψn)ψn, which Riley, Hobson and Bence write in the unnormalised form f=∑nanHn with
an=2nn!π1∫−∞∞f(x)Hn(x)e−x2dx.
Idea of the proofOriginal
Everything follows from one algebraic accident. The exponent −t2+2tx completes the square two different ways: as x2−(t−x)2, and as a product of two exponentials e2txe−t2 whose separate series are known.
The first completion makes t and x interchangeable up to a sign, because ∂te−(t−x)2=−∂xe−(t−x)2; differentiating n times in t and setting t=0 therefore turns into differentiating n times in x, which is the Rodrigues formula. The second completion multiplies out to the explicit coefficient sum, and the sum shows at a glance that Hn has degree n, 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 and ∂xg=2tg that the exponential satisfies, read coefficient by coefficient; eliminating Hn−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 s and t, by the weight e−x2, and the x-integral is a Gaussian that evaluates to πe2st. A function of the product st alone has no smtn term with m=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∑j!(2x)jtj,e−t2=s≥0∑s!(−1)st2s,
and collect the terms with tn, which are those with j+2s=n, that is 0≤s≤⌊n/2⌋ and j=n−2s:
g(x,t)=n≥0∑tns=0∑⌊n/2⌋(n−2s)!s!(−1)s(2x)n−2s.
Comparing with ∑nHntn/n! and multiplying by n! gives the stated sum. The rearrangement is legitimate because both series converge absolutely for every x and t.
Reading the sum: the s=0 term contributes (2x)n and no other term reaches degree n, so the degree is n and the leading coefficient 2n; every exponent n−2s has the parity of n, which gives both the absence of an xn−1 term and the parity relation Hn(−x)=(−1)nHn(x); and at x=0 the only surviving term is the one with n−2s=0, which exists only for even n=2m and equals (−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,
Item 2, the Rodrigues formula. Complete the square in the exponent the other way:
−t2+2tx=x2−(t−x)2,sog(x,t)=ex2e−(t−x)2.
The function e−(t−x)2 depends on t and x only through t−x, so ∂t acting on it equals −∂x acting on it. Applying this n times, and noting that ex2 is a constant for ∂t,
where the last step sets t=0 before differentiating in x, which is allowed because e−(t−x)2 is jointly smooth and the t-derivatives have already been taken.
Item 3, the recurrences. The exponential satisfies
∂t∂g=(2x−2t)g,∂x∂g=2tg,
by direct differentiation. Insert the series into the first identity. On the left, ∂t∑nHntn/n!=∑n≥0Hn+1tn/n! after shifting the index. On the right,
where the second sum was re-indexed using tn+1/n!=n′tn′/n′! with n′=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 t agree only if their coefficients agree, so Hn+1=2xHn−2nHn−1 for n≥1; for n=0 the second sum is empty and the identity reads H1=2xH0, which is the same statement with the convention H−1=0.
The second identity gives, in the same way,
n≥0∑Hn′n!tn=2n≥1∑nHn−1n!tn,
hence Hn′=2nHn−1 for n≥1; at n=0 the right-hand sum is empty and the identity reads H0′=0, which is again the stated relation with the convention H−1=0. Both recurrences therefore hold for every n≥0, as item 3 asserts.
Item 4, Hermite's equation. Use 2nHn−1=Hn′ to eliminate Hn−1 from the three-term recurrence:
Hn+1=2xHn−Hn′.
Differentiate this in x:
Hn+1′=2Hn+2xHn′−Hn′′.
On the left, the derivative recurrence at index n+1 says Hn+1′=2(n+1)Hn. Equating and cancelling 2Hn from both sides leaves
2nHn=2xHn′−Hn′′,
which is Hermite's equation. Each hypothesis used here is one of the two recurrences, so this holds for every n≥0.
Item 5, orthogonality and the norm. Take two copies of the generating function, in variables s and t, multiply by the weight, and integrate:
the last integral being finite because the Gaussian beats the exponential; Tonelli's theorem then licenses the interchange (Fubini and Tonelli theorems). Therefore
Both sides are power series in the two variables s and t, convergent on a neighbourhood of the origin, so their coefficients match one by one. The right-hand side has no smtn term with m=n, which gives orthogonality; and matching the sntn terms gives
By item 4, Hn′′−2xHn′=−2nHn, so ψn′′=cne−x2/2(x2−1−2n)Hn=(x2−2n−1)ψn, which rearranges to −ψn′′+x2ψn=(2n+1)ψn. □
Remark(Where the hypotheses are used)Original
n must be a non-negative integer. Riley, Hobson and Bence write Hermite's equation with a real parameter ν and substitute a power series y=∑mamxm, which forces
am+2=−(m+1)(m+2)2(ν−m)am.
If ν=n is a non-negative integer the chain of even or odd coefficients starting below n terminates at m=n, and that branch is the polynomial Hn. If ν is not a non-negative integer, no branch terminates, and for large m the ratio behaves like am+2/am∼2/m — which is exactly the ratio of consecutive coefficients of ex2=∑jx2j/j!, whose xm coefficient 1/(m/2)! satisfies cm+2/cm=1/(m/2+1)∼2/m. So the solution grows like ex2 and the Hermite function e−x2/2y like ex2/2: it is not in L2(R), and item 6 produces no eigenfunction. This is the whole content of energy quantisation in the application below: integrality of n 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 (−∞,∞); the half-line version ∫0∞e−x2HmHn is not δmn times anything, as ∫0∞e−x2H0H1dx=2∫0∞xe−x2dx=1=0 shows at once. In Sturm-Liouville language — Riley, Hobson and Bence note that Hermite's equation has the form with p=ρ=e−x2, q=0, λ=2n — the operator is self-adjoint on this interval precisely because p decays fast enough to kill the boundary terms at ±∞ for all polynomial data (Adjoint and self-adjoint differential operators). Cut the interval at 0 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 (∗) 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 span L2(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) needs item 7.
Remark(Physicists' Hn and probabilists' Hen)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 Hn. Kreyszig flags this explicitly, calling the terminology in the literature not unique, and defines the competing family by
He0(x)=1,Hen(x)=(−1)nex2/2dxndne−x2/2,n=1,2,…
These are the probabilists' polynomials: the weight is the standard normal density up to a constant, e−x2/2, rather than e−x2. Everything in this unit is stated for the physicists' polynomials Hn, 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,
The first few are He0=1, He1=x, He2=x2−1, He3=x3−3x, He4=x4−6x2+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 H1 or at the coefficient of the highest power: physicists' H1=2x and leading coefficient 2n; probabilists' He1=x and leading coefficient 1. Or look at the weight in the orthogonality relation: e−x2 against e−x2/2. 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.
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 Hn 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 n. Start from H0(x)=1, H1(x)=2x and iterate
Hk+1(x)=2xHk(x)−2kHk−1(x),k=1,…,n−1.
For ∣x∣ of order n or larger the values overflow; carry ψk instead, using the same recurrence rewritten for the normalised functions,
Compute an=(2nn!π)−1∫−∞∞f(x)Hn(x)e−x2dx, so that f=∑nanHn in mean square (item 7).
If f is even, all odd an vanish, and conversely: use the parity relation to halve the work.
If f is itself a polynomial of degree N, the expansion terminates at n=N; the fastest route is then not integration but repeated use of
xHn(x)=21Hn+1(x)+nHn−1(x),
which is the three-term recurrence solved for xHn. Starting from H0=1 and multiplying by x one factor at a time expands xN exactly, with rational coefficients and no integrals at all.
To compute moments and matrix elements. Any integral of the form ∫e−x2HmxkHndx is evaluated by applying xHn=21Hn+1+nHn−1 exactly k times to turn xkHn into a finite combination ∑jcjHj, and then reading off the answer as cm⋅2mm!π by item 5. Only one coefficient of the combination is ever needed, and the index shifts by at most one per step, so ∫e−x2HmxkHndx=0 whenever ∣m−n∣>k or m−n+k is odd.
Example(The spread of the third oscillator state)Original
Take the dimensionless quantum oscillator state ψ2, that is n=2, and compute ⟨x2⟩=∫x2ψ22 and ⟨x4⟩=∫x4ψ22 using nothing but the recurrence and item 5. Write N2=22⋅2!⋅π=8π for the norm constant of H2, so that ψ22=e−x2H22/N2.
Step 1: push the powers of x through. Apply xHn=21Hn+1+nHn−1 repeatedly, starting from H2:
The odd moments ⟨x⟩ and ⟨x3⟩ vanish, because the expansions of xH2 and x3H2 contain only odd-index Hermite polynomials and H2 is orthogonal to all of them.
Step 3: check. Direct symbolic integration of x2e−x2H2(x)2/(8π) and of x4e−x2H2(x)2/(8π) over R returns 5/2 and 39/4. The general pattern, obtained the same way for n=0,1,2,3, is
⟨x2⟩n=n+21,⟨x4⟩n=43(2n2+2n+1),
giving (21,43), (23,415), (25,439), (27,475) for n=0,1,2,3; all eight numbers were verified symbolically.
What the numbers mean.⟨x2⟩n=n+21 is the virial theorem in this problem: in the units of item 6 the energy is 21(−d2/dx2+x2) with eigenvalue n+21, and the mean potential energy 21⟨x2⟩=21(n+21) is exactly half of it, so kinetic and potential energies share the total equally. And the kurtosis of the position distribution,
⟨x2⟩2⟨x4⟩=25/439/4=2539=1.56,
is below the Gaussian value 3: the n=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=0 the same ratio is (3/4)/(1/4)=3, as it must be, since ψ02is 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 −2mℏ2ψ′′+2kz2ψ=Eψ and rescale z=bx with
b=(km)1/4ℏ1/2=mωℏ,ω=k/m,
which reduces the equation to −21φ′′+21x2φ=λφ — item 6 with λ=n+21. Undoing the substitution,
The argument of Hn is z/b and not bz: the quantity b is a length, so z/b is the only dimensionless combination available. (Arfken, Weber and Harris print this eigenfunction in eq. (18.19) as Nne−(αz)2/2Hn(αz) with α=ℏ1/2/(km)1/4, which is our b rather than 1/b; the argument needs the reciprocal of the α they print, and their own normalisation constant Nn=[α/(2nπn!)]1/2 of eq. (18.20) agrees with the reciprocal reading, since it is (b2nn!π)−1/2 exactly when α=1/b. We follow the substitution z=ℏ1/2x/(km)1/4 that they make on the preceding page, which fixes the reading unambiguously. That the ψn displayed above solve the dimensional equation with the energies quoted, and are normalised, was checked symbolically for n≤3.)
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 n a photon or phonon number; and the lowest energy is 21ℏω, 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†, 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=21(−d2/dx2+x2)=N+21 with N=A†A, the commutator is [H,A†]=[N,A†]=A†: applying A† moves one rung up, of height 1 in λ. Restoring units multiplies the ladder spacing by ℏω, so for the dimensional Hamiltonian the same relation reads [H,A†]=ℏωA†. The ladder has a bottom because ∥Aψ∥2≥0.
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 for infrared absorption is the statement that ∫e−x2HmxHndx vanishes unless ∣m−n∣=1, which is the k=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) of Hermite functions, labelled TEMmn. The m and n 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.
exact for every polynomial f of degree at most 2n−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 Hen of the convention remark, the Gram-Charlier and Edgeworth expansions write a near-Gaussian probability density as ϕ(x)∑ncnHen(x) with ϕ the standard normal density, the coefficients being combinations of the cumulants: c3 carries the skewness, c4 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! with no stray powers of two.
The polynomials orthogonal on the half-line with weight xke−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.
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≥0. The associated Laguerre polynomialsL0k,L1k,L2k,… are defined by the Rodrigues formula
Lnk(x)=n!exx−kdxndn(xn+ke−x),x>0.
The case k=0 gives the (ordinary) Laguerre polynomials, written Ln=Ln0:
Ln(x)=n!exdxndn(xne−x).
This is the convention of all four of our sources, and it is the one in which Ln(0)=1. In the language of General properties of orthogonal polynomials, Lnk is the classical family belonging to the weight
wk(x)=xke−xon(0,∞),
with the scale fixed by the choice Hassani writes as Kn=n!. Hassani runs the whole construction with a real parameter ν>−1 in place of the integer k, writing Lnν. The only item below that genuinely needs k to be an integer is the equivalent definition in item 2, which differentiates k times; everything else survives with each factorial in k read through the gamma function (The Gamma function), (n+k)!↦Γ(n+ν+1) and (k+m)!↦Γ(ν+m+1), so that the explicit sum of item 1 becomes a series in Γ's and its value at the origin becomes Lnν(0)=Γ(n+ν+1)/[n!Γ(ν+1)] (Hassani, Problem 8.32), just as item 6 already writes the norm as Γ(n+ν+1)/n!. We note below where the real case differs.
Note the sign: Ln starts at +1 and its leading term is (−1)nxn/n!, so these polynomials are not monic and alternate in overall sign at large x.
The Laguerre functionsφnk(x)=e−x/2xk/2Lnk(x) absorb half the weight, so that they are orthogonal with unit weight; Kreyszig's orthonormal sequence in L2[0,∞) is the case k=0, namely e−x/2Ln(x), which he obtains by Gram-Schmidt from xne−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≥0 be an integer and Lnk as defined above. Write L−1k≡0.
Explicit form. For n≥0,
Lnk(x)=m=0∑n(−1)m(n−m)!(k+m)!m!(n+k)!xm.
So Lnk is a polynomial of degree exactly n with leading coefficient (−1)n/n!, and
Lnk(0)=(nn+k)=n!k!(n+k)!,in particular Ln(0)=1.
Unlike the Legendre, Hermite and Chebyshev families, the Lnk have no parity: the interval (0,∞) is not symmetric.
Relation to the ordinary polynomials.Lnk(x)=(−1)kdxkdkLn+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 k to be a non-negative integer, and is the one item of this statement that does not survive to Hassani's real ν.
Generating function. For ∣h∣<1 and all x,
(1−h)k+1e−xh/(1−h)=n=0∑∞Lnk(x)hn.
Recurrences. For n≥1 and k≥0,
(n+1)Ln+1k(x)=(2n+k+1−x)Lnk(x)−(n+k)Ln−1k(x),x(Lnk)′(x)=nLnk(x)−(n+k)Ln−1k(x),(Lnk)′(x)=−Ln−1k+1(x),
and, for every n≥0,
Lnk+1=j=0∑nLjk.
The last two relations move between neighbouring values of the upper index; note that differentiating raisesk and lowers n. (With the convention L−1k≡0 the first three relations hold at n=0 as well; the fourth needs no convention.)
The associated Laguerre equation.y=Lnk solves
xy′′+(k+1−x)y′+ny=0,
which for k=0 is Laguerre's equation xy′′+(1−x)y′+ny=0. The point x=0 is a regular singular point and x=∞ an irregular one.
Orthogonality and norm. For all m,n≥0,
∫0∞e−xxkLmk(x)Lnk(x)dx=n!(n+k)!δmn.
For k=0 the right-hand side is simply δmn: the ordinary Laguerre polynomials are already orthonormal for the weight e−x, with no normalising constant at all. In Hassani's real-parameter version the constant is Γ(n+ν+1)/n!, which reduces to (n+k)!/n! at ν=k (The Gamma function).
Self-adjoint form and completeness.φnk=e−x/2xk/2Lnk satisfies
x(φnk)′′+(φnk)′+(−4x+22n+k+1−4xk2)φnk=0,
and the closely related Φnk=e−x/2x(k+1)/2Lnk satisfies the equation with no first-derivative term,
(Φnk)′′+(−41+2x2n+k+1−4x2k2−1)Φnk=0,
which is the form in which the hydrogen radial equation appears. For k=0, Kreyszig states that the normalised functions e−x/2Ln(x) form a total orthonormal sequence in L2[0,∞); 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−x exactly n times produces a sum over how many of the derivatives landed on the power and how many on the exponential; dividing out the xke−x that was never differentiated leaves a polynomial of degree n, 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−x satisfies xv′=(n+k−x)v — one line, because v is a power times an exponential. Differentiating that identity n+1 times with Leibniz's rule relates v(n), v(n+1) and v(n+2), and since v(n)isn!xke−xLnk 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 n times. The boundary terms vanish because what is being differentiated is a power at least xk at the origin and an exponential at infinity; and after n parts the other factor has been differentiated n times, which annihilates it if its degree is smaller. The one surviving case is m=n, where the nth derivative of a degree-n polynomial is a constant and the integral collapses to Γ(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), letting r derivatives fall on the power and n−r on the exponential:
Multiply by exx−k/n! and substitute m=n−r, so that xn+k−r⋅x−k=xm, (rn)=(n−m)!m!n! after the substitution, and (n+k−r)!=(k+m)!:
Lnk(x)=m=0∑n(−1)m(n−m)!(k+m)!m!(n+k)!xm.
The m=n term is (−1)n(n+k)!xn/(0!(k+n)!n!)=(−1)nxn/n!, which is the leading term and is non-zero, so the degree is exactly n. The m=0 term is (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:
Insert G=∑nLnkhn and collect the coefficient of hn on each side. On the left, (1−2h+h2)∑nnLnkhn−1 gives (n+1)Ln+1k−2nLnk+(n−1)Ln−1k; on the right, (k+1−x)Lnk−(k+1)Ln−1k. Equating and moving all but the first term to the right,
(n+1)Ln+1k=(2n+k+1−x)Lnk−(n+k)Ln−1k,
where the two coefficients were assembled as 2nLnk+(k+1−x)Lnk=(2n+k+1−x)Lnk and −(n−1)Ln−1k−(k+1)Ln−1k=−(n+k)Ln−1k.
Differentiating in x instead gives ∂G/∂x=−1−hhG, that is (1−h)∂xG=−hG, whose hn coefficient is
(Lnk)′−(Ln−1k)′=−Ln−1k.
Summing this from index 1 to n telescopes, and (L0k)′=0, so
(Lnk)′=−i=0∑n−1Lik.
Finally, comparing the generating functions for two neighbouring upper indices, Gk+1=Gk/(1−h), and expanding 1/(1−h)=∑jhj, the hn coefficient of the product is Lnk+1=∑j=0nLjk. Combining the last two displays gives (Lnk)′=−Ln−1k+1.
For the remaining recurrence, differentiate the series of item 1 term by term and multiply by x:
and (n−m)!n−(n−1−m)!1=(n−m)!n−(n−m)=(n−m)!m, which is the coefficient on the left.
Item 2. Iterating (Lnk)′=−Ln−1k+1 a total of k times starting from Ln+k0=Ln+k gives
dxkdkLn+k=(−1)kLnk,
which is the stated relation.
Item 5. Put v(x)=xn+ke−x, so that v(n)=n!xke−xLnk by the Rodrigues formula. Because v is a power times an exponential it satisfies the first-order identity
xv′=(n+k−x)v.
Differentiate this n+1 times with Leibniz's rule. On the left, only the first two Leibniz terms survive because x is linear:
(xv′)(n+1)=xv(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).
Equating and collecting,
xv(n+2)+(1−k+x)v(n+1)+(n+1)v(n)=0.(†)
Now write y=Lnk and express the three derivatives in terms of y. From v(n)=n!xke−xy,
It vanishes at ∞ because every derivative of xn+ke−x is a polynomial times e−x; and at 0 because differentiating xn+ke−x fewer than n+k times leaves a factor xn+k−(n−1−j)=xk+1+j, which is a positive power of x since k≥0 and j≥0. Hence
I=n!(−1)n∫0∞(Lmk)(n)(x)xn+ke−xdx.
If m<n the nth derivative of a polynomial of degree m is identically zero, so I=0; by symmetry of I in m and n this covers every m=n. If m=n, the nth derivative of Lnk is n! times its leading coefficient, namely n!⋅(−1)n/n!=(−1)n by item 1, a constant. Therefore
using the integral definition of the gamma function (The Gamma function). □
Remark(Where the hypotheses are used)Original
n 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 and run Frobenius about the regular singular point x=0 (The Frobenius method). The indicial equation is σ2=0, a repeated root, so there is only one Frobenius solution; it is an ordinary power series with
am+1=(m+1)2m−νam,
which terminates exactly when ν is a non-negative integer. If it does not terminate, am+1/am→1/m for large m, which is the coefficient ratio of ex, and the solution grows like ex. Arfken, Weber and Harris record the same fact for the associated equation — its non-polynomial solutions diverge proportionally to xkex — and then use precisely this to quantise the hydrogen spectrum: the radial function contains the factor e−ρ/2Lλ−L−12L+1(ρ), and only when λ−L−1 is a non-negative integer does e−ρ/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 logx and blows up at the origin. It is excluded by requiring regularity at x=0. The two exclusions are independent: integrality handles x=∞, regularity handles x=0, and a Sturm-Liouville problem on (0,∞) needs a condition at each end.
k>−1, not merely k≥0. Nothing in item 6 needs k to be an integer; what it needs is that xk be integrable near the origin and that the boundary terms vanish there. Both fail at k≤−1: already ∫01x−1e−xdx=∞, so the inner product itself is undefined. This is exactly Hassani's standing restriction ν>−1 in his Proposition 8.2.1(b), and it is why the natural parameter range is k>−1 rather than k≥0. Integrality of k is needed only for item 2, which expresses Lnk as a k-fold derivative. The factorials (n+k)! and (k+m)! of item 1, and the binomial coefficient (nn+k) with them, are not exceptions: for real ν>−1 they are read as Γ(n+ν+1) and Γ(ν+m+1), and the value at the origin becomes Γ(n+ν+1)/[n!Γ(ν+1)], which is how Hassani states it; item 6 is already written that way.
A different weight for every k. The orthogonality in item 6 holds within a fixed upper index. There is no sense in which L32 and L35 are orthogonal: they live in different weighted spaces. Arfken, Weber and Harris make the point explicitly — for each k the associated Laguerre polynomials form a new orthogonal set. In the hydrogen application this is the reason that radial functions with different L are not orthogonal through the radial integral alone; their orthogonality comes from the angular factor (Spherical harmonics).
Where k≥0 enters the boundary terms. In the proof of item 6 the surviving power at the origin was xk+1+j with j≥0. For −1<k<0 this is still a positive power, so the argument survives; for k≤−1 the j=0 boundary term need not vanish, and the integration by parts breaks down at the first step.
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 for the other family, those authors set
Lnk(x)=dxkdkLn(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)kLn+kk(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 and the subscript, by k. Lnk has degree n−k, while Lnk has degree n.
This unit uses Lnk 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)=(nn+k)>0 and the polynomial has degree n; in the other, Lnk has degree n−k and Lnk(0)=(−1)k(n−kn) alternates in sign with k. A hydrogen radial function written with the wrong one goes wrong twice over. It acquires a sign (−1)2L+1=−1, which a normalisation constant hides; and — far worse — the subscript is wrong by 2L+1. In our convention the correct factor for the state with principal number N and orbital number L is LN−L−12L+1, of degree N−L−1, and the other convention calls that same object −LN+L2L+1; so a formula carried across without translating the subscript becomes LN+L2L+1, of degree N+L. Since Lmk has exactly m simple zeros in (0,∞), the symptom is a radial function with N+L nodes instead of N−L−1 — for 2p (N=2, L=1) three nodes where the true function has none, and for 3d (N=3, L=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! is what puts the 1/n! in front of the Rodrigues formula. Dropping it (as some older texts do) multiplies Lnk by n! and changes the norm in item 6 from (n+k)!/n! to 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/x and a 1/x2 term and no obvious polynomial in sight. The recipe is to strip off the asymptotics at both ends and read off what is left.
Scale so that the exponential decay rate is 1/2. If the equation is u′′+(−4α2+rβ−r2γ)u=0, put x=αr. The equation becomes u′′+(−41+xβ/α−x2γ)u=0 in the new variable.
Match to the normal form of item 7. Compare with
Φ′′+(−41+2x2n+k+1−4x2k2−1)Φ=0.
Reading the 1/x2 coefficient gives k=4γ+1 — take the positive root, since the other sign makes x(k+1)/2 singular at the origin. Reading the 1/x coefficient gives n=21(2β/α−k−1).
Impose integrality. The solution is bounded at infinity only if n is a non-negative integer. This is the quantisation condition, and solving it for whatever physical parameter sits inside α or β gives the spectrum.
Write down the solution.Φ=e−x/2x(k+1)/2Lnk(x), and the original unknown is recovered by undoing step 1.
Normalise. Use item 6: ∫0∞e−xxk(Lnk)2dx=(n+k)!/n!, and multiply by whatever Jacobian the change of variable in step 1 introduced.
To expand a function on (0,∞): by item 6 the coefficients in f=∑nanLnk are
an=(n+k)!n!∫0∞f(x)Lnk(x)xke−xdx.
For k=0 the prefactor is 1. Evaluate Lnk at a point by the three-term recurrence of item 4, starting from L0k=1 and L1k=k+1−x, never by the explicit sum of item 1, whose terms alternate and grow before cancelling.
Example(The hydrogen 2p radial function)Original
Take hydrogen (Z=1) in the state with principal quantum number N=2 and orbital quantum number L=1, and build the radial function from scratch by the recipe above, then compute ⟨r⟩ and ⟨1/r⟩.
Step 1: reduce the radial equation. With u(r)=rR(r) the radial Schrodinger equation for a Coulomb potential is
u′′+(−4α2+a0r2−r2L(L+1))u=0,α2=−ℏ28mE,
with a0 the Bohr radius. Set x=αr. Comparing with the normal form, the 1/x2 coefficient gives k2−1=4L(L+1), that is k=2L+1=3; the 1/x coefficient gives 2n+k+1=4/(αa0), which with k=2L+1 reads 2(n+L+1)=4/(αa0), that is α=2/(Na0) with N=n+L+1 the principal quantum number. For N=2, L=1 we get n=N−L−1=0 and α=1/a0.
Step 2: write the function. With n=0 the Laguerre factor is L03=1, so
Step 5: check. Arfken, Weber and Harris quote, for the normalised hydrogenic radial function, ⟨r⟩=2a0[3N2−L(L+1)] and ⟨r−1⟩=1/(N2a0). At N=2, L=1 these give 2a0(12−2)=5a0 and 1/(4a0): both match. The normalisation, ⟨r⟩ and ⟨1/r⟩ were also recomputed symbolically from the general formula
and agree; the same check at N=3, L=0 returns ⟨r⟩=227a0, again matching the quoted formula.
Reading the numbers.⟨1/r⟩=1/(N2a0) depends only on N, not on L — which is why the Coulomb energy −21e2⟨1/r⟩/(4πϵ0) is degenerate in L, the accidental degeneracy of the hydrogen atom. By contrast ⟨r⟩=5a0 does depend on L: the 2s state (N=2, L=0) has ⟨r⟩=6a0, further out on average than 2p even though 2s has more weight very close to the nucleus. The two facts together are the whole story of why ⟨r⟩ is a poor predictor of chemical behaviour and ⟨1/r⟩ is a good one.
ApplicationOriginal
The hydrogen atom. Separating the Schrodinger equation for a charge Ze 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+1, as in the example. The bound states are labelled by N=n+L+1=1,2,3,…, the energies are
EN=−2N2ℏ2Z2m(4πϵ0e2)2,
and the radial function has n=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 N alone, because the quantisation condition in step 3 of the method involves only the combination 2n+k+1=2N; the degeneracy of level N is ∑L=0N−1(2L+1)=N2; and the node count is N−L−1, so the 1s, 2p, 3d, … 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.
exact for all polynomials f of degree at most 2n−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 should be rescaled to 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)∝(wr2)∣ℓ∣Ln∣ℓ∣(w22r2)e−r2/w2eiℓϕ.
The index n counts radial rings, the index ℓ is the azimuthal winding, and each photon in such a beam carries orbital angular momentum ℓℏ. 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∣ for different n at fixed ℓ, together with the orthogonality of eiℓϕ.
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 x, called Sonine polynomials in that literature and written Sν(r), with the two indices in the opposite positions to ours: Sν(r) is a multiple of our Lrν, of degree r in x. The two tensor ranks of the perturbation use different ν. The vector part, driven by the temperature gradient, is expanded in the S3/2(r), orthogonal for the weight x3/2e−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), orthogonal for the weight x5/2e−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. 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) is a complete orthonormal sequence in L2[0,∞) (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, so a scaling parameter must be fitted to the problem.
Disguised cosines: Tn(cosθ)=cosnθ gives polynomials with a three-term recurrence, orthogonality for the weight (1−x2)−1/2, and the minimax property that makes 21−nTn the monic polynomial of degree n that stays closest to zero on [−1,1].
turns it, because d/dx=(−1/sinθ)d/dθ, into d2y/dθ2+n2y=0, whose solutions are cosnθ and sinnθ. The Chebyshev polynomials of the first kind are the first of these written back in x:
Tn(x)=cos(narccosx),−1≤x≤1,n=0,1,2,…
The second solution, Vn(x)=sin(narccosx), is not a polynomial. Riley, Hobson and Bence and Arfken, Weber and Harris both define two further functions by dividing out the square root,
of which Un is a polynomial of degree n — the Chebyshev polynomial of the second kind — and Wn 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: Tn is the case μ=ν=−21, with weight (1−x2)−1/2 on [−1,1], and Un the case μ=ν=+21, with weight (1−x2)+1/2. Arfken, Weber and Harris instead obtain them as the ultraspherical polynomials Cn(α) at α=0 and α=1, the first of those needing a limit because the generating function degenerates at α=0; the Legendre polynomials (Legendre polynomials) sit between them at α=21.
Outside [−1,1] the definition is continued by Tn(x)=cosh(narccoshx) for x>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] 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≥0 and x∈[−1,1], with x=cosθ and θ∈[0,π]:
Polynomial character.Tn is a polynomial of degree n; its leading coefficient is 2n−1 for n≥1 and 1 for n=0. Un is a polynomial of degree n with leading coefficient 2n. Both have the parity of n: Tn(−x)=(−1)nTn(x) and Un(−x)=(−1)nUn(x).
Recurrences. With T0=1, T1=x, U0=1, U1=2x,
Tn+1(x)=2xTn(x)−Tn−1(x),Un+1(x)=2xUn(x)−Un−1(x).
The two families obey the same recurrence and differ only in the starting pair.
Special values.Tn(1)=1,Tn(−1)=(−1)n,T2m(0)=(−1)m,T2m+1(0)=0,Un(1)=n+1,Un(−1)=(−1)n(n+1),U2m(0)=(−1)m,U2m+1(0)=0.
Differential equations.Tn solves (1−x2)y′′−xy′+n2y=0, and Un solves (1−x2)y′′−3xy′+n(n+2)y=0.
Zeros and extrema (for n≥1). Tn has n simple zeros, all in (−1,1),
xj=cos2n(2j−1)π,j=1,…,n,
and n+1 points of extreme value in [−1,1],
yj=cosnjπ,j=0,…,n,withTn(yj)=(−1)j.
In particular ∣Tn(x)∣≤1 on [−1,1], with equality exactly at the yj.
Orthogonality.∫−111−x2Tm(x)Tn(x)dx=⎩⎨⎧0,π,π/2,m=n,m=n=0,m=n≥1,∫−11Um(x)Un(x)1−x2dx=2πδmn.
The exceptional value at m=n=0 in the first kind is not a blemish; it is the reason the constant term of a Chebyshev series is conventionally written 21a0.
Generating functions. For ∣h∣<1 and ∣x∣≤1,
1−2xh+h21−xh=n=0∑∞Tn(x)hn,1−2xh+h21=n=0∑∞Un(x)hn.
Arfken, Weber and Harris and Hassani give the first in the equivalent form 1−2xh+h21−h2=T0(x)+2∑n≥1Tn(x)hn, which is twice the display above minus one.
Relations between the two kinds.Tn′(x)=nUn−1(x),Tn(x)=Un(x)−xUn−1(x),(1−x2)Un−1(x)=xTn(x)−Tn+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≥1 and write ∥q∥∞=max−1≤x≤1∣q(x)∣. Among all monic real polynomials q of degree n,
qmonic,degq=nmin∥q∥∞=21−n,
and the minimum is attained at q=21−nTn and at no other monic polynomial of degree n.
Equivalently — and this is the form in which Kreyszig poses it — the best uniform approximation to xn on [−1,1] by a polynomial of degree at most n−1 is
y(x)=xn−21−nTn(x),
with maximum error 21−n; and if p(x)=βnxn+… is any real polynomial of degree exactly n, its best uniform approximation by a polynomial of degree at most n−1 is p(x)−βn21−nTn(x), with maximum error ∣βn∣21−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] 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−1 form an n-dimensional subspace satisfying the Haar condition, since a non-zero one of them has at most n−1 zeros. Stakgold and Holst state the same minimality and attribute it to the equi-oscillation of Tn, 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 cosnθis a polynomial in cosθ is the one thing needing an argument, and it is the addition formula: cos(n+1)θ+cos(n−1)θ=2cosθcosnθ. Read as a statement about polynomials this is the three-term recurrence, and running it from 1 and x both proves the polynomial character and doubles the leading coefficient at each step, which is where 2n−1 comes from. Orthogonality is then the orthogonality of cosines on [0,π]: the weight (1−x2)−1/2 is nothing but the Jacobian dx=−sinθdθ in disguise, and a Chebyshev series is literally a Fourier cosine series of f(cosθ).
The minimax property is the one genuinely different idea. Tn oscillates between +1 and −1, touching both n+1 times, and that is a lot of alternations for a polynomial of degree n. Suppose some monic q of degree n stayed strictly closer to zero than 21−nTn does. Then the difference 21−nTn−q would inherit the sign of 21−nTn at each of the n+1 alternation points, so it would change sign n times and have n roots. But the leading terms cancel, so the difference has degree at most n−1 and cannot have n roots unless it vanishes identically. The hypothesis is contradicted, and the only escape is that q was 21−nTn 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=cosθ with θ∈[0,π], so that θ=arccosx is a decreasing bijection of [−1,1] onto [0,π] and sinθ=1−x2≥0.
Item 2, the recurrences. Adding the two cosine addition formulas cos(n±1)θ=cosnθcosθ∓sinnθsinθ cancels the sine terms:
cos(n+1)θ+cos(n−1)θ=2cosθcosnθ.
Written in x, this is Tn+1+Tn−1=2xTn. Adding the two sine addition formulas sin(n±1)θ=sinnθcosθ±cosnθsinθ likewise gives sin(n+1)θ+sin(n−1)θ=2cosθsinnθ; dividing by sinθ (non-zero for 0<θ<π) and shifting the index by one gives Un+1+Un−1=2xUn, first on (−1,1) and then at x=±1 by continuity, both sides being continuous.
Item 1, from item 2.T0=1 and T1=x are polynomials of degrees 0 and 1 with leading coefficients 1 and 1=20. If Tn−1 and Tn are polynomials of degrees n−1 and n with Tn having leading coefficient 2n−1, then Tn+1=2xTn−Tn−1 has degree n+1 with leading coefficient 2⋅2n−1=2n, the subtracted term being of lower degree. Induction gives the claim for all n≥1. (Kreyszig proves the same statement in the equivalent trigonometric form, as the existence of an expansion cosnθ=2n−1cosnθ+∑j<nβnjcosjθ.) The argument for Un is identical, starting from U0=1 and U1=2x and giving leading coefficient 2n.
For the parity, note that x↦−x corresponds to θ↦π−θ, under which cosnθ↦cos(nπ−nθ)=(−1)ncosnθ; and sin((n+1)(π−θ))/sin(π−θ)=(−1)nsin((n+1)θ)/sinθ by the same computation. Alternatively the recurrence shows by induction that Tn contains only powers of the parity of n.
Item 3. Put θ=0, θ=π and θ=π/2 in Tn=cosnθ: these are x=1, x=−1, x=0, giving 1, cosnπ=(−1)n and cos(nπ/2), which is (−1)m for n=2m and 0 for n odd. For Un=sin((n+1)θ)/sinθ the values at θ=0,π need a limit: as θ→0, sin((n+1)θ)/sinθ→n+1 by L'Hopital's rule or by comparing the leading terms of the two sines, and as θ→π one gets (−1)n(n+1) the same way. At θ=π/2, sin((n+1)π/2)/1 is (−1)m for n=2m and 0 for n odd.
Item 4. Differentiating Tn=cosnθ with d/dx=(−1/sinθ)d/dθ reverses the substitution made in the definition: y=cosnθ satisfies yθθ+n2y=0, and the chain rule converts that back to (1−x2)y′′−xy′+n2y=0. For Un, note that Vn+1=1−x2Un solves Chebyshev's equation with n replaced by n+1, since Vn+1=sin((n+1)θ) is the other solution of yθθ+(n+1)2y=0. Substituting Vn+1=(1−x2)1/2Un and its derivatives
into (1−x2)Vn+1′′−xVn+1′+(n+1)2Vn+1=0 and dividing by (1−x2)1/2 collects to
(1−x2)Un′′−3xUn′−Un+(n+1)2Un=0,
and (n+1)2−1=n(n+2).
Item 5.Tn(x)=0 means cosnθ=0, that is nθ=(2j−1)π/2 for an integer j; the values of θ in (0,π) are θj=(2j−1)π/(2n) for j=1,…,n, and these give n distinct values xj=cosθj in (−1,1) because cos is injective on [0,π]. A polynomial of degree n has at most n roots, so these are all of them and each is simple. Similarly ∣Tn∣=∣cosnθ∣ attains the value 1 exactly when nθ is a multiple of π, that is at θ=jπ/n for j=0,…,n, where Tn=cosjπ=(−1)j; and ∣cosnθ∣≤1 everywhere.
Item 6. Substitute x=cosθ, so dx=−sinθdθ and 1−x2=sinθ, which cancels:
∫−111−x2TmTndx=∫0πcosmθcosnθdθ.
This is the standard cosine orthogonality on [0,π] (Fourier series), equal to 0 for m=n, to π/2 for m=n≥1, and to π for m=n=0 where the integrand is 1. For the second kind, UmUn1−x2dx=−sin((m+1)θ)sin((n+1)θ)dθ after the same substitution, and ∫0πsin((m+1)θ)sin((n+1)θ)dθ=2πδmn, with no exceptional case because m+1≥1 always.
Item 7. Sum the geometric-like series directly. For ∣h∣<1 and x=cosθ, ∑n≥0cos(nθ)hn=Re∑n≥0(heiθ)n=Re1−heiθ1, and multiplying numerator and denominator by the conjugate 1−he−iθ gives
1−2hcosθ+h21−hcosθ=1−2xh+h21−xh.
Taking the imaginary part instead gives ∑n≥1sin(nθ)hn=hsinθ/(1−2xh+h2); dividing by sinθ and shifting the index gives the generating function for Un.
Item 8.Tn′(x)=(−1/sinθ)d(cosnθ)/dθ=nsinnθ/sinθ=nUn−1(x). The second relation is the sine addition formula sin((n+1)θ)−cosθsin(nθ)=cos(nθ)sinθ divided by sinθ. The third comes from the cosine addition formula in the form cosθcos(nθ)−cos((n+1)θ)=sinθsin(nθ): the left side is xTn−Tn+1, and the right side is sinθ⋅sinθUn−1=(1−x2)Un−1. □
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 Tn=21−nTn. By item 1 of the first statement, Tn is monic of degree n, and by item 5 it satisfies ∥Tn∥∞=21−n, attained with alternating signs at the n+1 points yj=cos(jπ/n):
Tn(yj)=(−1)j21−n,j=0,1,…,n.
Note 1=y0>y1>⋯>yn=−1, since cos is decreasing on [0,π].
The minimum is not smaller. Let q be monic of degree n and suppose, for contradiction, that ∥q∥∞<21−n. Put
d=Tn−q.
Both are monic of degree n, so the xn terms cancel and degd≤n−1. At each yj,
(−1)jd(yj)=21−n−(−1)jq(yj)≥21−n−∥q∥∞>0,
so d(yj) is non-zero and has the sign of (−1)j. Since d is continuous and changes sign on each of the n intervals (yj+1,yj), j=0,…,n−1, the intermediate value theorem gives a root of d strictly inside each, hence n distinct roots. A polynomial of degree at most n−1 with n distinct roots is identically zero, so d≡0 and q=Tn; but then ∥q∥∞=21−n, contradicting the assumed strict inequality. Therefore ∥q∥∞≥21−n for every monic q of degree n, and Tn attains the bound.
Uniqueness. Minimising ∥q∥∞ over monic q of degree n is the same problem as minimising ∥xn−y∥∞ over y in the subspace Y of polynomials of degree at most n−1, by the correspondence q=xn−y. Now dimY=n, and a non-zero element of Y has at most n−1 zeros in [−1,1], so Y satisfies the Haar condition (Kreyszig's Definition 6.3-2). Kreyszig's theory then gives a unique best uniform approximation to xn out of Y — 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 q, which by the previous paragraph is Tn.
The two restatements. The first is just the correspondence above read backwards: the best approximation is y=xn−Tn, whose degree really is at most n−1 because the leading terms cancel, and the error is ∥Tn∥∞=21−n. For the second, write p=βnxn+(lower) with βn=0. Then p−y ranges, as y runs over Y, over all polynomials of degree exactly n with leading coefficient βn, that is over βnq with q monic; and ∥βnq∥∞=∣βn∣∥q∥∞ is minimised exactly when ∥q∥∞ is. □
Remark(Where the hypotheses are used)Original
The interval must be [−1,1], and rescaling costs a factor. Everything above is tied to [−1,1] through x=cosθ. On a general interval [a,b], substitute
x=b−a2t−a−b,
which maps [a,b] onto [−1,1]. The monic polynomial of degree n in t of smallest sup-norm on [a,b] is then (2b−a)nTn(x), and its sup-norm is
2(4b−a)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∥∞ over polynomials of degree n is 0, attained by q≡0. 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)=1.
The norm must be the maximum norm. In the L2 norm with weight (1−x2)−1/2 the minimising monic polynomial is also Tn, because orthogonal polynomials always minimise the weighted L2 norm among monic polynomials of their degree (General properties of orthogonal polynomials); but in the unweighted L2 norm on [−1,1] it is the monic Legendre polynomial instead (Legendre polynomials), and in the L1 norm it is 2−nUn. Different norms, different extremal polynomials. What is special about Tn 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 Tn is monic and that it attains ±∥Tn∥∞ alternately at n+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 n alternation points, because then d would be forced to have only n−1 roots, which a polynomial of degree n−1 can have. The margin is exactly one: Tn has n+1 extreme points on a closed interval, including both endpoints, and dropping either endpoint from [−1,1] destroys the proof. This is also why the minimax polynomial for a general function equi-oscillates at n+2 points rather than n+1 when the approximating space has dimension n+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(cosθ)=sinθsin((n+1)θ),
a polynomial of degree n with Un(1)=n+1. Riley, Hobson and Bence and Arfken, Weber and Harris give the non-polynomial sin(nθ) the separate name Vn.
Kreyszig, in footnote 1 on p. 348, attaches the label second kind to sinnθ instead, writing it Un(t) as well. That is the function the other three call Vn, and it is not a polynomial.
This unit follows the first group, so Un here always means sin((n+1)θ)/sinθ, the degree-n polynomial tabulated in item 1, orthogonal for the weight (1−x2)+1/2. Nothing in this unit uses Kreyszig's Un; his contribution here is the minimax theorem, which concerns Tn only and is unaffected. The quick test on a formula from an unfamiliar source: evaluate at x=1. Our Un(1)=n+1; Kreyszig's vanishes identically there, since sinnθ→0 as θ→0.
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 T comes from one of those transliterations, as Kreyszig remarks. Second, the normalisation of the orthogonality integral: all our sources use the Tn with Tn(1)=1 and therefore all get π at n=0 and π/2 for n≥1, but an orthonormal convention rescaling Tn by 2/π 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 f continuous on [−1,1]:
Write the Chebyshev series f(x)=21a0+∑n≥1anTn(x) with
an=π2∫−111−x2f(x)Tn(x)dx=π2∫0πf(cosθ)cos(nθ)dθ.
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(cosθ) (Fourier series). Everything known about the convergence of Fourier series transfers, with one bonus — f(cosθ) is automatically even and 2π-periodic, so there is no jump at the ends of the interval and no Gibbs phenomenon there, however badly f behaves at ±1 in the x variable.
In practice evaluate the an by the discrete cosine transform: sample f at the N+1 points yj=cos(jπ/N) and transform. This costs O(NlogN) with a fast algorithm (The discrete Fourier transform and the FFT) and returns the coefficients of the polynomial interpolatingf at those points, which for smooth f differs from the truncated series only in the tail.
Truncate after TN. If the coefficients are decaying fast, the error is dominated by the first omitted term, so it is close to aN+1TN+1(x): of size ∣aN+1∣, nearly equal in magnitude at N+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.
Evaluate the truncated series by the recurrence of item 2, never by expanding into powers of x: the expansion into monomials involves large cancelling coefficients and loses precision.
To economise a polynomial. Given p of degree n with leading coefficient βn, and a tolerance ε:
Compute ∣βn∣21−n. If it exceeds ε, stop: the degree cannot be reduced within tolerance.
Otherwise replace p by p−βn21−nTn, which has degree at most n−1 and differs from p by at most ∣βn∣21−n anywhere on [−1,1]. This is Kreyszig's formula (9).
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 costs ∣βn∣, while economising it costs only ∣βn∣21−n — a saving of a factor 2n−1.
Example(A cubic for ex on [−1,1])Original
Approximate f(x)=ex on [−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=π2∫0πecosθcos(kθ)dθ=2Ik(1),
where Ik is the modified Bessel function of the first kind, whose standard integral representation is Ik(z)=π1∫0πezcosθcos(kθ)dθ (Modified Bessel functions). So the coefficients are known in closed form:
(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)=21a0+a1T1+a2T2+a3T3,
and substituting T2=2x2−1, T3=4x3−3x and collecting powers,
Its largest error on [−1,1], found by sampling on a fine grid, is
[−1,1]max∣ex−p3(x)∣=0.0060656,
attained at x=1; at x=−1 the error is 0.0049733, and the interior extrema are of comparable size with alternating signs. The first omitted coefficient predicted this: ∣a4∣=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+2x2+6x3 has
[−1,1]maxex−(1+x+2x2+6x3)=e−38=0.0516152,
attained at x=1 and tiny near x=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
The Chebyshev truncation is within 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)=x4, with n=4 and β4=1. Step 1 of the economisation recipe gives β421−4=81, and step 2 replaces x4 by
x4−81T4(x)=x4−81(8x4−8x2+1)=x2−81.
So the best uniform cubic (indeed quadratic) approximation to x4 on [−1,1] is x2−81, with error exactly 81=0.125 everywhere the error is extreme — at the five points ±1,±1/2,0, with alternating signs, as item 5 requires. Dropping the x4 term instead would have cost 1: economisation is eight times better here, the predicted factor 2n−1=23.
ApplicationOriginal
Spectral methods on a bounded interval. A Chebyshev spectral method represents the unknown by its values at the extrema yj=cos(jπ/N) and differentiates by transforming to coefficients, differentiating the series with Tn′=nUn−1, and transforming back (Spectral methods). For a smooth solution the error falls faster than any power of N, 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 yj near ±1, with spacing O(N−2) there against 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 yj 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 N rather than 2N−1 (Numerical quadrature). Its practical advantages are that the nodes are nested — doubling N 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=b, the error after k steps is ek=Pk(A)e0 for some polynomial Pk with Pk(0)=1, so the worst-case error reduction is governed by maxλ∈σ(A)∣Pk(λ)∣. Stakgold and Holst set this up as a mini-max problem over the spectral interval [λmin,λmax] and identify the optimum as the Chebyshev polynomial shifted and scaled onto that interval and divided by its value at the origin,
The numerator is bounded by 1 on the spectrum by item 5; the denominator is Tk evaluated outside[−1,1], where the cosh continuation makes it grow geometrically. The resulting bound is
∥ek∥A≤2(κ+1κ−1)k∥e0∥A,
with κ 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 and λmax in advance (Solving linear systems numerically). Stakgold and Holst also record the degenerate case k=1, where T1(x)=x reproduces the optimal relaxation parameter αopt=2/(λmin+λmax) 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)). Inside the passband Tn oscillates between ±1, so the response ripples between 1 and 1/(1+ϵ2) — equal ripple, which is the minimax property transcribed into engineering — and outside, where Tn grows like 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 against n, 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 sinx 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−1 computed above, and it is why the coefficients printed in function-approximation handbooks are never the Taylor coefficients.
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.
In (L) the origin is an ordinary point whose nearest singular neighbours are x=±1, so there are two solutions in powers of x, each with radius of convergence at least 1 — exactly 1 for general ℓ, which is the case Riley, Hobson and Bence record. Substituting y=∑k≥0akxk gives the two-step recurrence
ak+2=(k+1)(k+2)k(k+1)−ℓ(ℓ+1)ak,
and the choices (a0,a1)=(1,0) and (0,1) produce an even solution ye and an odd solution yo respectively. When ℓ=n the numerator vanishes at k=n, so exactly one of the two series terminates — ye if n is even, yo if n 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).
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 n is
Their stated reason for these particular constants is that the resulting Qn then satisfy, verbatim, the recurrences already established for Pn. Part 6 of the statement below identifies αn and βn 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−x2, so a second solution is
Qn(x)=Pn(x)∫x(1−t2)Pn(t)2dt,(R)
and they remark that (R) fixes both the scale and the functional form of Qn. By construction Qn is of the opposite parity to Pn, since it is the series that Pn is not.
Outside the interval. Formulas for Qn on (−1,1) contain ln1−x1+x. For real x>1 this is complex, and Arfken, Weber and Harris replace it by lnx−1x+1, discarding the resulting constant multiple ±iπPn of the first solution. In the complex plane the logarithm forces a branch cut, traditionally [−1,1] (Branch points and branch cuts); the functions on (−1,1) are the average of the two sides of that cut. Hassani arrives at Qλ for arbitrary complex degree by specialising the second-kind Jacobi functions Qλ(α,β) to α=β=0,
with F the hypergeometric function of The hypergeometric function; at λ=n this agrees with the x>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≥0 be an integer. Parts 1–6 concern the interval −1<x<1; part 7 steps outside it, to x∈(1,∞). Then:
Closed form. There is a polynomial Rn of degree exactly n−1 (with R0≡0) such that
Qn(x)=Pn(x)L(x)−Rn(x),Rn(x)=m=1∑nm1Pm−1(x)Pn−m(x).
In particular
Q0=L,Q1=xL−1,Q2=P2L−23x,Q3=P3L−25x2+32.
Arfken, Weber and Harris write the polynomial part as
Rn=1⋅n2n−1Pn−1+3(n−1)2n−5Pn−3+⋯, which is the same polynomial resolved in the Legendre basis.
Logarithmic singularity.Rn(1)=Hn, and as x→1−Qn(x)=21ln1−x2−Hn+O((1−x)ln(1−x)).
By the parity in part 4 the behaviour at the other end is
Qn(x)=(−1)n+1[21ln1+x2−Hn]+O((1+x)ln(1+x)) as x→−1+.
So Qn is unbounded at both ends for every n — Arfken, Weber and Harris record this as Qn(1)=∞ — while remaining square integrable on (−1,1), the singularity being only logarithmic.
Recurrences. The Qn satisfy the same relations as the Pn:
(n+1)Qn+1=(2n+1)xQn−nQn−1,(2n+1)Qn=Qn+1′−Qn−1′,
together with (1−x2)Qn′=n(Qn−1−xQn) and the rest of the family in Legendre polynomials.
Parity and values at the origin.Qn(−x)=(−1)n+1Qn(x): the Qn of even degree are odd functions and conversely. Hence Q2p(0)=0, while
Q2p+1(0)=(−1)p+1(2p+1)!!(2p)!!.
There is no singularity at x=0, which is an ordinary point of (L).
Wronskian. On (−1,1),
W[Pn,Qn](x)=Pn(x)Qn′(x)−Pn′(x)Qn(x)=1−x21.
Arfken, Weber and Harris state this in the form An/(1−x2) with An independent of x; with the normalisation above An=1 for every n. The Wronskian never vanishes, so {Pn,Qn} is a basis of the solution space of (L) at ℓ=n: every solution on (−1,1) is c1Pn+c2Qn. The companion contiguous relation of Arfken, Weber and Harris is
n[PnQn−1−Pn−1Qn]=1.
The normalising constants.αn=1/Pn(0) for even n and βn=−1/Pn′(0) for odd n; these are precisely the values that make An=1 in part 5.
Outside the interval. With the convention L↦21lnx−1x+1 for x>1, each Qn is finite and positive on (1,∞), decreasing to 0; Arfken, Weber and Harris obtain the expansion in inverse powers of x whose leading term is
Qn(x)∼(2n+1)!!n!x−n−1(x→∞).
Idea of the proofOriginal
Everything follows from one generating function. Put G(x,h)=(1−2xh+h2)−1/2=∑nPnhn and look for a second family packed the same way. Two observations do it.
First, PnL almost solves Legendre's equation: because (1−x2)L′=1, substituting PnL into (L) leaves the clean residue 2Pn′ and nothing else. So the second solution must be PnL corrected by whatever cancels 2Pn′, and because 2Pn′ 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 G multiplied by A(x,h)=∫0hG(x,s)ds, an integral that is elementary — complete the square under the square root — so the whole family ∑Qnhn=G⋅[L−A] is available in closed form. Differentiating that product in x and in h gives the two recurrences exactly as for the Pn; reading off the coefficient of hn gives the closed form with Rn=∑mm1Pm−1Pn−m; and setting x=1 in Rn gives ∑mm1=Hn because every Pj(1)=1. That last line is the whole content of the logarithmic singularity.
The Wronskian is separate and cheaper: (1−x2)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−, where the polynomial pieces are killed by the factor 1−x2 and only Pn(1)2=1 survives.
Proof(The generating function of the second solutions)Original
Write G=(1−2xh+h2)−1/2, and let
M[Ψ]:=∂x[(1−x2)Ψx]+∂h[h2Ψh].
If Ψ=∑n≥0ψn(x)hn then ∂h[h2Ψh]=∑nn(n+1)ψnhn, so
M[Ψ]=n≥0∑{[(1−x2)ψn′]′+n(n+1)ψn}hn:(1)
M[Ψ]=0 says precisely that every coefficient ψn solves (L) with ℓ=n. For G itself this is the statement proved in Legendre polynomials, so we may use M[G]=0. We shall also use the two elementary derivatives
Gx=hG3,Gh=(x−h)G3,(2)
each obtained by differentiating (1−2xh+h2)−1/2 once.
Step 1: the candidate. Set
A(x,h):=∫0h1−2xs+s2ds,Λ:=L−A,Φ:=GΛ.
Completing the square, 1−2xs+s2=(s−x)2+(1−x2), so the integral is a standard inverse hyperbolic sine and
Step 2: Φ is analytic in h on the unit disc. For ∣x∣<1 the roots of 1−2xh+h2 in h are x±i1−x2, both of modulus 1, so 1−2xh+h2 has an analytic branch on ∣h∣<1 equal to 1 at h=0. The denominator in (3) cannot vanish there: h−x+1−2xh+h2=0 would force 1−2xh+h2=(x−h)2, that is x2=1. Being non-vanishing on a disc and equal to 1−x>0 at h=0, it admits a single-valued analytic logarithm. Hence Φ=∑n≥0qn(x)hn with radius of convergence at least 1.
Step 3: the two derivatives of Λ. Differentiating (3) in h recovers the integrand, Λh=−Ah=−G. For the x-derivative, write u=1−2xh+h2 and differentiate the logarithm:
Λx=1−x2−x+u(h−x+u)u+h.
Multiply numerator and denominator of the second term by u−h+x. The denominator becomes u[u−(h−x)2]=u(1−x2), and the numerator becomes u+xu−h2+hx=1−xh+xu. The two terms in x cancel and
Λh=−G,Λx=1−x2(1−xh)G.(4)
Step 4: M[Φ]=0. For any product GΛ, expanding each derivative by the Leibniz rule and collecting,
M[GΛ]=ΛM[G]+2{(1−x2)GxΛx+h2GhΛh}+GM[Λ].
The first term vanishes. For the second, (2) and (4) give
(1−x2)GxΛx=h(1−xh)G4,h2GhΛh=−h2(x−h)G4,
and since (1−xh)−h(x−h)=1−2xh+h2=G−2, their sum is hG2; the second term is therefore 2hG2. For the third, (4) gives (1−x2)Λx=(1−xh)G and h2Λh=−h2G, so
Hence GM[Λ]=−2hG2, the two surviving terms cancel, and M[Φ]=0. By (1), every coefficient qn solves (L) with ℓ=n.□
Proof(Closed form, recurrences, parity and the singularity)Original
We now read the properties off Φ=GΛ=LG−GA.
Closed form (part 1). Integrating G=∑k≥0Pkhk term by term from 0 to h,
A(x,h)=k≥0∑k+1Pk(x)hk+1=m≥1∑mPm−1(x)hm.
Multiplying this by G and taking the coefficient of hn in Φ=LG−GA,
qn=PnL−Rn,Rn=m=1∑nm1Pm−1Pn−m(R0=0).
Each summand has degree (m−1)+(n−m)=n−1, and the leading coefficients of Legendre polynomials are positive, so the leading coefficients cannot cancel: Rn is a polynomial of degree exactly n−1 for n≥1. Taking n=0,1,2,3 gives R0=0, R1=1, R2=P0P1+21P1P0=23x and R3=P2+21x2+31P2=25x2−32, which are the four displayed cases.
The three-term recurrence (part 3). From (2) and (4),
Φh=GhΛ+GΛh=(x−h)G3Λ−G2,
and multiplying by 1−2xh+h2=G−2,
(1−2xh+h2)Φh=(x−h)Φ−1.(5)
This is the identity that produces the recurrence for the Pn, except for the extra constant −1 on the right, which only affects the coefficient of h0. Comparing coefficients of hn with n≥1,
(n+1)qn+1−2xnqn+(n−1)qn−1=xqn−qn−1,
which rearranges to (n+1)qn+1=(2n+1)xqn−nqn−1. At n=0 the same comparison gives q1=xq0−1.
The derivative recurrence (part 3). In the same way,
Φx=GxΛ+GΛx=hG3Λ+(1−xh)G2/(1−x2), so
(1−2xh+h2)Φx=hΦ+1−x21−xh.(6)
For n≥2 the coefficient of hn reads qn′−2xqn−1′+qn−2′=qn−1, which is the relation the Pn satisfy; eliminating between it and the three-term recurrence exactly as in Legendre polynomials yields (2n+1)qn=qn+1′−qn−1′ and (1−x2)qn′=n(qn−1−xqn).
Parity and the origin (part 4).L is odd and Pj(−x)=(−1)jPj(x), so every summand of Rn has parity (−1)n−1 and
For even n this makes qn odd, so qn(0)=0. The odd indices are linked to one another by taking neven in the three-term recurrence: with n=2p and x=0 the middle term drops out and (2p+1)q2p+1(0)=−2pq2p−1(0), so from q1(0)=−1,
which is Arfken, Weber and Harris's value. (For p=1 and p=2 this is 32 and −158.)
The logarithmic singularity (part 2). Since Pj(1)=1 for every j,
Rn(1)=m=1∑nm1⋅1⋅1=Hn,
with no computation at all. Now split L as L(x)=21ln1−x2+21ln21+x. As x→1− the second piece is O(1−x), and Pn(x)−1=O(1−x) because Pn is a polynomial with Pn(1)=1; hence
Pn(x)L(x)=21ln1−x2+O((1−x)ln(1−x)).
Likewise Rn(x)=Hn+O(1−x). Subtracting gives part 2, and the statement at x=−1 follows from the parity. In particular ∣qn(x)∣→∞ at both ends, while ∫−11qn2 converges because ln2 is integrable. □
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 is constant. Let u,v be any two solutions of (L) at ℓ=n and W=uv′−u′v. Then W′=uv′′−u′′v and, using (L) for each of u and v to replace (1−x2)v′′ and (1−x2)u′′,
(1−x2)W′=u[2xv′−n(n+1)v]−[2xu′−n(n+1)u]v=2xW.
Hence dxd[(1−x2)W]=(1−x2)W′−2xW=0. This is Arfken, Weber and Harris's observation that pW is constant for a self-adjoint equation, here with p=1−x2; their Exercise 7.6.9 states the Legendre case as W=An/(1−x2).
Step 2: the constant is 1. Take u=Pn, v=qn=PnL−Rn. Using (1−x2)L′=1,
the terms carrying L having cancelled. The right-hand side is a polynomial, so we may evaluate it at x=1: the bracket is finite there and is multiplied by 1−x2=0, while Pn(1)2=1. Hence An=1 for every n and
Pnqn′−Pn′qn=1−x21.
Since this never vanishes on (−1,1), Pn and qn 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]=1 of Arfken, Weber and Harris is the same statement rewritten with the recurrences of part 3.
Step 3: qn=Qn. Both qn and Qn solve (L) at ℓ=n, and both have parity (−1)n+1 — for Qn by its definition as the non-terminating series, for qn by part 4. Since Pn has the opposite parity, a solution of definite parity (−1)n+1 is unique up to scale, so qn and Qn are proportional and it is enough to match one value at x=0.
Let n be even. Then Qn=αnyo with yo(0)=0, yo′(0)=1, so Qn(0)=0 and Qn′(0)=αn; also Pn′(0)=0. Putting x=0 in Step 2 for the pair Pn,Qn — whose Wronskian constant we may call An — gives An=Pn(0)αn. Now Legendre polynomials gives P2m(0)=(−1)m(2m)!/[22m(m!)2], that is Pn(0)=(−1)n/2n!/(2n[(n/2)!]2), which is exactly 1/αn. So An=1: Riley, Hobson and Bence's constant αn is the one that normalises the Wronskian, αn=1/Pn(0), and qn and Qn — proportional, with the same Wronskian against Pn — coincide.
Let n be odd. Then Qn=βnye with ye(0)=1, ye′(0)=0, so Qn(0)=βn, Qn′(0)=0, and Pn(0)=0; the Wronskian at the origin is −Pn′(0)βn. Setting x=0 in the derivative recurrence (1−x2)Pn′=n(Pn−1−xPn) gives Pn′(0)=nPn−1(0), so with n=2p+1,
the last equality being the definition of βn with (n−1)/2=p and (−1)(n+1)/2=(−1)p+1. Hence the Wronskian constant is again 1 and qn=Qn. This proves part 6 and completes the identification. □
Remark(Where the hypotheses are used)Original
n 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 ℓboth series are infinite and both blow up at x=±1; the functions Pλ, Qλ of Hassani's hypergeometric formulas still exist, but nothing in this unit's proofs survives, because Rn was a polynomial precisely because G generates polynomials.
∣x∣<1. Used three times. The series defining Qn converges only there. The generating function argument needs ∣x∣≤1 so that the singularities of G in h sit on ∣h∣=1, and needs ∣x∣<1 strictly so that h−x+u does not vanish — at x=±1 it does, which is the analytic shadow of the singularity of Qn at the ends. And L=artanhx is real only there; beyond x=1 one must switch to 21lnx−1x+1, which differs by 21ln(−1) and so changes the solution by a multiple of Pn — a real change of function, not a change of notation.
Regularity at the ends is what selects Pn. Part 2 is the general-n version of the n=0 calculation in Legendre polynomials: the obstruction is not a failure of square integrability, since ∫−11Qn2 converges for every n, but the survival of the boundary form p(uv′−u′v), with p=1−x2, at the endpoints x=±1: 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=Pn, v=Qn the boundary form equals 1 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=0 in y=c1Pn+c2Qn, and hence why the separation constant is quantised.
Where the normalisation enters. Nothing in parts 1–4 depends on the scale of Qn except through the starting values Q0=L, Q1=xL−1. The content of the constants αn,βn is exactly part 6: they are the unique choice making the Wronskian constant equal 1 for every degree, and therefore the unique choice under which Pn and Qn obey one and the same recurrence. A different scaling for each n — for instance normalising Qn to have unit value at x=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 Pn; what pins that multiple
down is the parity, since Pn and Qn have opposite parities and so no multiple of Pn
can be added without destroying it. Also, Pn has n simple zeros inside (−1,1), so the
integrand of (R) is not locally integrable at them; Qn itself nevertheless has a
finite limit at each, because the double pole of Pn−2 integrates to a simple pole in the
quotient Qn/Pn — which is what (R) computes — and the prefactor Pn cancels exactly
that pole. It is the quotient, not Qn, that blows up there. We have taken the generating-function route instead of (R) partly for
this reason: it never divides by Pn and so never meets either difficulty. Arfken, Weber and
Harris state that (R) fixes the scale and the form of Qn without comment on either point.
Rn(1)=Hn needs Pj(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−1. The claim that the leading coefficients in Rn=∑mm1Pm−1Pn−m cannot cancel uses that all of them are positive, which is read off Rodrigues' formula. Without that, Rn could in principle drop degree, and the asymptotic statements in part 7 would need separate proof.
Method(Computing with Qn)Original
On (−1,1): set L=artanhx, take Q0=L and Q1=xL−1, and iterate upward,
Qn+1=n+1(2n+1)xQn−nQn−1.
This is stable, and for the same reason as for Pn: on (−1,1) the two solutions of the recurrence are of comparable size (both decay only like n−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 L and the polynomial Rn separately and subtract — rather than letting L and PnL both overflow.
On (1,∞): do not iterate upward. There Pn grows like (x+x2−1)n while Qn decays like its reciprocal, so an upward sweep for Qn is contaminated by the dominant solution and loses everything. At x=5 in double precision the upward recurrence already gives Q10 wrong by four orders of magnitude and Q20 wrong by twenty-three. Instead either use the closed form of part 1 with L replaced by 21lnx−1x+1, or run the recurrence downward from a high starting index with arbitrary seed values and rescale at the end using the known Q0: downward iteration is stable exactly when the wanted solution is the decaying one.
To obtain the polynomial part Rn in closed form: use Rn=∑m=1nm1Pm−1Pn−m directly, or iterate Rn+1=[(2n+1)xRn−nRn−1]/(n+1) from R0=0, R1=1 — the same recurrence, since PnL satisfies it on its own. Check the result with Rn(1)=Hn, which catches almost every arithmetic slip.
To solve a boundary-value problem for (L) on [a,b]⊂(−1,1): write the general solution c1Pn+c2Qn and impose the two conditions. Because part 5 gives pW=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+bQn and v=cPn+dQn carries the denominator pW[u,v]=(ad−bc)pW[Pn,Qn]=ad−bc, a number fixed once by the two boundary conditions rather than a determinant to be evaluated at every point.
For the associated equation with m=0, use Qnm=(1−x2)m/2dmQn/dxm, the second-kind counterpart of the construction in Associated Legendre functions; these are again singular at x=±1 and again discarded on the full sphere.
Example(Building Q3, and four checks)Original
Step 1: the polynomial part. From R0=0, R1=1 and Rn+1=[(2n+1)xRn−nRn−1]/(n+1),
R2=23x⋅1−1⋅0=23x,R3=35x⋅23x−2⋅1=25x2−32.
The sum formula agrees: R3=P0P2+21P1P1+31P2P0=34⋅23x2−1+21x2=25x2−32. Hence, with L=artanhx,
Q3(x)=21(5x3−3x)L(x)−25x2+32.
Check 1 (the harmonic number).R3(1)=25−32=615−4=611, and H3=1+21+31=611. Agreed.
The logarithms cancel exactly, and 34=1/(1−41) as part 5 requires.
Check 4 (the singularity). Part 2 predicts Q3(x)≈21ln1−x2−611 near x=1. At x=1−10−5 that is 21ln(2×105)−1.8333333=6.1030363−1.8333333=4.2697030, while the closed form gives 4.2693843. The discrepancy is 3.2×10−4, of the predicted size O((1−x)ln(1−x))=1.2×10−4 up to a constant of order one. Every number above was verified symbolically and in extended precision.
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≤π, so x=cosθ reaches both ±1, and the general separated solution c1Pn+c2Qn must have c2=0 because Qn is infinite there. Riley, Hobson and Bence take this step at the start of their treatment of Pn (Separation of variables in standard coordinates). Part 5 is what makes it rigorous rather than cosmetic: it is not that Qn is "badly behaved", it is that the bilinear boundary form p(uv′−u′v) for the pair Pn,Qn equals exactly 1 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 Qn is gone, so both solutions survive and the degree is no longer forced to be an integer. Physically the Q-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 Qn is the answer rather than the discard. Prolate spheroidal coordinates based on foci at z=±c use ξ=(r1+r2)/2c≥1 and η=(r1−r2)/2c∈[−1,1], where r1,r2 are the distances to the two foci. Laplace's equation separates with Pn(η) in the angular variable and, in ξ, the same Legendre equation again — now on (1,∞), where Pn grows and Qn decays. An exterior field therefore expands as ∑nanPn(η)Qn(ξ), Arfken, Weber and Harris's property Qn(∞)=0 being exactly the radiation-free condition at infinity. The simplest case is a uniformly charged rod of half-length c and total charge q along the axis. Integrating 1/∣r−r′∣ over the rod gives the elementary closed form lnr1+r2−2cr1+r2+2c, and in these coordinates r1+r2=2cξ, so
V=4πε0cq⋅21lnξ−1ξ+1=4πε0cqQ0(ξ).
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/ξ, and ξ→r/c far away, so V→q/(4πε0r) as it must; and the equipotentials ξ=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) do the same job for the multipole terms of a spheroidal conductor, the spheroidal analogue of the r−n−1 of Multipole expansions.
Green's functions. Part 5 says pW[Pn,Qn]=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) is
G(x,s)=pW[u,v]u(x<)v(x>)
with u,v the combinations of Pn,Qn satisfying the left and right conditions; the prefactor 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 Pn regular at the inner end and Qn decaying at the outer, is what makes Pn(x<)Qn(x>) the standard kernel in exterior and annular problems.