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The Aharonov-Bohm effect is the phase $q\Phi /\hslash$ acquired when a charge $q$ encircles magnetic flux $\Phi$, even when the magnetic field vanishes along the charge's path.

In the magnetic Aharonov-Bohm effect, a coherent electron wave takes two routes around inaccessible magnetic flux. Recombination converts their relative wave phase into displaced interference fringes. In 1959, Yakir Aharonov and David Bohm proposed a shielded solenoid geometry and identified the obstruction to removing the magnetic vector potential throughout the accessible region.^[ref]

Use International System of Units, electric charge $q$, and magnetic minimal coupling $H=(-i\hslash \nabla -q\mathbf{A}{)}^2/(2m)$. Let ${\gamma }_1,{\gamma }_2$ run from the same source $S$ to the same detector point $D$. The oriented closed curve $C$ follows ${\gamma }_1$ and then ${\gamma }_2$ backwards. For a static magnetic vector potential, the magnetic contribution to the relative wave phase isHere ${\Phi }_C={\int }_{\Sigma }\mathbf{B}\cdot d\mathbf{S}$ is the oriented magnetic flux through a spanning surface. Equal or calibrated nonmagnetic propagation phases are assumed. For an electron, $q=-e$, where $e>0$ is the elementary charge; reversing $C$ reverses the sign. The observable is the gauge-invariant holonomy $e^{i\delta }$, rather than a unique local value of the magnetic vector potential.

The accessible region is a multiply connected region: $C$ cannot contract to a point without entering the excluded flux region. Stokes theorem does not set the line integral to zero, because a spanning surface must cross that excluded region, where the magnetic field need not vanish. There is real magnetic flux in the apparatus; only the accessible paths are idealized as free of the magnetic field. The exterior of a finite straight solenoid also permits routes around its ends, so its exact topology and leakage differ from those of an infinite solenoid or an excluded solid torus.

An explicit model makes the obstruction visible. Outside an infinite flux tube of radius $a$, choose cylindrical coordinates $(r,\theta ,z)$ andIts curl vanishes there, but a loop with winding number $n$ has ${\oint }_C\mathbf{A}\cdot d\mathbf{l}=n\Phi$. Locally $\mathbf{A}=\nabla (\Phi \theta /(2\pi ))$; the angular coordinate changes by $2\pi$ on a circuit. For generic $q\Phi /\hslash$, that candidate gauge removal changes the wave function's single-valued boundary condition. On a simply connected region, a smooth static magnetic vector potential with zero curl instead has a globally defined scalar primitive and is a pure gauge potential. Thus local zero magnetic field and globally removable magnetic vector potential are different conditions.

For an original interference example, take mutually coherent scalar amplitudes with constant individual intensities $I_1,I_2$ and an imposed carrier phase $ky$, as in off-axis holography. ThenThe interference fringe spacing is $d=2\pi /k$. An electron enclosing $\Phi =h/(4e)$ has $\delta =-\pi /2$, shifting maxima by $d/4$ for this orientation. Increasing the magnetic flux by $h/e$ changes $\delta$ by $-2\pi$ and restores the same pattern: the electron Aharonov-Bohm effect has magnetic flux quantum $h/e$. This uses $h=2\pi \hslash$, the Planck constant. It does not require the source's magnetic flux to vary continuously during a single exposure.

For $I_1=1$, $I_2=1/4$, the extrema are $I_{\max }=9/4$, $I_{\min }=1/4$, and the interferometric visibility is $4/5$. A fixed Aharonov-Bohm phase translates these fringes without changing their contrast. Averaging unresolved fluctuating phases can reduce the contrast; a darkened pattern alone cannot identify enclosed magnetic flux. These formulas assume coherence, indistinguishable alternatives and stable beam envelopes.

The appeal of the Aharonov-Bohm effect is that a local Lorentz force calculation along either ideal path misses information carried by their global arrangement. It gives a physical example of holonomy and supplies the encircling phase used in Chern--Simons flux attachment. This alone does not establish intrinsic topological order. The static comparison also matters: changing magnetic flux can induce an electric field, so a time-dependent experiment requires the scalar-potential contribution as well as the displayed spatial line integral.

Analytic torsion is a spectral invariant constructed from regularized determinants of the Laplacians on differential forms with suitable flat coefficients. It is the analytic counterpart of combinatorial Reidemeister torsion, whose refinements include Turaev torsion; equality theorems relate the spectral and combinatorial invariants under appropriate hypotheses on the manifold and coefficients.

A four-vertex model proves the essential gauge invariance without solving an electron scattering problem. Use two directed paths $S\to U\to D$ and $S\to L\to D$. Assign an edge phase $a_{ij}$ and a complex unit link $U_{ij}=e^{i{a}_{ij}}$ to each directed edge, with $U_{ji}=U_{ij}^{-1}$. These links represent discretized magnetic line integrals multiplied by $q/\hslash$. Under a gauge transformation with arbitrary vertex phases ${\chi }_i$,Both path products acquire the same factor $e^{i({\chi }_D-{\chi }_S)}$. Their gauge-invariant ratio is the closed-loop holonomyEvery intermediate vertex phase cancels. For any fixed nonnegative real amplitudes $a,b$, the interference quantity $|a{P}_1+b{P}_2{|}^2$ is unchanged too. A gauge-dependent open-path phase consequently cannot be interpreted as an observable by itself.

For an exact example, measure phases in units of $\pi$ and order edges as $(SU,UD,SL,LD)$. Choose edge phases $(0,0,0,1/2)$ and vertex phases $({\chi }_S,{\chi }_U,{\chi }_L,{\chi }_D)=(1/3,-1/4,2/5,1/7)\pi$. The transformed edge phases areThe upper and lower sums become $-4\pi /21$ and $13\pi /42$; their difference remains $-\pi /2$. With amplitudes $a=1$, $b=1/2$, the intensity remains $5/4$. Sweeping one physical loop phase through $2\pi$ instead produces extrema $9/4$ and $1/4$, agreeing with the interferometric visibility example above.

On a connected regular radial interval, trivial holonomy around circular fibres forces the continuous function $f^{\prime }$ to be integer-valued, hence constant. The resulting Riemannian metrics have $f(r)=nr+b$, with $n\in \mathbb{Z}$ and $f>0$ on the interval. Their Gaussian curvature is $-f^{\prime \prime }/f=0$. For $n\ne 0$, $\rho =f/|n|$ and $\theta =|n|\varphi$ give a local isometry to the Euclidean plane; for $n=0$, $r,b\varphi$ give a flat cylindrical chart. The polar developing map is injective on the regular domain when $|n|=1$, and is a multiple covering when $|n|>1$. Trivial holonomy therefore does not imply a global Euclidean chart. The slope is a discrete parameter, not an unrestricted real one.

A cone is intrinsically flat away from its vertex but has nontrivial angular holonomy.

For the Riemannian metric $d{r}^2+f(r{)}^{2}d{\varphi }^2$, where $f>0$ and $\varphi$ has period $2\pi$, parallel transport around $r=r_0$ acts on the orthonormal basis components of a tangent vector by a rotation through $-2\pi f^{\prime }(r_0)$. Indeed, those components obey $u^{\prime }=f^{\prime }(r_0)v$ and $v^{\prime }=-f^{\prime }(r_0)u$ as functions of $\varphi$. Thus all tangent vectors return to themselves exactly when $f^{\prime }(r_0)\in \mathbb{Z}$. The integer criterion concerns holonomy and is stronger than local flatness.

The holonomy of a flat connection is invariant under based homotopy of loops, giving a group homomorphism from the fundamental group into the structure group after choosing a frame at the base point. Changing that frame conjugates the representation. Its associated bundle is constructed from the universal cover and this monodromy. It differs from the natural representation of a Riemannian holonomy group on a tangent fiber.

A foliation of a symplectic manifold is Lagrangian when every leaf has half the ambient dimension and the restricted symplectic form vanishes. The tangent distribution is a real polarization in geometric quantization. Restricted prequantum connections are flat on individual leaves, but nontrivial holonomy can obstruct global parallel sections.

A leaf of a Lagrangian foliation is an immersed Lagrangian submanifold tangent to its integrable distribution. A prequantum line bundle restricts to a flat connection along such a leaf, because its curvature is proportional to the symplectic form. Nonzero global parallel sections require trivial holonomy along every leaf loop.

A parallel spinor field is annihilated by the spinor covariant derivative. Its value at one point determines it on a connected domain by parallel transport. The spinor curvature identity implies $R_{\mu \nu AB}{\gamma }^{AB}ϵ=0$, but this local condition alone need not settle global existence: the spin structure and holonomy also matter. Lorentzian parallel spinors need not force the whole Ricci tensor to vanish, unlike the positive-definite case.

The only nonconstant entry of the metric tensor is $g_{\varphi \varphi }=R^2{\sin }^2\theta$, whose partial derivative is ${\partial }_{\theta }{g}_{\varphi \varphi }=2R^2\sin \theta \cos \theta$. The nonzero Christoffel symbols areAlong a fixed latitude with $0<\theta <\pi$, parallel transport givesSet $X=V^{\theta }$ and $Y=\sin \theta V^{\varphi }$. Since $\theta$ is fixed, this linear ordinary differential equation becomesIt is a rotation at constant rate $-\cos \theta$. Its general solution, also valid at the equator where the rate is zero, isThe initial data give $A=1$ and $B=0$. At $\theta =\pi /3$, a complete circuit has $\varphi \cos \theta =\pi$, soThe physical squared length uses the metric tensor, rather than the Euclidean sum of coordinate components:It is constant and equals $R^2$ for these data. More generally, metric compatibility gives $dg(V,V)/d\varphi =({\nabla }_{{\partial }_{\varphi }}g)(V,V)+2g({\nabla }_{{\partial }_{\varphi }}V,V)=0$. The final change of direction is holonomy, exemplified by parallel transport around a latitude of the unit sphere; parallel transport preserves length without requiring the direction to return unchanged.

For a principal bundle $\pi :E\to B$, the vertical space is $V_e=\ker d{\pi }_e$. A principal connection specifies smooth complements $T_{e}E=H_e\oplus V_e$ with $(R_g{)}_{\ast }{H}_e=H_{eg}$. Equivalently its Lie-algebra-valued connection one-form $\mathcal{A}$ satisfies $\mathcal{A}({\xi }_E)=\xi$ and $R_g^{\ast }\mathcal{A}={\mathrm{Ad}}_{g^{-1}}\mathcal{A}$. Every base vector has a unique horizontal lift because $d\pi :H_e\to T_{\pi (e)}B$ is an isomorphism. A base curve lifts from a chosen initial fiber point by solving $\mathcal{A}(\dot{e})=0$; this defines parallel transport. Closed curves can give nontrivial holonomy, and failure of horizontal fields to close under brackets is measured by the curvature of a principal connection.

For a smooth strictly convex rigid body, assume contact with the plane is unique and varies smoothly with orientation. Write a configuration as $(R,c_{\parallel })\in SO(3)\times {\mathbb{R}}^2$. Contact determines the reference point's height $h(R)$, so the full reference position is $c=(c_{\parallel },h(R))$. Let $r(R)$ be the spatial vector from that point to the contacting material point and let $\hat{\Omega }=\dot{R}{R}^{-1}$ encode the spatial angular velocity. The instantaneous contact velocity isThe no-slip condition is its vanishing. Its vertical component follows already from differentiating contact: the normal component of the changing contact point along the body surface is zero, giving $\dot{h}=-(\Omega \times r{)}_z$. The two remaining equations areTranslations of $c_{\parallel }$ define a free right ${\mathbb{R}}^2$ action, with quotient $SO(3)$. The translational connection for a convex body rolling on a plane isHere $\Omega$ denotes the vector-valued right Maurer-Cartan form on $SO(3)$. This one-form is invariant under translations and returns a translation vector on a vertical translation, so it satisfies both principal connection axioms. A prescribed orientation path has the unique rolling horizontal liftThis is a kinematic connection; inertia and gravity determine which of the allowed paths is dynamically realized.

For a sphere of radius $a$, $r=-an$, where $n$ is the upward unit normal. The same no-slip condition gives ${\dot{c}}_{\parallel }=a\Omega \times n$. It leaves the component $\Omega \cdot n$ arbitrary. If one additionally forbids twisting about $n$, then $\Omega =n\times {\dot{c}}_{\parallel }/a$ and a plane path uniquely lifts to an orientation path by $\dot{R}=\hat{\Omega }R$. This is the familiar $SO(3)$ principal connection over the plane; its horizontal lifts have noncommuting rotation generators and hence nonzero curvature. No-slip alone does not supply that latter two-dimensional horizontal distribution. Corners or nonsmooth changes of contact require piecewise treatment rather than the smooth bundle construction.

By definition $H_A^0=\ker d_A$ consists of parallel adjoint sections. Evaluation at one point identifies it with the subspace of $\mathfrak{su}(2)$ fixed by the full holonomy group. For holonomy equal to a maximal circle, write its elements as $\mathrm{diag}(\lambda ,{\lambda }^{-1})$. The fixed Lie algebra isThe off-diagonal real plane rotates with weight two and has no nonzero vector fixed by the whole circle. A parallel section is uniquely determined by its value, and every invariant value gives a globally well-defined parallel section by parallel transport. Therefore ${\dim }_{\mathbb{R}}{H}_A^0=1$. The full-circle holonomy hypothesis is essential: central holonomy would fix all of $\mathfrak{su}(2)$.

Use $D_x={\partial }_x+A_x$, $D_y={\partial }_y+A_y$, acting on columns. Expanding the two compositions and cancelling their ordinary derivative terms givesCompatibility means local solvability for arbitrary initial vectors, equivalently the existence of an invertible parallel bundle frame $S$ with $D_{x}S=D_{y}S=0$. Such a bundle frame gives $F_{xy}S=0$, hence $F_{xy}=0$. A single specified parallel vector only implies $F_{xy}v=0$: for example $A_x=0$, $A_y=x\mathrm{diag}(1,0)$ and $v=(0,1{)}^T$ have $D_{x}v=D_{y}v=0$ although $F_{xy}=\mathrm{diag}(1,0)\ne 0$. This is why arbitrary-data compatibility is the necessary interpretation of “consistent.”

Conversely assume $F_{xy}=0$. On $x=0$, solve ${\partial }_{y}S=-A_{y}S$ with $S(0,0)=1$. For each $y$, extend it along $x$ by ${\partial }_{x}S=-A_{x}S$. Linear ordinary differential equations with smooth coefficients give global invertible matrices on these finite coordinate paths. Put $R=D_{y}S$. Since $D_{x}S=0$,Uniqueness of the homogeneous transport equation gives $R=0$. Thus $S$ is a full parallel bundle frame. This proves the parallel-frame compatibility criterion without assuming a nonzero vector spans the whole fiber.

Geometrically $A=A_{x}dx+A_{y}dy$ is a gauge connection on the trivial real rank-$n$ vector bundle, and $F_{xy}dx\wedge dy$ is its gauge curvature. The displayed nonlinear equation is flat connection. Writing $g=S^{-1}$ gives its general solution on ${\mathbb{R}}^2$:Conversely substitution, or the Maurer-Cartan equation, proves that every such gauge potential is flat. The parallel sections are $v=g^{-1}c$ for constant columns $c$. Multiplying $g$ on the left by a constant matrix leaves $A$ unchanged. On a nonsimply connected domain a flat connection can have nontrivial holonomy, so a single-valued global pure-gauge expression needs an extra holonomy condition; the specified plane has no such obstruction.

For the Bogomolny equations, complexify the auxiliary vector space and introduce the spectral parameter $\lambda$. Let $\Phi$ denote multiplication by the Higgs matrix, while $D_i\Phi$ denotes its adjoint gauge covariant derivative. Define the Lax pair for the Bogomolny equationsThe auxiliary equations are $L_0(\lambda )\psi =L_1(\lambda )\psi =0$. To check their compatibility, use $[D_i,D_j]=F_{ij}$ and $[D_i,\Phi ]=D_i\Phi$. PutThe constant, linear and quadratic terms in the matrix commutator are respectivelyThus vanishing for every $\lambda$ forces $a=0$, $b=0$, $c=0$, and those conditions also make the matrix commutator zero. They are exactly $\frac{1}{2}{ϵ}_{ijk}{F}_{jk}=D_i\Phi$. No spectral-parameter derivative is needed. Homogenizing in the two projective coordinates of $\lambda \in {\mathbb{CP}}^1$ includes the point at infinity.

The dimensional-reduction interpretation is equally explicit. On Euclidean ${\mathbb{R}}^4$ with orientation $1234$, take fields independent of $x^4$ and set $A_4=-\Phi$. Then $F_{i4}=-D_i\Phi$. Anti-self-duality of gauge curvature says $F_{12}=-F_{34}$, $F_{23}=-F_{14}$, and $F_{31}=-F_{24}$, which reduce to the same three Bogomolny equations. Translation symmetry reduction therefore turns the four-dimensional ASDYM zero-curvature system into the parameter-dependent three-dimensional system above.

Let $\mathcal{J}=\star J$ be the current $(d-1)$-form on an oriented $d$-dimensional spacetime. For a slab bounded by two spacelike hypersurfaces ${\Sigma }_1,{\Sigma }_2$ and a side boundary $B$, the Stokes theorem givesThus $Q(\Sigma )={\int }_{\Sigma }\mathcal{J}$ is conserved if the side flux vanishes, for example for a spatially compact current or appropriate falloff at infinity. With side flux it instead obeys the corresponding charge-balance equation. This is the differential-form version of current conservation.

For the Abelian Chern--Simons theory, take $A\mapsto A+d\chi$ with a smooth compactly supported gauge parameter. The change in the gauge-field term is a boundary term because $d\chi \wedge dA=d(\chi dA)$. For the source,Discarding the boundary terms, the action variation isGauge invariance for every such $\chi$ and $c\ne 0$ therefore gives $d\star J=0$. If $c=0$, the current decouples and gauge invariance places no condition on it. On a spacetime with boundary, boundary conditions or boundary degrees of freedom are also needed to handle the discarded terms; on a nontrivial compact gauge bundle, large-gauge invariance is an additional global issue.

For the field equation, useThe one-form $\delta A$ commutes with the two-form $\mathcal{J}$ under the wedge product, soup to the boundary term. ConsequentlyApplying $d$ and using $d^2=0$ again gives current conservation for $c\ne 0$. The factor two comes from varying both occurrences of $A$.

Orient a spatial region $D$ so its boundary is $\gamma$, and define $Q_D={\int }_D\mathcal{J}$. Applying Stokes theorem and the field equation yieldsFor a U(1) connection, in a convention with unit minimal charge, the gauge-invariant quantity is its holonomy $e^{i\Phi }$, the Wilson-loop or Aharonov-Bohm phase. A large gauge transformation can change the chosen representative of $\Phi$ by $2\pi n$. Therefore $\Phi$ is a phase angle modulo $2\pi$, rather than an absolute gauge-invariant real number. A particle of charge $q$ has phase $e^{iq\Phi }$ in the corresponding normalization. The flux-charge relation thus attaches a gauge phase to enclosed charge.

The metric variation requires specifying the independent source. The pure Chern-Simons term $A\wedge dA$ has no metric dependence, hence contributes zero stress-energy tensor. If the one-forms named in the question have fixed covariant components $A_{\mu },J_{\mu }$, the Hodge star in the source does depend on the metric:Using $\delta \sqrt{-g}=-\sqrt{-g}{g}_{\mu \nu }\delta g^{\mu \nu }/2$ givesUnder that literal fixed-one-form convention,There is another common convention in the topological source theory: hold the conserved two-form $\mathcal{J}=\star J$, equivalently the vector current density, fixed as the metric varies. Then both $A\wedge dA$ and $\mathcal{J}\wedge A$ are metric-independent, and $T_{\mu \nu }=0$ for this action. These are different variations, not contradictory calculations. The Chern-Simons source stress convention explains why a source prescription is necessary; a dynamical matter source would contribute its own action and stress as well.

Geometric quantization starts from a classical symplectic manifold $(P,\omega )$ and seeks a Hilbert space and quantum observables that reflect its geometry. It has two distinct stages: prequantization, which represents the full Poisson algebra on sections of a line bundle, and a choice of polarization in geometric quantization, which reduces the excessive phase-space dependence of those sections.

Use the conventions ${\iota }_{X_f}\omega =df$ and $\{f,g\}=\omega (X_f,X_g)$. A prequantum line bundle is a Hermitian line bundle $L\to P$ with unitary connection of curvatureExistence requiresfor every closed integral two-cycle $\Sigma$, equivalently integrality of the symplectic cohomology class. Necessity follows from the integral curvature periods of a unitary line bundle; conversely an integral class defines such a bundle, and one adjusts its connection to give the specified curvature. This is a restriction on the classical system, not a choice of local coordinates. For an exact cotangent symplectic form $\omega =-d\theta$, with $\theta =p_{j}d{q}_j$, the bundle can be trivial and a connection is $\nabla =d-i\theta /\hslash$. On a sphere with symplectic area $4\pi s$, the condition becomes $2s/\hslash \in \mathbb{Z}$. The sign of the curvature convention changes if one changes the Hamiltonian-vector-field convention or uses the dual bundle; the integrality condition is unaffected.

The Kostant-Souriau prequantum operator for a real observable isIts derivative part describes the classical Hamiltonian flow and the multiplication term corrects the bracket. Indeed,Since $X_f(g)=-\{f,g\}$, expanding the operator commutator givesWith the Hermitian bundle metric and Liouville measure, these operators are formally symmetric for real $f$ because Hamiltonian flows preserve symplectic volume. On noncompact spaces, domains and self-adjointness still require attention; the algebraic identity initially holds on an appropriate common smooth domain.

Prequantization alone is too large. Sections depend on $2n$ phase-space variables, and the canonical commutation relations are represented reducibly. A polarization is an involutive maximal isotropic distribution $\mathcal{F}\subset T_{\mathbb{C}}P$ of complex rank $n$. Polarized sections satisfyThe curvature vanishes on pairs of polarization vectors because $\omega {|}_{\mathcal{F}}=0$, and involutivity makes the equations locally compatible. A real polarization is tangent to a Lagrangian foliation. A complex polarization in a Kähler manifold chooses one complex tangent type. With the curvature sign used here, choose the holomorphic tangent distribution, giving antiholomorphic wavefunctions. The conjugate line-bundle and sign convention gives the more usual holomorphic description.

For $T^{\ast }Q$, the vertical real polarization is spanned locally by ${\partial }_{p_j}$. Since the canonical one-form has no vertical component, its polarized sections are functions of $q$ alone. In the above trivialization,One does not simply integrate these states over every momentum to retain the prequantum norm: they are constant along noncompact polarization leaves and that integral would diverge. The polarized inner product instead uses the configuration-space measure or an intrinsic half-density description. This recovers the Schrödinger realization.

Only observables whose Hamiltonian flows preserve $\mathcal{F}$ act directly on the polarized space. For example, the free-particle flow generated by $p^2/2$ sends vertical tangent vectors into directions with a configuration-space component, so its raw prequantum operator does not preserve the vertical polarization. To obtain the usual kinetic operator one needs an additional procedure, such as transporting the polarization and pairing back, rather than pretending every classical observable restricts automatically. The Blattner-Kostant-Sternberg pairing and related projection methods address this problem. The impossibility of quantizing all classical observables with every desired algebraic property is another reason to specify the admissible observable algebra.

A metaplectic correction tensors $L$ with a square root of the polarization's canonical determinant bundle when such a square root exists. These half-forms make the pairing more intrinsic and supply corrections to operator transport. They account for familiar Maslov and zero-point shifts; their existence is an extra topological condition. For a flat phase plane, take $z=q+ip$ and replace $\theta =pdq$ by the gauge-equivalent ${\theta }_s=(pdq-qdp)/2$. ThenThe antiholomorphic function therefore has Gaussian norm, the complex-conjugate Bargmann-Fock space. This explicit calculation fixes which polarization matches our curvature sign. Choosing the opposite tangent type with the same sign would give growing Gaussian sections, not the desired Hilbert space. On compact positive Kähler examples, the corresponding holomorphic description (or its conjugate in these conventions) gives finite-dimensional state spaces.

Finding a suitable polarization is therefore a substantial part of the construction. A globally smooth real Lagrangian foliation may fail to exist; for example a regular real polarization on $S^2$ would require a line field that its tangent topology forbids. Singular foliations and noncompact leaves complicate both states and inner products. Although the restricted prequantum connection is locally flat on a Lagrangian leaf, a global nonzero parallel section requires trivial holonomy. In an integrable system this selects the Bohr-Sommerfeld quantization leaves; including the half-form correction gives the familiar action conditionin the usual Maslov-index convention. Complex polarizations require compatible global complex geometry and positivity. Different choices can lead to quite different realizations, and a natural unitary comparison is not automatic. Thus integral curvature provides prequantization; polarization and its global compatibility determine the actual quantum state space.

For $n\ne 0$, put $\rho =f(r)/|n|$ and $\theta =|n|\varphi$. Then $d{\rho }^2=d{r}^2$ andThese give a local isometry wherever $\rho >0$. For $n=0$, $f=b>0$, and the local coordinates $X=r$, $Y=b\varphi$ similarly give the Euclidean plane metric tensor. Equivalently, the Gaussian curvature of the original Riemannian metric is $-f^{\prime \prime }/f=0$ throughout this family.

A local isometry need not be a global isometry. If $|n|=1$, the displayed polar map is injective modulo $\varphi \sim \varphi +2\pi$ and identifies the regular domain with a Euclidean annulus, possibly a punctured disc or the punctured Euclidean plane, according to the radial interval. If $|n|>1$, it is a $|n|$-fold covering of that annulus: distinct angles differing by $2\pi /|n|$ have the same image. This developing map cannot be made injective by a different choice of Euclidean coordinates, since two local Euclidean isometries on a connected flat domain differ by a rigid motion once their value and differential are fixed at one point. For $n=0$, the domain is a flat cylinder; its circular curves are closed geodesics, whereas the Euclidean plane has no nonconstant closed geodesic.

All these metrics are locally Euclidean; they are not all globally Euclidean. Even the regular $|n|=1$ domain is an annular region rather than the entire Euclidean plane. A regular axis can be added without a conical singularity only when $|n|=1$; with the full radial range this completion becomes the Euclidean plane. Trivial holonomy alone therefore does not supply a global injective Euclidean chart.

If every regular circle has trivial holonomy, the preceding calculation says that the continuous function $f^{\prime }$ takes values in $\mathbb{Z}$. On a connected radial interval it must therefore be constant. Integrating gives the complete familyConversely, every such function has identity holonomy around every circle. With the opposite sign convention for $f$, replace the positivity condition by nonvanishing; the Riemannian metric is unchanged by changing its sign.

The slope parameter is discrete. Thus the printed parameter count can mean one integer and one real parameter, but it cannot mean two freely varying real parameters. Allowing $f=ar+b$ with arbitrary real $a$ gives a locally flat family, not trivial circular holonomy: for example $a=1/2$ gives a rotation by $\pi$. This distinction is captured by flat circular metrics with trivial holonomy.

Work on a connected regular radial interval, where $f$ is smooth and nonzero; replacing $f$ by $|f|$ allows us to take $f>0$. The only nonzero Christoffel symbols areAlong the circle, write a tangent vector in the single-valued orthonormal basis $E_r={\partial }_r$, $E_{\varphi }=f^{-1}{\partial }_{\varphi }$ as $V=u{E}_r+v{E}_{\varphi }$. The equations for parallel transport reduce toConsequently $u+iv$ acquires the factor $e^{-2\pi i{f}^{\prime }(r_0)}$ after one circuit. The basis itself agrees at $\varphi =0$ and $2\pi$, so no extra basis change is missing. The holonomy is the identity on every tangent vector precisely whenThe first condition makes this a regular circle of the Riemannian metric; the second is the actual transport condition. In particular, zero slope is sufficient but not necessary. This is holonomy around circular fibres.

Ordinary component differences at neighboring points depend on how the local bases vary. The covariant derivative corrects for that variation:Each upper tensor index adds a connection term and each lower index subtracts one. A scalar has ${\nabla }_{a}f={\partial }_{a}f$, while the derivative of a tensor transforms as a tensor with one additional covariant index. The Christoffel symbols themselves are not tensor components; their inhomogeneous transformation cancels that of the ordinary derivative.

For the Levi-Civita connection,These express metric compatibility and zero torsion. Along $x^a(\lambda )$, parallel transport is the initial value problemIt preserves inner products for the Levi-Civita connection. Parallel transport of the tangent itself gives the affinely parametrized geodesic equation, ${\ddot{x}}^a+{\Gamma }_{bc}^a{\dot{x}}^b{\dot{x}}^c=0$.

Transport generally depends on the path. On a round sphere, transport a tangent vector around the triangle bounded by two meridians $90$ degrees apart and the intervening equatorial arc. All three sides are geodesics; the enclosed spherical area is $\pi a^2/2$. The returning vector is rotated by $90$ degrees, with sign set by traversal orientation. Locally this holonomy is controlled by curvature, even though the connection can be made zero at a single point. A vector can therefore have changing coordinate components while remaining parallel, and constant components do not generally mean parallel transport.

This construction compares directions at separate points through a chosen connection, rather than assuming that a curved manifold supplies a path-independent comparison.^[ref]