Chiral model

In nuclear physics, the chiral model, introduced by Feza Gürsey in 1960, is a phenomenological model describing effective interactions of mesons in the chiral limit (where the masses of the quarks go to zero), but without necessarily mentioning quarks at all. It is a nonlinear sigma model with the principal homogeneous space of the Lie group SU(''N'') as its target manifold, where ''N'' is the number of quark flavors. The Riemannian metric of the target manifold is given by a positive constant multiplied by the Killing form acting upon the Maurer–Cartan form of SU(''N'').

The internal global symmetry of this model is SU(''N'')_''L'' × SU(''N'')_''R'', the left and right copies, respectively; where the left copy acts as the left action upon the target space, and the right copy acts as the right action. The left copy represents flavor rotations among the left-handed quarks, while the right copy describes rotations among the right-handed quarks, while these, L and R, are completely independent of each other. The axial pieces of these symmetries are spontaneously broken so that the corresponding scalar fields are the requisite Nambu−Goldstone bosons.

This model admits topological solitons called skyrmions.

Departures from exact chiral symmetry are dealt with in chiral perturbation theory.

An outline of the original, 2-flavor model

The chiral model of Gürsey (1960; also see Gell-Mann and Lévy) is now appreciated to be an effective theory of QCD with two light quarks, ''u'', and ''d''. The QCD Lagrangian is approximately invariant under independent global flavor rotations of the left- and right-handed quark fields,
:$\begin q_L \mapsto q_L'= L q_L = \exp q_L \\ q_R \mapsto q_R'= R q_R = \exp q_R \end$
where ''τ'' denote the Pauli matrices in the flavor space and ''θ''_''L'', ''θ''_''R'' are the corresponding rotation angles.

The corresponding symmetry group $\text(2)_L\times\text(2)_R$ is the chiral group, controlled by the six conserved currents
:$L_\mu^i = \bar q_L \gamma_\mu \tfrac q_L, \qquad R_\mu^i = \bar q_R \gamma_\mu \tfrac q_R,$
which can equally well be expressed in terms of the vector and axial-vector currents
:$V_\mu^i = L_\mu^i + R_\mu^i, \qquad A_\mu^i = R_\mu^i - L_\mu^i.$

The corresponding conserved charges generate the algebra of the chiral group,
: \left Q_^i, Q_^j \right= i \epsilon^ Q_I^k \qquad \qquad \left Q_^i, Q_^j \right= 0,
with ''I=L,R'', or, equivalently,
: \left Q_^i, Q_^j \right= i \epsilon^ Q_V^k, \qquad \left Q_^i, Q_^j \right= i \epsilon^ Q_V^k, \qquad \left Q_^i, Q_^j \right= i \epsilon^ Q_A^k.

Application of these commutation relations to hadronic reactions dominated current algebra calculations in the early seventies of the last century.

At the level of hadrons, pseudoscalar mesons, the ambit of the chiral model, the chiral $\text(2)_L \times \text(2)_R$ group is spontaneously broken down to $\text(2)_V$, by the QCD vacuum. That is, it is realized ''nonlinearly'', in the Nambu–Goldstone mode: The ''Q_V'' annihilate the vacuum, but the ''Q_A'' do not! This is visualized nicely through a geometrical argument based on the fact that the Lie algebra of $\text(2)_L\times\text(2)_R$ is isomorphic to that of SO(4). The unbroken subgroup, realized in the linear Wigner–Weyl mode, is $\text(3) \subset \text(4)$ which is locally isomorphic to SU(2) (V: isospin).

To construct a non-linear realization of SO(4), the representation describing four-dimensional rotations of a vector
:$\begin \\ \sigma \end \equiv \begin \pi_1 \\ \pi_2 \\ \pi_3 \\ \sigma \end,$
for an infinitesimal rotation parametrized by six angles
:$\left \, \qquad i =1, 2, 3,$
is given by
: \begin \\ \sigma \end \stackrel \begin \\ \sigma' \end = \left \mathbf_4+ \sum_^3 \theta_i^V V_i + \sum_^3 \theta_i^A A_i \right\begin \\ \sigma \end
where
:$\sum_^3 \theta_i^V V_i =\begin 0 & -\theta^V_3 & \theta^V_2 & 0 \\ \theta^V_3 & 0 & -\theta_1^V & 0 \\ -\theta^V_2 & \theta_1^V & 0 & 0 \\ 0 & 0 & 0 & 0 \end \qquad \qquad \sum_^3 \theta_i^A A_i = \begin 0 & 0 & 0 & \theta^A_1 \\ 0 & 0 & 0 & \theta^A_2 \\ 0 & 0 & 0 & \theta^A_3 \\ -\theta^A_1 & -\theta_2^A & -\theta_3^A & 0 \end.$

The four real quantities define the smallest nontrivial chiral multiplet and represent the field content of the linear sigma model.

To switch from the above linear realization of SO(4) to the nonlinear one, we observe that, in fact, only three of the four components of are independent with respect to four-dimensional rotations. These three independent components
correspond to coordinates on a hypersphere ''S''³, where and are subjected to the constraint
:$^2 + \sigma^2 = F^2,$
with ''F'' a ( pion decay) constant of dimension mass.

Utilizing this to eliminate yields the following transformation properties of under SO(4),
:$\begin \theta^V: \boldsymbol \mapsto \boldsymbol'= \boldsymbol + \boldsymbol^V \times \boldsymbol \\ \theta^A: \boldsymbol \mapsto \boldsymbol'= \boldsymbol + \boldsymbol^A \sqrt \end \qquad \boldsymbol^ \equiv \left \, \qquad i =1, 2, 3.$

The nonlinear terms (shifting ) on the right-hand side of the second equation underlie the nonlinear realization of SO(4). The chiral group $\text(2)_L \times \text(2)_R \simeq \text(4)$ is realized nonlinearly on the triplet of pions— which, however, still transform linearly under isospin $\text(2)_V \simeq \text(3)$ rotations parametrized through the angles $\.$ By contrast, the $\$ represent the nonlinear "shifts" (spontaneous breaking).

Through the spinor map, these four-dimensional rotations of can also be conveniently written using 2×2 matrix notation by introducing the unitary matrix
:$U = \frac \left( \sigma \mathbf_2 + i \boldsymbol \cdot \boldsymbol \right),$
and requiring the transformation properties of ''U'' under chiral rotations to be
:$U \longrightarrow U' = L U R^\dagger,$
where $\theta_L=\theta_V- \theta_A, \theta_R= \theta_V+ \theta_A.$

The transition to the nonlinear realization follows,
:$U = \frac \left( \sqrt \mathbf_2 + i \boldsymbol \cdot \boldsymbol \right) , \qquad \mathcal_\pi^ = \frac \langle \partial_\mu U \partial^\mu U^\dagger \rangle,$
where $\langle \ldots \rangle$ denotes the trace in the flavor space. This is a non-linear sigma model.

Terms involving $\textstyle \partial_\mu \partial^\mu U$ or $\textstyle \partial_\mu \partial^\mu U^\dagger$ are not independent and can be brought to this form through partial integration.
The constant ''F''²/4 is chosen in such a way that the Lagrangian matches the usual free term for massless scalar fields when written in terms of the pions,
:$\mathcal_\pi^ = \frac \partial_\mu \boldsymbol \cdot \partial^\mu \boldsymbol + \frac\left( \partial_\mu \boldsymbol \cdot \boldsymbol \right)^2 + \mathcal ( \pi^6 ).$

Alternate Parametrization

An alternative, equivalent (Gürsey, 1960), parameterization
:$\boldsymbol\mapsto \boldsymbol~ \frac,$
yields a simpler expression for ''U'',
:$U=\mathbf \cos , \pi/F, + i \widehat\cdot \boldsymbol \sin , \pi/F, =e^.$

Note the reparameterized transform under
:$L U R^\dagger=\exp(i\boldsymbol_A\cdot \boldsymbol/2 -i\boldsymbol_V\cdot \boldsymbol/2 ) \exp(i\boldsymbol\cdot \boldsymbol/F ) \exp(i\boldsymbol_A\cdot \boldsymbol/2 +i\boldsymbol_V\cdot \boldsymbol/2 )$
so, then, manifestly identically to the above under isorotations, ; and similarly to the above, as
:$\boldsymbol \longrightarrow \boldsymbol +\boldsymbol_A F+ \cdots =\boldsymbol +\boldsymbol_A F ( , \pi/F, \cot , \pi/F, )$
under the broken symmetries, , the shifts. This simpler expression generalizes readily (Cronin, 1967) to light quarks, so $\textstyle \text(N)_L \times \text(N)_R/\text(N)_V.$

References

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*Georgi, H. (1984, 2009). ''Weak Interactions and Modern Particle Theory'' (Dover Books on Physics)
online
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{{Quantum field theories

Quantum field theory
Quantum chromodynamics
Nuclear physics