Lectures April 29, May

Transcription

1 Lectures April 29, May Electromagnetism controls most of physics from the atomic to the planetary scale, we have spent nearly a year exploring the concrete consequences of Maxwell s equations the Lorentz force. But electromagnetism holds one clue to set our imagination free to go further. In the gauge invariance of the vector potential, it points to the elegant structure of LOCAL SYMMETRY a generalization of a global symmetry of electromagnetism which is not obvious classically.

2 Topics we need to introduce We first need to discuss several topics from outside the realm of classical electromagnetism: Lattice approach to field theories Internal field variables How to make global local This will give us minimal substitution covariant derivatives electromagnetic field strengths as a form of curvature non-abelian gauge theories (QCD, GSW, stard model)

3 Lattice approach to field theories Field theories may be approached as the continuum limit of discrete degrees of freedom on a lattice. Matter fields have their degrees of freedom defined only at sites x = (an x, an y, an z ), with integers n i lattice spacing a. For relativistic treatments we will also discretize time. What are the degrees of freedom φ j at each site j? φ j lives in some domain: reals? complex numbers? a finite set (like for spin 1/2 Ising model)? a real vector?

4 Spins on a Lattice Consider having at each site a spin S j of fixed length but of variable direction, with an interaction H = J S i S j nn between nearest-neighbor lattice sites i j. Without any coupling to spatial things (such as ( r i r j ) ( S j S i ), spins are in separate space, not necessarily 3D. Invariant under rotations by an orthogonal matrix R, S j RS j, each term S i S j in the Hamiltonian in unchanged. This holds also in the continuum limit, with S( x) a field taking independent (though not uncorrelated) values at each spatial point. Here the nearest-neighbor coupling becomes µα µs α µ S α. That keeps our space real space uncoupled. If the space is R 3, then we might also have ( S) 2 terms, then space spin would be coupled with an L S term.

5 If there is no coupling of the spins with space, the symmetry under rotation in spin space is called an symmetry. ( ) ψp Isotopic spin. Ψ :=. Strong interactions are ψ n invariant under SU(2) transformations ( I ) Ψ(x µ ) e i ω I Ψ(x µ ), i are pauli spin matrices σ i acting on the doublet Ψ. The isospin rotation has to be the same at all points in spacetime for this to be a symmetry. Thus it is a global symmetry.

6 In Mechanics, p i E i / t. Then a free nonrelativistic with E = p 2 /2m is described by a wave function i ψ t = 2 2m 2 ψ which is Schrödinger s equation for a free. Notice ψ must be complex. If ψ(x µ ) is a solution, so is e iλ ψ(x µ ), as long as λ is a constant. symmetry under phase change. Ms. Noether tells us: conserved current. For our non-relativistic : J = ρ v i qψ ψ.

7 Conserved current More generally, the phase induces φ = iλφ, the current J µ = L µ φ φ is just like the kinetic energy part of the lagrangian density with one derivative left out. For example, the Dirac L = i ψγ µ µ ψ m ψψ gives a conserved current J µ = ψγ µ ψ, while for the Klein-Gordon lagrangian for scalar field, L = 2 ( µ φ)( µ φ ) m 2 φ φ gives J µ = i 2m (φ µ φ φ µ φ ). This conserved current is a consequence of the global symmetry.

8 Minimal Substitution If the is charged in the presence of an external field A µ (x ν ), this interaction can be incorporated by minimal substitution into the free lagrangian, which is to say that p p qa/c, E E qφ, so for a non-relativistic i ψ t = (qφ 2 ( ) ) 2 + iqa/ c ψ. 2m qφ is the potential energy. Might recognize the A term from Schnetzer s derivation of Hamiltonian. Important point: Not invariant under A µ A µ µ Λ. Only effect is on derivative adds a piece iq µ Λ/c to each derivative operator, or µ e iqλ/c µ e iqλ/c. So if ψ satisfies the equation in the original gauge, ψ = e iqλ/c ψ saitisfies the equation in the transformed gauge. So we have a symmetry under a more involved gauge transformation.

9 Gauge invariance with matter Conclusion: A µ A µ + µ Λ (Φ Φ 1 c ψ e iqλ/c ψ Λ t, A A + Λ) is an invariance of the theory, is the correct form of a gauge transformation. Here we have the required local phase change, or rotation, as a result of deming gauge invariance, but we now ask the reverse can we turn a global symmetry of the matter fields into a local symmetry by having gauge fields with a gauge invariance.

10 Let us develop the idea of a locally invariant gauge theory more formally, starting with matter fields with a symmetry group. Let the matter fields be a set of N real fields φ i (x µ ) with an symmetry group G which acts with representation M, that is, G G acts on the φ fields by G : φ i (x) φ i(x) = j M ij (G)φ j (x). (1) To be a symmetry the must be invariant. The usual field theory kinetic term is 1 2 µ,i ( µφ i ) ( µ φ i ), invariant if M is a constant orthogonal matrix k M kim kj = δ ij. Any potential term of the forms V ( i φ2 i ) is also invariant. Individual components of φ have no intrinsic meaning, as M mixes them up.

11 V might not be invariant under all of O(N) but only under a subgroup. Example: SU(3) of color acts on triplet of complex quark fields, 6 real components, kinetic energy is invariant under O(6) SU(4), but full lagrangian is invariant only under an SU(3) subgroup. So we have a symmetry group G which has generators L b which form a basis of the Lie algebra G. There are as many independent L b s as the dimension of the group or algebra. For example, 3 for rotations in 3D or SU(2), 8 for SU(3), N(N 1)/2 for SO(N), rotations in N dimensions.

12 On our lattice, the degrees of freedom φ i (x) are replaced by φ i ( n) for n Z 4 with x µ = an µ. The mass term m2 d 4 x φ 2 i (x) m2 a 4 φ 2 i ( n). 2 2 i i n Z 4 Similarly for any other potential (single site) term. Kinetic energy: replace derivative by finite difference. The simplest substitution is to replace µ φ i (x) 1 ( φ i ( n + a ) µ ) φ i ( n), where µ is 1 in the µ direction 0 in the others. Note the best x ν = an ν aδν µ, the center of the link between the two sites. ( φ) 2 gives mass-like terms nearest-neighbor terms i φ i( n + µ )φ i ( n). Each is invariant under rotations if the same rotation is used at all sites, so we have a global symmetry.

13 In a relativistic field theory, all information is local. Why should symmetry at one site know about a distant one? Could we have a local symmetry? Mass terms potential terms, which depend only on a single site, are okay because they don t know the difference between local global. But the nearest neighbor term φ i ( n + µ )φ i ( n) i M ik (G( n + µ ))M ij (G( n)) φ k ( n + µ )φ j ( n) ( = M 1 (G( n + ) µ ))M(G( n)) φ k( n + µ )φ j ( n) kj is not invariant because M 1 (G( n + µ ))M(G( n)) 1 if the G s ( hence the M s) vary from point to point.

14 The problem, of course, is that we need a measure of how much φ changes from point to point which does not depend on the arbitrary basis vectors used at each point to describe φ in terms of components. We need a method of looking at φ( x) as if it were at x, so we can subtract it from φ( x ) to find the change. That is, we need a definition of parallel transport moving the object (say a vector) φ from one point to another without rotating or otherwise distorting it. We considered this problem in Lecture 4 when we considered the gradient operator on vectors in curvilinear coordinates. Let s recall what we did then.

15 Rule for parallel transport As an example, it might help to think of an ordinary vector in the plane, expressed in polar coordinates. Consider the unit basis vectors e r e θ at the point P. If we transport e r to the point P while keeping it parallel to what it was, we arrive at the vector labelled e Ge r, which is not the same as the r unit radial vector e r at the point G er P. P G Note that if we have a vector e V = V r e r + V θ e θ at P which θ e is unchanged (parallel transported) P r from the vector V = V r e r + V θ e θ at P, we do not have V r = V r. Here we had an a priori rule for parallel transport. But if we allow gauge invariance, the rule becomes a dynamical variable, a new degree of freedom.

16 For each pair of nearest neighbors, that is, for each link, we need a group element which specifies how to parallel-transport the basis vectors. Thus we have a lot of new degrees of freedom, but we do have a local symmetry. This gives us a gauge field theory. The matter fields in a gauge theory transform according to some representation of the group, but the gauge fields, which define the parallel transport, take values in the group itself, or in the Lie algebra formed from the generators of the group. So for example, for colored quarks, the quarks transform as a three-complex-dimension representation of the SU(3) group, but there are 8 gluons, because there are 8 generators of SU(3).

17 We have seen that each link will have a group element G which defines parallel transport, so that transporting φ i ( n) to the neighboring site n + µ gives j M ij(g)φ j ( n), so the change in φ which might enter our lagrangian is φ i = φ i ( n + µ ) j M ij (G)φ j ( n). In the continuum limit the fields change little from one site to the next, so φ i ( n + µ ) φ i ( n) + a µ φ i. parallel transport should not require more than a small rotations, proportional to the lattice spacing, so we can write 1 G = e iaga, M(G) = M(e iaga ) 1 + iagm(a). 1 The scale of generators A is somewhat arbitrary we include a factor of g, which will be called the fundamental charge, here, but some people do not.

18 Here the generator A G, the Lie algebra generating G. So to o(a 1 ), φ i = a µ φ i ig j M ij (A)φ j. In the continuum limit we define 1/a times this to be the covariant derivative. We need a covariant derivative for each link, of which there are four emerging from each site. So we have a 4-vector of Lie-algebra valued fields A µ, which have Lie-algebra components multiplying the basis generators L b, so A µ (x ν ) = b A(b) µ (x ν )L b. Then the covariant derivative, acting on the matter field, is (D µ φ) j = µ φ j ig A (b) µ M jk (L b )φ k kb D µ φ = µ φ iga (b) µ M(L b )φ.

19 Gauge Transformations On the lattice the terms we were just discussing are φ( n + µ ) M(G L ) φ( n), where G L is the group transformation associated with the link ( n, n + µ ) that implements a parallel transport. What happens to G L under local group transformations? Think of gauge transformations passively that is, φ( n) doesn t change but its description in terms of components φ i ( n) does. transport across a link doesn t change either, but its group element does. If G L transports φ p to φ q at site q with M(G L ) φ(x p ) in our original set of basis vectors, we do a gauge transformation φ p φ p = M(G p ) φ p, φ q φ q = M(G q ) φ q, the parallel transport in the new basis is φ q = M(G q ) M(G L ) M(G p ) 1 φ p = M(G q G L G 1 p ) φ p, so the gauge field which does parallel transport p q is

20 Gauge transform Now our gauge transform acts on both matter fields gauge fields: φ(x p ) M(G p ) φ(x p ) Λ : φ(x q ) M(G q ) φ(x q ) G L G q G L G 1 p This gauge transformation is a local symmetry of the gauge field theory. Let s verify that this is an invariance of the nearest neighbor term: φ(x q ) M(G L ) φ(x p ) = φ i (x q )M ij (G L )φ j (x p ) M ik (G q )φ k (x q )M ij (G q G L G 1 p )M jl (G p )φ l (x p ) = φ k (x q )M 1 ki (G q)m ij (G q G L G 1 p )M jl (G p )φ l (x p ) = φ k (x q )M kl (G L )φ l (x p ) = φ(x q ) M(G L ) φ(x p ), where we have used the orthogonality of M(G q ) the fact that the M s are a representation, therefore M 1 ki (G q)m ij (G q G L G 1 p )M jl (G p ) = M kl (G L ).

21 In the Continuum In the continuum limit, G L = e iaga L, M(G L ) = M(e iaga L ) 1 + iagm(a L ). The gauge transform group element Λ is not necessarily a small change, but we can still write it as the exponential of a generator, Λ(x µ ) = e iλ(xµ) with λ(x µ ) a Lie algebra element, λ(x µ ) = b λ(b) (x µ )L b. We may assume that λ (b), though not small, changes slowly (o(a)) across a link. So if the gauge transformation takes φ(x) φ (x) A (b) µ (x) A µ (b) (x), we have φ (x) = e i P b λ(b) (x)m(l b ) φ(x), (b) iaga µ (x) e = e iλ(x+ 1 2 a µ) e iaga(b) µ (x)l b e iλ(x 1 2 a µ). We placed x µ in the middle of a link.

22 To first order in a In exping in powers of a, we need care that λ(x), µ λ(x), A µ do not commute, so we exp Λ rather than λ, e iagaµ 1 + iaga µ, e iλ(x± 1 2 a µ) e iλ(x) ± 1 2 a µ [ e iλ(x)], plugging these in, we get 1+iagA µ = (e iλ + 12 ) a µe iλ (1+iagA µ ) (e iλ 12 ) a µe iλ = 1 + iage iλ A µ e iλ a ( µ e iλ) e iλ 1 2 aeiλ ( µ e iλ) Note from µ ( e iλ e iλ) = 0 that the third fourth terms are equal, so we can drop the third double the fourth.

23 Gauge transform of Gauge Field This gives Electromagnetism? A µ = e iλ A µ e iλ + i g eiλ µ e iλ = e iλ (A µ + i g µ ) e iλ Gauge transformations for charged wavefunctions were just phase shifts, rotations in two dimensions. There is only one generator L 1 with G = e iλl 1, so everything commutes. The gauge transformation takes ( ) ( ) ( ) φ φ Re Φ cos λ sin λ Re Φ = Im Φ = sin λ cos λ Im Φ or Φ = e iλ Φ, A µ = e (A iλ µ + i ) g µ e iλ = A µ + g 1 µ λ.

24 Gauge transform in E&M So for electromagnetism, Φ = e iλ Φ A µ = A µ + g 1 µ λ. is what we are used to. But this simplicity only holds for an Abelian group, where all generators commute. More generally, we need φ (x ν ) = e im(λ(xν )) φ(x ν ) A µ(x ν ) = e iλ(xν) A µ (x ν )e iλ(xν) + i g eiλ(xν) µ e iλ(xν )

25 Terms in L We have learned how to formulate the interaction of matter fields with the gauge fields, both on the lattice in the continuum. But what about the pure gauge field part of L? Can t depend on just one link a be gauge invariant, because choosing a gauge transform G 1 = G a makes G a = 1I, nothing to depend on. Simplest way to get rid of the gauge dependance of G a = e iagax(xa) on G 2 = e iλ(x 2) is to premultiply it by G b, G b G a G 3 G b G 1 2 G 2G a G 1 1 = G 3 G b G a G 1 1. Independent of G 2, but... 1 a 2 G G G G a 3 b 2 a 1 a 2 G G G G G G b a 3 b a

26 There is still a gauge dependence on the endpoints of the path, however, so the best thing to do is close the path. To do so, we are traversing some links backwards from the way they were defined, but from that definition in terms of parallel transport it is clear that the group element associated with taking a link back- d b d b 4 c 3 4 c 3 wards is the inverse of the element taken going forwards. So 1 a 2 1 a 2 the group element associated with the closed path on the right (which is called a plaquette) is G P = G 1 d G 1 c G b G a, which transforms under gauge transformations as ) 1 ( G3 G c G 1 ) 1 G P G P = ( G 4 G d G ( G3 G b G 1 ) ( 2 G2 G a G 1 ) 1 = G 1 G 1 d G 1 4 G 4G 1 c G 1 3 G 3G b G 1 2 G 2G a G 1 1 = G 1 G 1 d G 1 c G b G a G 1 1 = G 1 G P G 1 1. So the plaquette group element is not invariant.

27 Invariant function on placquette G P G 1 G p G 1 1 is not invariant, but is simpler. In the continuum limit each G L is small transformation, G c differs from G a proportional to the lattice spacing, so G P close to the identity, G P 1I a generator in the Lie algebra. The Killing form acting on that generator will provide us with an invariant. Let us define F µν = ia 2 g 1 (G P 1) to be the field-strength tensor, where µ ν are the directions of links a b respectively. Let us take x in the center of the placquette. d 4 c 3 1 x a 2 µ b ν

28 Exping each link to order O(a 2 ) G a 1+iagA µ(x 1 2 a ν) 1 2 a2 g 2 A 2 µ(x 1 2 a ν) 1+iagA µ(x) 1 2 ia2 g νa µ(x) 1 2 a2 g 2 A 2 µ(x) 4 c 3 a 2 µ G 1 c 1 iaga µ(x+ 1 2 a ν) 1 2 a2 g 2 A 2 µ(x+ 1 2 a ν) d 1 x b ν 1 iaga µ(x) 1 2 ia2 g νa µ(x) 1 2 a2 g 2 A 2 µ(x), so, to order a 2, G P = 1 iaga ν(x) ia2 g µa ν(x) 1 «2 a2 g 2 A 2 ν(x) 1 iaga µ(x) 1 2 ia2 g νa µ(x) 1 «2 a2 g 2 A 2 µ(x) 1 + iaga ν(x) ia2 g µa ν(x) 1 «2 a2 g 2 A 2 ν(x) 1 + iaga µ(x) 1 2 ia2 g νa µ(x) 1 «2 a2 g 2 A 2 µ(x) = 1 + a 2 g {g [A µ(x), A ν(x)] + i µa ν(x) i νa µ(x)}

29 Evaluation of F µν Thus F µν (x) = µ A ν (x) ν A µ (x) ig [A µ (x), A ν (x)]. Note that F µν is a Lie-algebra valued field, F µν (x) = b F (b) µν (x)l b. An antisymmetric tensor, F µν (x) = F νµ (x). Because the Lie algebra is defined in terms of the structure constants, cab d by [L a, L b ] = ic d ab L d, the field-strength tensor may also be written F µν (d) = µ A (d) ν ν A (d) µ + gcab d A(a) µ A (b) ν.

30 Before we turn to the, let me point out a crucial relationship between the covariant derivatives the field-strength. If we take the commutator of covariant derivatives D µ = µ iga (b) µ L b at the same point but in different directions, [D µ, D ν ] = [ µ iga µ, ν iga ν ] = ig µ A ν g 2 A µ A ν (µ ν) = g 2 [A µ, A ν ] ig µ A ν + ig ν A µ = igf µν. Notice that although the covariant derivative is in part a differential operator, the commutator has neither first or second derivatives left over to act on whatever appears to the right. It does need to be interpreted, however, as specifying a representation matrix that will act on whatever is to the right.

31 The Density for F We have seen that the field strength F µν is an element of the Lie algebra which transforms by conjugation, F µν GF µν G 1 under gauge transformation G at one vertex. For the we need an invariant quadratic function on the algebra fortunately that is just the Killing form, β(l a, L b ) = 2δ ab if we have normalized our generators L a in the usual way. This is familiar in the form L 2 is invariant for the rotations. Thus we may take L = 1 F (b) (b) µν µν F 4 b to be the pure gauge term. Obviously that agrees with our E&M lagrangian, where there is only one term in the sum over b.

32 of Almost Everything All together, if we have a non-abelian gauge (Yang-Mills) field interacting with Dirac s that transform under a representation t b (that is, t b = D(L b ) in the representation for the fermions), together with some scalar s φ transforming under a possibly different representation t, the langrangian density is L = 1 4 F µν (b) F (b) µν + i ψγ ( µ µ iga (b) µ t b) ψ + 1 [( µ iga (b) µ 2 t b) ] T [( φ µ iga (b) µ t b) ] φ. Now you know everything about the non-abelian theories that run the universe (except for gravity).