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Section6: Electromagnetic Radiation ( ) Section6: Electromagnetic Radiation Potential formulation of Maxwell equations Now we consider a general solution of Maxwell’s equations. Namely we are interested how the sources (charges and currents) generate electric and magnetic fields. For simplicity we restrict our considerations to the vacuum. In this case Maxwell’s equations have the form: ρ ∇ ⋅E = (6.1) ε 0 ∇ ⋅B = 0 (6.2) ∂B ∇×E =− (6.3) ∂t ∂E1 ∂ E ∇×=Bµ J + µε = µ J + (6.4) 0 00∂t 0 c2 ∂ t Maxwell’s equations consist of a set of coupled first-order partial differential equations relating the various components of electric and magnetic fields. They can be solved as they stand in simple situations. But it is often convenient to introduce potentials, obtaining a smaller number of second-order equations, while satisfying some of Maxwell’s equations identically. We are already familiar with this concept in electrostatics and magnetostatics, where we used the scalar potential Φ and the vector potential A. Since ∇ ⋅B = 0 still holds, we can define B in terms of a vector potential: B= ∇× A (6.5) Then the Faraday's law (6.3) can be written ∂A ∇×E + = 0 . (6.6) ∂t This means that the quantity with vanishing curl in (6.6) can be written as the gradient of some scalar function, namely, a scalar potential Φ . ∂A E + = −∇Φ (6.7) ∂t or ∂A E = −∇Φ − . (6.8) ∂t The definition of B and E in terms of the potentials A and Φ according to (6.5) and (6.8) satisfies identically the two homogeneous Maxwell equations (6.2) and (6.3). The dynamic behavior of A and Φ will be determined by the two inhomogeneous equations (6.1) and (6.4). Putting Eq. (6.8) into (6.1) we find that ∂ ρ ∇Φ+2 () ∇⋅A =− . (6.9) ∂ ε t 0 This equation replaces Poisson equation (to which it reduces in the static case). Substituting Eqs. (6.5) and (6.8) into (6.4) yields ∂Φ ∂ 2 ∇×∇×() =µ −1 ∇ − 1 A A0 J . (6.10) c2∂ t ct 2 ∂ 2 1 Now using the vector identity ∇×∇×( A) =∇∇⋅( A) −∇ 2 A , and rearranging the terms we find 1∂Φ 1 ∂ 2A ∇∇⋅() −∇2 =µ − ∇ − A A0 J , (6.11) c2∂ t ct 2 ∂ 2 or ∂2 ∂Φ ∇−2 1A −∇∇⋅+ 1 =− µ A A 0 J . (6.12) ct2∂ 2 ct 2 ∂ We have now reduced the set of four Maxwell equations to two equations (6.9) and (6.12). But they are still coupled equations. The uncoupling can be accomplished by exploiting the arbitrariness involved in the definition of the potentials. Since B is defined through (6.5) in terms of A, the vector potential is arbitrary to the extent that the gradient of some scalar function A can be added. Thus B is left unchanged by the transformation, A→ A′ = A +∇Λ . (6.13) For the electric field (6.8) to be unchanged as well, the scalar potential must be simultaneously transformed, ∂Λ Φ→Φ′ =Φ− . (6.14) ∂t The freedom implied by (6.13) and (6.14) means that we can choose a set of potentials ( A, Φ ) such that 1 ∂Φ ∇⋅+A = 0 . (6.15) c2 ∂ t This will uncouple the pair of equations (6.9) and (6.12) and leave two inhomogeneous wave equations, one for Φ and one for A: 1 ∂2 Φ ρ ∇Φ−2 =− . (6.16) 2∂ 2 ε c t 0 1 ∂2A ∇−2A =− µ J . (6.17) c2∂ t 2 0 Equations (6.16) and (6.17), plus (6.15), form a set of equations equivalent in all respects to Maxwell’s equations. Gauge Transformations The transformation (6.13) and (6.14) is called a gauge transformation , and the invariance of the fields under such transformations is called gauge invariance. The relation (6.15) between A and Φ is called the Lorentz condition . To see that potentials can always be found to satisfy the Lorentz condition, suppose that the potentials A and Φ that satisfy (6.9) and (6.12) do not satisfy (6.15). Then let us make a gauge transformation to potentials A′ and Φ′ and demand that A′ and Φ′ satisfy the Lorentz condition: 1∂Φ′ 1 ∂Φ 1 ∂2 Λ ∇⋅A′ + ==∇⋅+∇Λ+0 A 2 − . (6.18) ct2 ∂ ctct2 ∂ 2 ∂ 2 Thus, provided a gauge function Λ can be found to satisfy ∂2 Λ ∂Φ 2 1 1 ∇Λ− =−∇⋅A + , (6.19) ct2∂ 2 ct 2 ∂ the new potentials A′ and Φ′ will satisfy the Lorentz condition and wave equations (6.16) and (6.17). 2 Even for potentials that satisfy the Lorentz condition (6.15) there is arbitrariness. Evidently the restricted gauge transformation , A→ A +∇Λ , (6.20) ∂Λ Φ → Φ − , (6.21) ∂t where 1 ∂2 Λ ∇Λ−2 = 0 (6.22) c2∂ t 2 preserves the Lorentz condition, provided A and Φ satisfy it initially. All potentials in this restricted class are said to belong to the Lorentz gauge. The Lorentz gauge is commonly used, first because it leads to the wave equations (6.9) and (6.12), which treat A and Φ on equivalent footings. Another useful gauge for the potentials is the so-called Coulomb gauge. This is the gauge in which ∇ ⋅A = 0 . (6.23) From (6.16) we see that the scalar potential satisfies the Poisson equation, 1 ∂2 Φ ρ ∇Φ−2 =− (6.24) 2∂ 2 ε c t 0 with solution, 1ρ (,)r′ t Φ(r ,t ) = d3 r ′ . (6.25) πε ∫ − ′ 4 0 r r The scalar potential is just the instantaneous Coulomb potential due to the charge density ρ(r ,t ) . This is the origin of the name “Coulomb gauge ”. There is a peculiar thing about the scalar potential (6.25) in the Coulomb gauge: it is determined by the distribution of charge right now. That sounds particularly odd in the light of special relativity, which allows no message to travel faster than the speed of light. The point is that Φ by itself is not a physically measurable quantity – all we can measure is E, and that involves A as well. Somehow it is built into the vector potential, in the Coulomb gauge, that whereas Φ instantaneously ∂A reflects all changes in ρ , the combination −∇Φ − does not; E will change only after sufficient time ∂t has elapsed for the “news” to arrive. The advantage of the Coulomb gauge is that the scalar potential is particularly simple to calculate; the disadvantage is that A is particularly difficult to calculate. The differential equation for A (6.12) in the Coulomb gauge reads ∂2 ∂Φ ∇−2 1A =−µ +∇ 1 A0 J . (6.26) ct2∂ 2 c 2 ∂ t Retarded potentials In the Lorentz gauge A and Φ satisfy the inhomogeneous wave equation , with a “source” term (in place of zero) on the right: 1 ∂2 Φ ρ ∇Φ−2 =− , (6.27) 2∂ 2 ε c t 0 3 1 ∂2A ∇−2A =− µ J . (6.28) c2∂ t 2 0 From now we will use the Lorentz gauge exclusively, and the whole of electrodynamics reduces to the problem of solving the inhomogeneous wave equations for specified sources. In the static case Eqs. (6.27), (6.28) reduce to Poison’s equation ρ ∇2 Φ = − , (6.29) ε 0 ∇2 = − µ A0 J , (6.30) with familiar solutions 1ρ (r′ ) Φ(r ) = d3 r ′, (6.31) πε ∫ − ′ 4 0 r r µ J( r ′ ) A( r ) = 0 d3 r ′ . (6.32) 4π ∫ r− r ′ In dynamic case, electromagnetic “news” travel at the speed of light. In the nonstatic case, therefore, it’s not the status of the source right now that matters, but rather its condition at some earlier time tr (called the retarded time ) when the “message” left. Since this message must travel a distance r− r ′ , the delay is r− r ′ / c : r− r ′ t= t − . (6.33) r c The natural generalization of Eqs. (6.31), (6.32) to the nonstatic case is therefore 1ρ (,)r′ t Φ(r ,t ) = r d3 r ′ , (6.34) πε ∫ − ′ 4 0 r r µ J( r ′ ,t ) A( r ,t ) = 0 r d3 r ′ . (6.35) 4π ∫ r− r ′ ρ ′ ′ ′ Here (r ,tr ) and J( r ,tr ) is the charge and current density that prevailed at point r at the retarded time tr. Because the integrands are evaluated at the retarded time, these are called retarded potentials . Note that the retarded potentials reduce properly to Eqs. (6.31), (6.32) in the static case, for which ρ and J are independent of time. That all sounds reasonable – and surprisingly simple. But so far we did not prove that all this is correct. To prove this, we must show that the potentials in the form (6.34), (6.35) satisfy the inhomogeneous wave equations (6.27), (6.28) and meet the Lorentz condition (6.15). In calculating the Laplacian we have to take into account that the integrands in Eqs. (6.34), (6.35) depend on r in two places: explicitly, in the − ′ = − − ′ denominator r r , and implicitly, through tr tr r / c . Thus, 1 1 1 ∇Φ(,)rt = ∇ρ (,)(,) r′ tt + ρ r ′ ∇ dr3 ′ , (6.36) πε ∫ −′r r − ′ 4 0 rr rr and ∂ρρρ ∂∂∂t 1 ∂ ρ 1 ∂− ρ r r ′ ∇ρ(rxx′ ,t ) =ˆ = ˆ r =∇=− t ∇−=− rr ′ . (6.37) ri∑∂ ∑ i ∂∂∂ r ∂ ∂− ′ ixi i txt rir ct r ct r r r 4 Taking into account that 1 r− r ′ ∇ = − , (6.38) r− r ′ r− r ′ 3 we find 1 1 ∂ρ rr −′ rr − ′ ∇Φ(,)rt = − − ρ (,) r ′ tdr 3 ′ , (6.39) πε ∫ ∂ −′2r − ′ 3 4 0 c t r rr rr Taking the divergence we obtain: 1 1 rr−∂∂−′ρ ρ rr ′′ rr − rr − ′ ∇Φ2 = − ⋅∇ +∇⋅ − ⋅∇+∇⋅ρ ρ 3 ′ (r ,t ) ∫ 2 2 3 3 d r (6.40) 4πε c−′∂ t ∂ t −− ′′ − ′ 0 rrr r rr rr rr Similar to (6.37): ∂∂ρ 2 ρ1 ∂− 2 ρ r r ′ ∇ = ∇=−t .
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