Clebsch–Gordan coefficients

From Wikipedia, the free encyclopedia

In physics, the Clebsch–Gordan coefficients are sets of numbers that arise in angular momentum coupling under the laws of quantum mechanics.

In more mathematical terms, the CG coefficients are used in representation theory, particularly of compact Lie groups, to perform the explicit direct sum decomposition of the tensor product of two irreducible representations into irreducible representations, in cases where the numbers and types of irreducible components are already known abstractly. The name derives from the German mathematicians Alfred Clebsch (1833–1872) and Paul Gordan (1837–1912), who encountered an equivalent problem in invariant theory.

In terms of classical mathematics, the CG coefficients, or at least those associated to the group SO(3), may be defined much more directly, by means of formulae for the multiplication of spherical harmonics. The addition of spins in quantum-mechanical terms can be read directly from this approach. The formulas below use Dirac's bra-ket notation.

Clebsch–Gordan coefficients are the expansion coefficients of total angular momentum eigenstates in an uncoupled tensor product basis.

Below, this definition is made precise by defining angular momentum operators, angular momentum eigenstates, and tensor products of these states.

From the formal definition of angular momentum, recursion relations for the Clebsch–Gordan coefficients can be found. To find numerical values for the coefficients a phase convention must be adopted. Below the Condon–Shortley phase convention is chosen.

Angular momentum operators

Angular momentum operators are self-adjoint operators ${\mathrm {j}}_{x}$ , ${\mathrm {j}}_{y}$ , and ${\mathrm {j}}_{z}$ that satisfy the commutation relations

$[{\mathrm {j}}_{k},{\mathrm {j}}_{l}]\equiv {\mathrm {j}}_{k}{\mathrm {j}}_{l}-{\mathrm {j}}_{l}{\mathrm {j}}_{k}=i\hbar \sum _{m}\varepsilon _{{klm}}{\mathrm {j}}_{m},\quad {\mathrm {where}}\quad k,l,m\in (x,y,z)$

where $\varepsilon _{{klm}}$ is the Levi-Civita symbol. Together the three operators define a "vector operator":

${\mathbf {j}}=({\mathrm {j}}_{x},{\mathrm {j}}_{y},{\mathrm {j}}_{z})$

By developing this concept further, one can define an operator as an "inner product" of ${{\mathbf j}}$ with itself:

${\mathbf {j}}^{2}={\mathrm {j}}_{x}^{2}+{\mathrm {j}}_{y}^{2}+{\mathrm {j}}_{z}^{2}.\,$

It is an example of a Casimir operator.

We also define raising ( ${\mathrm {j}}_{{+}}$ ) and lowering ( ${\mathrm {j}}_{{-}}$ ) operators:

${\mathrm {j}}_{\pm }={\mathrm {j}}_{x}\pm i{\mathrm {j}}_{y}.\,$

Angular momentum states

It can be shown from the above definitions that ${\mathbf {j}}^{2}$ commutes with ${\mathrm {j}}_{x}$ , ${\mathrm {j}}_{y}$ and ${\mathrm {j}}_{z}$

$[{\mathbf {j}}^{2},{\mathrm {j}}_{k}]=0\ {\mathrm {for}}\ k=x,y,z$

When two Hermitian operators commute, a common set of eigenfunctions exists. Conventionally ${\mathbf {j}}^{2}$ and ${\mathrm {j}}_{z}$ are chosen. From the commutation relations the possible eigenvalues can be found. The result is:

${\begin{alignedat}{2}{\mathbf {j}}^{2}|j\,m\rangle =\hbar ^{2}j(j+1)|j\,m\rangle &\;\;\;j=0,{\frac {1}{2}},1,{\frac {3}{2}},2,\ldots \\{\mathrm {j}}_{z}|j\,m\rangle =\hbar m|j\,m\rangle &\;\;\;m=-j,-j+1,\ldots ,j.\end{alignedat}}$

The raising and lowering operators change the value of $m$ :

${\mathrm {j}}_{\pm }|j\,m\rangle =\hbar C_{\pm }(j,m)|j\,m\pm 1\rangle$

with

$C_{\pm }(j,m)={\sqrt {j(j+1)-m(m\pm 1)}}={\sqrt {(j\mp m)(j\pm m+1)}}.$

A (complex) phase factor could be included in the definition of $C_{\pm }(j,m)$ . The choice made here is in agreement with the Condon and Shortley phase conventions. The angular momentum states must be orthogonal (because their eigenvalues with respect to a Hermitian operator are distinct) and are assumed to be normalized:

$\langle j_{1}\,m_{1}|j_{2}\,m_{2}\rangle =\delta _{{j_{1},j_{2}}}\delta _{{m_{1},m_{2}}}.$

Note that the italicized $j_{1}$ and $j_{2}$ denote integers or half-integers, which label the total angular momentum of the system (e.g. $j_{1}=1/2$ for an electron and $j_{1}=1$ for a photon). On the other hand, the roman ${\mathrm {j}}_{x},{\mathrm {j}}_{y},{\mathrm {j}}_{z},{\mathrm {j}}_{+},{\mathrm {j}}_{-}$ and ${\mathbf {j}}^{2}$ denote operators (the last being a vector operator).

Tensor product space

Let $V_{1}$ be the $(2j_{1}+1)$ -dimensional vector space spanned by the states

$|j_{1}m_{1}\rangle ,\quad m_{1}=-j_{1},-j_{1}+1,\ldots ,j_{1},$

and $V_{2}$ the $(2j_{2}+1)$ -dimensional vector space spanned by

$|j_{2}m_{2}\rangle ,\quad m_{2}=-j_{2},-j_{2}+1,\ldots ,j_{2}.$

The tensor product of these spaces, $V_{{12}}\equiv V_{1}\otimes V_{2}$ , has a $(2j_{1}+1)(2j_{2}+1)$ -dimensional uncoupled basis

$|j_{1}m_{1}\rangle |j_{2}m_{2}\rangle \equiv |j_{1}m_{1}\rangle \otimes |j_{2}m_{2}\rangle ,\quad m_{1}=-j_{1},\ldots j_{1},\quad m_{2}=-j_{2},\ldots j_{2}.$

Angular momentum operators acting on $V_{{12}}$ can be defined by

$({\mathrm {j}}_{i}\otimes 1)|j_{1}m_{1}\rangle |j_{2}m_{2}\rangle \equiv ({\mathrm {j}}_{i}|j_{1}m_{1}\rangle )\otimes |j_{2}m_{2}\rangle$

and

$(1\otimes {\mathrm {j}}_{i})|j_{1}m_{1}\rangle |j_{2}m_{2}\rangle \equiv |j_{1}m_{1}\rangle \otimes ({\mathrm {j}}_{i}|j_{2}m_{2}\rangle )\quad {\mathrm {for}}\quad i=x,y,z.$

Total angular momentum operators are defined by

${\mathrm {J}}_{i}={\mathrm {j}}_{i}\otimes 1+1\otimes {\mathrm {j}}_{i}\quad {\mathrm {for}}\quad i=x,y,z.$

The total angular momentum operators satisfy the required commutation relations

$[{\mathrm {J}}_{k},{\mathrm {J}}_{l}]=i\hbar \epsilon _{{klm}}{\mathrm {J}}_{m},\quad {\mathrm {where}}\quad k,l,m\in (x,y,z),$

and hence total angular momentum eigenstates exist:

${\begin{aligned}{\mathbf {J}}^{2}|(j_{1}j_{2})JM\rangle &=\hbar ^{2}J(J+1)|(j_{1}j_{2})JM\rangle ,\\{\mathrm {J}}_{z}|(j_{1}j_{2})JM\rangle &=\hbar M|(j_{1}j_{2})JM\rangle ,\quad {\mathrm {for}}\quad M=-J,\ldots ,J.\end{aligned}}$

It can be derived that the total angular momentum quantum number $J$ must satisfy the triangular condition

$|j_{1}-j_{2}|\leq J\leq j_{1}+j_{2}.$

The total number of total angular momentum eigenstates is equal to the dimension of $V_{{12}}$ :

$\sum _{{J=|j_{1}-j_{2}|}}^{{j_{1}+j_{2}}}(2J+1)=(2j_{1}+1)(2j_{2}+1).$

The total angular momentum states form an orthonormal basis of $V_{{12}}$ :

$\langle J_{1}M_{1}|J_{2}M_{2}\rangle =\delta _{{J_{1}J_{2}}}\delta _{{M_{1}M_{2}}}.$

Formal definition of Clebsch–Gordan coefficients

The total angular momentum states can be expanded with the use of the completeness relation in the uncoupled basis

$|(j_{1}j_{2})JM\rangle =\sum _{{m_{1}=-j_{1}}}^{{j_{1}}}\sum _{{m_{2}=-j_{2}}}^{{j_{2}}}|j_{1}m_{1}j_{2}m_{2}\rangle \langle j_{1}m_{1}j_{2}m_{2}|JM\rangle$

The expansion coefficients $\langle j_{1}m_{1}j_{2}m_{2}|JM\rangle$ are called Clebsch–Gordan coefficients.

Applying the operator

${\mathrm {J}}_{z}={\mathrm {j}}_{z}\otimes 1+1\otimes {\mathrm {j}}_{z}$

to both sides of the defining equation shows that the Clebsch–Gordan coefficients can only be nonzero when

$M=m_{1}+m_{2}.\,$

Recursion relations

The recursion relations were discovered by physicist Giulio Racah from the Hebrew University of Jerusalem in 1941.

Applying the total angular momentum raising and lowering operators

${\mathrm {J}}_{\pm }={\mathrm {j}}_{\pm }\otimes 1+1\otimes {\mathrm {j}}_{\pm }$

to the left hand side of the defining equation gives

${\mathrm {J}}_{\pm }|(j_{1}j_{2})JM\rangle =\hbar C_{\pm }(J,M)|(j_{1}j_{2})JM\pm 1\rangle =\hbar C_{\pm }(J,M)\sum _{{m_{1}m_{2}}}|j_{1}m_{1}\rangle |j_{2}m_{2}\rangle \langle j_{1}m_{1}j_{2}m_{2}|JM\pm 1\rangle .$

Applying the same operators to the right hand side gives

${\begin{aligned}{\mathrm {J}}_{\pm }&\sum _{{m_{1}m_{2}}}|j_{1}m_{1}\rangle |j_{2}m_{2}\rangle \langle j_{1}m_{1}j_{2}m_{2}|JM\rangle \\&=\hbar \sum _{{m_{1}m_{2}}}\left[C_{\pm }(j_{1},m_{1})|j_{1}m_{1}\pm 1\rangle |j_{2}m_{2}\rangle +C_{\pm }(j_{2},m_{2})|j_{1}m_{1}\rangle |j_{2}m_{2}\pm 1\rangle \right]\langle j_{1}m_{1}j_{2}m_{2}|JM\rangle \\&=\hbar \sum _{{m_{1}m_{2}}}|j_{1}m_{1}\rangle |j_{2}m_{2}\rangle \left[C_{\pm }(j_{1},m_{1}\mp 1)\langle j_{1}{m_{1}\mp 1}j_{2}m_{2}|JM\rangle +C_{\pm }(j_{2},m_{2}\mp 1)\langle j_{1}m_{1}j_{2}{m_{2}\mp 1}|JM\rangle \right].\end{aligned}}$

where

$C_{\pm }(j,m)={\sqrt {j(j+1)-m(m\pm 1)}}$

Combining these results gives recursion relations for the Clebsch–Gordan coefficients

$C_{\pm }(J,M)\langle j_{1}m_{1}j_{2}m_{2}|JM\pm 1\rangle =C_{\pm }(j_{1},m_{1}\mp 1)\langle j_{1}{m_{1}\mp 1}j_{2}m_{2}|JM\rangle +C_{\pm }(j_{2},m_{2}\mp 1)\langle j_{1}m_{1}j_{2}{m_{2}\mp 1}|JM\rangle .$

Taking the upper sign with $M=J$ gives

$0=C_{+}(j_{1},m_{1}-1)\langle j_{1}{m_{1}-1}j_{2}m_{2}|JJ\rangle +C_{+}(j_{2},m_{2}-1)\langle j_{1}m_{1}j_{2}m_{2}-1|JJ\rangle .$

In the Condon and Shortley phase convention the coefficient $\langle j_{1}j_{1}j_{2}J-j_{1}|JJ\rangle$ is taken real and positive. With the last equation all other Clebsch–Gordan coefficients $\langle j_{1}m_{1}j_{2}m_{2}|JJ\rangle$ can be found. The normalization is fixed by the requirement that the sum of the squares, which corresponds to the norm of the state $|(j_{1}j_{2})JJ\rangle$ must be one.

The lower sign in the recursion relation can be used to find all the Clebsch–Gordan coefficients with $M=J-1$ . Repeated use of that equation gives all coefficients.

This procedure to find the Clebsch–Gordan coefficients shows that they are all real (in the Condon and Shortley phase convention).

Explicit expression

For an explicit expression of the Clebsch–Gordan coefficients and tables with numerical values, see table of Clebsch–Gordan coefficients.

Orthogonality relations

These are most clearly written down by introducing the alternative notation

$\langle JM|j_{1}m_{1}j_{2}m_{2}\rangle \equiv \langle j_{1}m_{1}j_{2}m_{2}|JM\rangle$

The first orthogonality relation is

$\sum _{{J=|j_{1}-j_{2}|}}^{{j_{1}+j_{2}}}\sum _{{M=-J}}^{{J}}\langle j_{1}m_{1}j_{2}m_{2}|JM\rangle \langle JM|j_{1}m_{1}'j_{2}m_{2}'\rangle =\langle j_{1}m_{1}j_{2}m_{2}|j_{1}m_{1}'j_{2}m_{2}'\rangle =\delta _{{m_{1},m_{1}'}}\delta _{{m_{2},m_{2}'}}$

(using the completeness relation that $1\equiv \sum _{x}|x\rangle \langle x|$ )and the second

$\sum _{{m_{1}m_{2}}}\langle JM|j_{1}m_{1}j_{2}m_{2}\rangle \langle j_{1}m_{1}j_{2}m_{2}|J'M'\rangle =\langle JM|J'M'\rangle =\delta _{{J,J'}}\delta _{{M,M'}}.$

Special cases

For $J=0$ the Clebsch–Gordan coefficients are given by

$\langle j_{1},m_{1};j_{2},m_{2}|00\rangle =\delta _{{j_{1},j_{2}}}\delta _{{m_{1},-m_{2}}}{\frac {(-1)^{{j_{1}-m_{1}}}}{{\sqrt {2j_{2}+1}}}}.$

For $J=j_{1}+j_{2}$ and $M=J$ we have

$\langle j_{1},j_{1};j_{2},j_{2}|j_{1}+j_{2},j_{1}+j_{2}\rangle =1.$

For $j_{1}=j_{2}=J/2$ and $m_{2}={-m_{1}}$ we have

$\langle j_{1},m_{1};j_{1},{-m_{1}}|2j_{1},0\rangle ={\frac {(2j_{1})!^{2}}{(j_{1}-m_{1})!(j_{1}+m_{1})!{\sqrt {(4j_{1})!}}}}.$

For $j_{1}=j_{2}=m_{1}={-m_{2}}$ we have

$\langle j_{1},j_{1};j_{1},{-j_{1}}|J,0\rangle =(2j_{1})!{\sqrt {{\frac {2J+1}{(J+2j_{1}+1)!(2j_{1}-J)!}}}}.$

For $j_{2}=1,m_{2}=0$ we have

${\begin{aligned}\langle j_{1},m;1,0|j_{1}+1,m\rangle &={\sqrt {{\frac {(j_{1}-m+1)(j_{1}+m+1)}{(2j_{1}+1)(j_{1}+1)}}}},\\\langle j_{1},m;1,0|j_{1},m\rangle &={\frac {m}{{\sqrt {j_{1}(j_{1}+1)}}}},\\\langle j_{1},m;1,0|j_{1}-1,m\rangle &=-{\sqrt {{\frac {(j_{1}-m)(j_{1}+m)}{j_{1}(2j_{1}+1)}}}}.\end{aligned}}$

Symmetry properties

${\begin{aligned}\langle j_{1}m_{1}j_{2}m_{2}|JM\rangle &=(-1)^{{j_{1}+j_{2}-J}}\langle j_{1}\,{-m_{1}}j_{2}\,{-m_{2}}|J\,{-M}\rangle \\&=(-1)^{{j_{1}+j_{2}-J}}\langle j_{2}m_{2}j_{1}m_{1}|JM\rangle \\&=(-1)^{{j_{1}-m_{1}}}{\sqrt {{\frac {2J+1}{2j_{2}+1}}}}\langle j_{1}m_{1}J\,{-M}|j_{2}\,{-m_{2}}\rangle \\&=(-1)^{{j_{2}+m_{2}}}{\sqrt {{\frac {2J+1}{2j_{1}+1}}}}\langle J\,{-M}j_{2}m_{2}|j_{1}\,{-m_{1}}\rangle \\&=(-1)^{{j_{1}-m_{1}}}{\sqrt {{\frac {2J+1}{2j_{2}+1}}}}\langle JMj_{1}\,{-m_{1}}|j_{2}m_{2}\rangle \\&=(-1)^{{j_{2}+m_{2}}}{\sqrt {{\frac {2J+1}{2j_{1}+1}}}}\langle j_{2}\,{-m_{2}}JM|j_{1}m_{1}\rangle \end{aligned}}$

A convenient way to derive these relations is by converting the Clebsch–Gordan coefficients to 3-jm symbols using the equation given below. The symmetry properties of 3-jm symbols are much simpler. Care is needed when simplifying phase factors, because the quantum numbers can be integer or half integer, e.g., $(-1)^{{2j}}$ is equal to 1 for integer $j$ and equal to −1 for half-integer $j$ . The following relations, however, are valid in either case:

$(-1)^{{4j}}=(-1)^{{2(j-m)}}=1$

and for $j_{1}$ , $j_{2},$ and $J$ appearing in the same Clebsch–Gordan coefficient:

$(-1)^{{2(j_{1}+j_{2}+J)}}=(-1)^{{2(m_{1}+m_{2}+M)}}=1.$

Relation to 3-jm symbols

Clebsch–Gordan coefficients are related to 3-jm symbols which have more convenient symmetry relations.

$\langle j_{1}m_{1}j_{2}m_{2}|j_{3}m_{3}\rangle =(-1)^{{j_{1}-j_{2}+m_{3}}}{\sqrt {2j_{3}+1}}{\begin{pmatrix}j_{1}&j_{2}&j_{3}\\m_{1}&m_{2}&-m_{3}\end{pmatrix}}.$

Relation to Wigner D-matrices

$\int _{0}^{{2\pi }}d\alpha \int _{0}^{\pi }\sin \beta d\beta \int _{0}^{{2\pi }}d\gamma D_{{MK}}^{J}(\alpha ,\beta ,\gamma )^{\ast }D_{{m_{1}k_{1}}}^{{j_{1}}}(\alpha ,\beta ,\gamma )D_{{m_{2}k_{2}}}^{{j_{2}}}(\alpha ,\beta ,\gamma )={\frac {8\pi ^{2}}{2J+1}}\langle j_{1}m_{1}j_{2}m_{2}|JM\rangle \langle j_{1}k_{1}j_{2}k_{2}|JK\rangle .$

Other Properties

$\sum _{m}(-1)^{{j-m}}\langle jmj{-m}|J0\rangle ={\sqrt {2j+1}}~\delta _{{J0}}$

SU(N) Clebsch–Gordan coefficients

For arbitrary groups and their representations, Clebsch–Gordan coefficients are not known in general. However, algorithms to produce Clebsch–Gordan coefficients for the Special unitary group are known.^[1] ^[2] A web interface for tabulating SU(N) Clebsch–Gordan coefficients is readily available.

References

↑ Alex, A.; M. Kalus, A. Huckleberry, and J. von Delft (February 2011). "A numerical algorithm for the explicit calculation of SU(N) and SL(N,C) Clebsch–Gordan coefficients". J. Math. Phys. 82: 023507. arXiv:1009.0437. Bibcode:2011JMP....52b3507A. doi:10.1063/1.3521562. Cite uses deprecated parameters (help)
↑ Kaplan, L. M.; Resnikoff, M. (1967). "Matrix products and explicit 3, 6, 9, and 12j coefficients of the regular representation of SU(n)". J. Math. Phys. 8: 2194. doi:10.1063/1.1705141.