3.1 Computation of steady axisymmetric flow

We convert the NSEs into stream function–vorticity form. From (2.1a,b ), the steady vorticity transport equation is

where ${\bf\omega}=\boldsymbol{{\rm\nabla}}\times \boldsymbol{v}$ is the vorticity vector. A stream function ${\it\psi}$ is defined in cylindrical coordinates, such that

where ${\it\rho}$ is the distance from the $z$ -axis and $v_{{\it\rho}}$ and $v_{z}$ represent the fluid velocity components.

Combining (3.2a,b ) with (3.1) gives

where ${\it\omega}$ is the component of ${\bf\omega}$ in the azimuthal direction ${\it\varphi}$ about the $z$ -axis; the other ( ${\it\rho}$ and $z$ ) components of ${\bf\omega}$ vanish by symmetry. Expressing ${\it\omega}$ in terms of ${\it\psi}$ gives

Equations (3.3) and (3.4) are coupled partial differential equations, with the former being nonlinear. These may be simultaneously solved for the scalar quantities ${\it\psi}$ and ${\it\omega}$ to give the flow field given appropriate boundary conditions.

In a co-moving frame, the squirmer surface ( $r=1$ ) is a streamline with tangential velocity given by (1.2). Thus, ${\it\psi}|_{r=1}=0$ and $[\boldsymbol{{\rm\nabla}}{\it\psi}\boldsymbol{\cdot }\boldsymbol{n}]_{r=1}=v_{s}\sin {\it\theta}=3\sin ^{2}{\it\theta}(1+{\it\beta}\cos {\it\theta})/2$ . The values of ${\it\beta}$ and $Re$ are specified constants, so the swimming stroke is represented as a fixed boundary condition. By axisymmetry, ${\it\psi}|_{{\it\rho}=0}=0$ and ${\it\omega}|_{{\it\rho}=0}=0$ on the $z$ -axis. Finally, the flow is uniform in the far-field, so ${\it\psi}|_{r\rightarrow \infty }=-U{\it\rho}^{2}/2$ and ${\it\omega}|_{r\rightarrow \infty }=0$ .

A spectral element method (Karniadakis & Sherwin Reference Karniadakis and Sherwin2005) was used to spatially discretize (3.3), (3.4) and the boundary conditions. The shape functions were defined as a tensor product of $N$ th order Lagrange polynomials supported at the $N+1$ Gauss–Lobatto integration points over the square $[-1,1]^{2}$ parametric space of each quadrilateral element. Integration over each element was carried out using the corresponding Gauss–Lobatto quadrature rule to produce a system of nonlinear algebraic equations. This system was solved iteratively using Newton–Raphson iteration. Iteration was terminated when the $L^{2}$ -norm of the relative errors in ${\it\psi}$ and ${\it\omega}$ over all discretization points was reduced below $10^{-6}$ .

Figure 2.

Structured polar grids (a) were used for both axisymmetric and 3-D (by revolving them azimuthally) computations of the flow. For axisymmetric computations with $Re>10$ , a different mesh (b) was used with greater resolution in the wake of the squirmer to more accurately resolve the details of the flow in this region.

The spatial domain was discretized into high-order computational grids using the software package ‘Gmsh’ (Geuzaine & Remacle Reference Geuzaine and Remacle2009). Three different grids were used depending on the value of $Re$ . For $Re\leqslant 0.1$ , a polar grid extending to $R_{\infty }=1000$ and consisting of $9\times 9$ node quadrilateral elements was used. The elements were distributed evenly in the ${\it\theta}$ -direction ( $N_{{\it\theta}}=10$ ) and progressed geometrically outward in the $r$ -direction ( $N_{r}=20$ ). A similar grid was used for $0.1\leqslant Re\leqslant 10$ , with $R_{\infty }=100$ (figure 2 a). For $Re>10$ , a different mesh was used to provide better resolution in the squirmer wake. Here, a boundary layer grid was used along the squirmer surface, with $N_{r}=10$ and $N_{{\it\theta}}=51$ , extending $R_{BL}=0.25$ radii from the squirmer surface (here the subscript $BL$ stands for boundary layer). The radial grid size grew geometrically with $r$ , and was initially ${\it\Delta}_{r0}=0.01$ at the squirmer surface. The remainder of the grid was unstructured, with upstream boundaries extending to $R_{\infty }=32$ , and a rectangular wake region extending a distance of $100$ radii behind the squirmer (figure 2 b). The far-field boundary conditions were enforced at the exterior boundary of the mesh. We refer the reader to the Appendix for details on grid convergence.

The far-field boundary condition of uniform, oncoming flow cannot be directly applied because the steady-state swimming speed $U$ is unknown a priori. Since the flow is assumed to be steady and axisymmetric, we instead enforce that $F_{z}$ is equal to zero. Expressing (2.2) in terms of ${\it\omega}$ for an axisymmetric flow field gives (Khair & Chisholm Reference Khair and Chisholm2014)

A secant method was used to iteratively compute the value of $U$ at which (3.5) vanishes. At each iteration, the flow is solved with $U=U^{\langle n\rangle }$ , where $n$ is the iteration number, and (3.5) is evaluated to give $F_{z}^{\langle n\rangle }$ . An improved estimate for $U$ is given by linear interpolation: $U^{\langle n+1\rangle }=(U^{\langle n\rangle }F_{z}^{\langle n-1\rangle }-U^{\langle n-1\rangle }F_{z}^{\langle n\rangle })/(F_{z}^{\langle n-1\rangle }-F_{z}^{\langle n\rangle })$ . Iteration was terminated when $|U^{\langle n\rangle }-U^{\langle n-1\rangle }|$ was reduced below $10^{-5}$ .

Computations for each value of ${\it\beta}$ were started initially with $Re=0.01$ . Two initial guesses of the swimming speed are required, which were made as $U^{\langle 0\rangle }=0.99$ and $U^{\langle 1\rangle }=1.01$ , since $U$ is close to unity at small $Re$ . An initial guess for the stream function and vorticity fields of uniformly zero was sufficient for convergence of the computed flow in this case. A simple continuation strategy was employed by incrementally increasing $Re$ . Initial guesses for $U$ and the flow field at a given $Re$ were supplied by using the values computed at the last largest values of $Re$ for which a converged solution was successfully reached.

3.2 Computation of unsteady 3-D flows

Unsteady 3-D flows were explored using the JADIM code described in detail in Magnaudet, Rivero & Fabre (Reference Magnaudet, Rivero and Fabre1995) and Legendre & Magnaudet (Reference Legendre and Magnaudet1998). The JADIM code has been extensively used and validated in previous studies concerning the 3-D flow dynamics of spheroidal and disk-shaped bodies with no-slip (solid) or slip (bubble) surfaces in uniform, shear or turbulent flows (see e.g. Legendre & Magnaudet Reference Legendre and Magnaudet1998; Merle, Legendre & Magnaudet Reference Merle, Legendre and Magnaudet2005; Legendre, Merle & Magnaudet Reference Legendre, Merle and Magnaudet2006; Hallez & Legendre Reference Hallez and Legendre2011). In particular, the wake transition from axisymmetric to 3-D flow for a fixed body has been considered in Mougin & Magnaudet (Reference Mougin and Magnaudet2001), Magnaudet & Mougin (Reference Magnaudet and Mougin2007) and Fabre, Auguste & Magnaudet (Reference Fabre, Auguste and Magnaudet2008). In the case of a sphere, a first bifurcation resulting in loss of axial symmetry in the flow is detected at a critical Reynolds number (based on the sphere radius and speed of translation $U$ ) of $Re_{U}^{(c1)}=105$ , in agreement with linear stability analysis (Natarajan & Acrivos Reference Natarajan and Acrivos1993) and previous numerical studies (Johnson & Patel Reference Johnson and Patel1999; Tomboulides & Orszag Reference Tomboulides and Orszag2000). A second (Hopf) bifurcation is observed at $Re_{U}^{(c2)}=135$ , leading to time-dependent flow, which is also in good agreement with previous numerical findings (Johnson & Patel Reference Johnson and Patel1999; Tomboulides & Orszag Reference Tomboulides and Orszag2000), according to which the Hopf bifurcation lies in within the range $135<Re_{U}^{(c2)}<137$ . In Magnaudet & Mougin (Reference Magnaudet and Mougin2007), the vortex shedding process for a sphere at $Re_{U}=150$ corresponds to a Strouhal number of $St_{U}=fa/U=0.0665$ , where $f$ is the dimensional frequency of vortex shedding. This falls within 2–3 % of that reported by Johnson & Patel (Reference Johnson and Patel1999) and Tomboulides & Orszag (Reference Tomboulides and Orszag2000) for the same $Re$ .

Briefly, the JADIM code solves the incompressible NSEs (2.1a,b ) in terms of velocity and pressure variables. The spatial discretization employs a staggered grid on which the equations are integrated using a second-order accurate finite-volume method. Fluid incompressibility is satisfied after each time step by solving a Poisson equation for an auxiliary potential. Time advancement is achieved through a second-order accurate Runge–Kutta/Crank–Nicholson algorithm. At each time step, the swimming speed $U$ is updated by integrating (2.2). For each simulation, the squirmer was started from rest with swimming stroke (1.2) and allowed to accelerate. Simulations were terminated after a steady time-averaged value of the swimming speed was reached.

A polar grid extending to $R_{\infty }=150$ and rotated around the $z$ -axis was used for computation (figure 2 a). Nodes were distributed uniformly in the ${\it\theta}$ -direction and in a geometric progression in the $r$ -direction. The effect of the number of nodes ( $N_{r}=150$ along the radial direction, $N_{{\it\theta}}=250$ along the polar direction and $N_{{\it\varphi}}=64$ along the azimuthal direction), as well as $R_{\infty }$ and the radial grid size ${\it\Delta}_{r0}=0.001$ at the body surface, were checked in order to ensure grid independence of the results (see the Appendix).

The transition from steady axisymmetric to unsteady 3-D flow was investigated by running the simulation for a given period of time while allowing numerical error to perturb the initially axisymmetric flow profile. If the flow is unstable for a given ${\it\beta}$ and $Re$ , such perturbations are expected to grow over time, resulting in a flow field that is potentially 3-D and/or unsteady. Such is the case for a no-slip sphere in uniform flow, where distinct axisymmetric, steady 3-D and unsteady 3-D flow regimes are respectively encountered as $Re$ is increased (Natarajan & Acrivos Reference Natarajan and Acrivos1993; Tomboulides & Orszag Reference Tomboulides and Orszag2000). Specifically, simulations were performed with $Re$ increased in coarse increments until a transition, if one occurred, was identified. Then, $Re$ was increased in finer increments within the interval in which the transition occurred. This process was repeated until a satisfactory estimate of the critical transition $Re$ was procured. The simulation time was increased as the critical $Re$ of transition was approached, as it generally required longer times for perturbations to grow and hence for the flow to reach a final transitioned state.

4.1 Swimming speed of a squirmer

The calculated swimming speed $U$ versus $Re$ of a squirmer with ${\it\beta}=\pm 0.5$ and $\pm 5$ is shown in figure 3. There, $U$ is normalized by $2B_{1}/3$ , which is the swimming speed at $Re=0$ for all ${\it\beta}$ , or the swimming speed of a neutral squirmer at arbitrary $Re$ . At $Re=0$ , $U$ is independent of ${\it\beta}$ because the equations governing the flow are linear. Thus, the two terms in the swimming stroke (1.2) contribute to the flow field independently; only the first (treading) term generates propulsion, while the second only produces vorticity. This is not the case as $Re$ is increased from zero: pushers (pullers) monotonically increase (decrease) in speed if $Re\lesssim O(1)$ , in agreement with results from asymptotic analyses (Wang & Ardekani Reference Wang and Ardekani2012; Khair & Chisholm Reference Khair and Chisholm2014). The increase or decrease in swimming speed is amplified as $|{\it\beta}|$ increases. However, as $Re$ is increased beyond an $O(1)$ value, significantly different behaviour of pushers versus pullers is observed. For all pushers and pullers with ${\it\beta}<1$ , $U$ continues to vary monotonically with increasing $Re$ , eventually reaching a terminal value. The computed swimming speed is nearly identical for axisymmetric and 3-D computations, suggesting that there is no departure from steady axisymmetric flow. In contrast, a non-monotonic trend is observed for pullers with ${\it\beta}>1$ , and no limiting value for $U$ is apparent up to $Re=1000$ . Moreover, the axisymmetric and 3-D computations give drastically different results, suggesting the destabilization of the axisymmetric steady flow (see § 4.3 for more detail). We remind the reader that $Re$ for a squirmer is defined as $2{\it\varrho}B_{1}a/(3{\it\mu})$ , in contrast to the Reynolds number based on the translational speed $U$ , which we denote $Re_{U}={\it\varrho}Ua/{\it\mu}$ . Note that $Re$ and $Re_{U}$ are the same order of magnitude since $U\sim O(1)$ .

Figure 3.

Swimming speed, $U$ , normalized by $2B_{1}/3$ , for ${\it\beta}=\pm 0.5$ (a) and $\pm 5$ (b). The ‘▿’ markers represent pushers and the ‘▵’ markers represent pullers. Hollow markers represent steady axisymmetric solutions, and filled markers represent unsteady 3-D solutions. We follow these conventions for the remainder of the article. The dashed line represents the speed of a neutral ( ${\it\beta}=0$ ) squirmer. For a ${\it\beta}=+5$ puller, the steady axisymmetric flow destabilizes at $Re\approx 20$ , and hence the steady axisymmetric and unsteady 3-D solutions diverge. Time averages of $U$ are taken in the case of unsteady flow. Dotted lines show the asymptotic result of Khair & Chisholm (Reference Khair and Chisholm2014) for $U$ to $O(Re^{2})$ .

Figure 4.

Streamlines of axisymmetric flow past a squirmer with ${\it\beta}=\pm 5$ ; (a,c,e,g) pusher: ${\it\beta}=-5$ , (b,d,f,h) puller: ${\it\beta}=5$ ; $Re=0.1$ (a,b), 10 (c,d), 100 (e,f), 1000 (g,h). Dashed streamlines represent negative values of the stream function. The tick marks in (g) at $Re=1000$ follow along streamlines of irrotational flow past a sphere.

Distinct contributions to the thrust and drag on a squirmer are provided by the two terms on the right-hand side of (3.5). The first term, which equals $8{\rm\pi}Re{\it\beta}/15$ after integration, depends solely on the swimming stroke and vanishes when $Re=0$ . The second term also vanishes if $Re=0$ due to the antisymmetric, purely diffusive, distribution of the vorticity, and it is hence associated with forces arising from the flow asymmetry produced by inertia at finite $Re$ . Thus, (3.5) is satisfied identically in Stokes flow, and a squirmer propels itself at the same speed regardless of ${\it\beta}$ . However, pushers increase in speed with $Re$ while pullers decrease at finite $Re\lesssim O(1)$ . In the former case, the first term represents a drag force because it is negative when ${\it\beta}<0$ . Thus, the redistribution of vorticity caused by inertia is responsible for the extra thrust that increases the swimming speed with $Re$ . The opposite occurs for a puller, where ${\it\beta}>0$ : the contribution of the first term is a thrust, but it is outweighed by drag produced by the inertial redistribution of vorticity. As $Re$ is increased beyond an $O(1)$ value, the monotonic trend continues for a pusher until a limiting speed is reached. In contrast, the swimming speed of a puller becomes non-monotonic. A fuller explanation of these trends, especially when $Re$ is large, requires a closer examination of the flow fields generated by squirmers and how they differ for pushers versus pullers.

4.2 Flows generated by pushers and pullers

Streamlines illustrating the steady axisymmetric flow around pushers and pullers are shown in figures 4 and 5, and contours of constant vorticity are shown in figures 6 and 7. At $Re=0.1$ , symmetries in the near-field flow are apparent due to the dominance of viscous forces over inertial forces: reversing the sign of ${\it\beta}$ causes the streamlines to be mirrored along the ${\it\rho}$ -axis. Also, the vorticity is fore–aft antisymmetric. Pushers generate positive vorticity ahead of their direction of travel and negative vorticity behind, while pullers do the opposite. Closed-streamline recirculatory regions appear in front of pushers and behind pullers if $|{\it\beta}|>1$ . Streamlines separate from the squirmer surface at the point where the stroke $v_{s}({\it\theta})$ changes sign (Magar et al. Reference Magar, Goto and Pedley2003).

Figure 5.

Similar to figure 4 except with ${\it\beta}=\pm 0.5$ ; (a) pusher: ${\it\beta}=-0.5$ ; $Re=1000$ , (b) puller: ${\it\beta}=0.5$ ; $Re=1000$ .

Figure 6.

Vorticity contours for axisymmetric flow with ${\it\omega}=\pm \{0.1,0.2,0.5,1,2,5,\ldots ,200\}$ . Dashed lines represent negative values. The dotted line in (h) at $Re=1000$ encircles the region where there is an approximately constant value of ${\it\omega}/{\it\rho}=2.9\pm 0.05$ , indicating that the wake bubble behind a ${\it\beta}=5$ puller has a structure resembling a Hill’s vortex; (a,c,e,g) pusher: ${\it\beta}=-5$ , (b,d,f,h) puller: ${\it\beta}=5$ ; $Re=0.1$ (a,b), 10 (c,d), 100 (e,f), 1000 (g,h).

Figure 7.

Similar to figure 6 except with ${\it\beta}=\pm 0.5$ ; (a) pusher: ${\it\beta}=-0.5$ ; $Re=1000$ , (b) puller: ${\it\beta}=0.5$ ; $Re=1000$ .

Figure 8.

The maximum value of $|{\it\omega}|$ , normalized by $|{\it\beta}|$ , from axisymmetric computations at ${\it\beta}=\pm 0.5$ and $\pm 5$ and from 3-D computations for ${\it\beta}=+5$ . In Stokes flow, $\max |{\it\omega}/{\it\beta}|=9|{\it\beta}|/4$ , as indicated by the dashed line to which the data collapses as $Re\rightarrow 0$ . The dotted line indicates a slope of one-half, revealing that ${\it\omega}\sim \sqrt{Re}$ at large $Re$ .

The flow patterns and swimming speed observed as $Re$ is increased depend critically on ${\it\beta}$ . For pushers at $Re\gg 1$ , the majority of the vorticity, along with the upstream closed-streamline region that is present if ${\it\beta}<-1$ , is concentrated into a laminar boundary layer of thickness $O(1/\sqrt{Re})$ . This vorticity is then transported into a narrow downstream wake due to the motion of the swimming stroke (figures 6 a,c,e,g and 7 a). The streamlines outside the boundary layer and wake tend toward a potential flow profile, and no standing wake eddy is present (figures 4 a,c,e,g and 5 a). Thus, the flow around a pusher apparently resembles that past an inviscid spherical bubble at large $Re_{U}$ , which exhibits the same characteristics (Moore Reference Moore1963; Leal Reference Leal1989). The key similarity is that the mobile surfaces of a bubble and a pusher squirmer cause advection of vorticity downstream, thus preventing it from accumulating into a recirculating wake. However, for a bubble, the shear-free surface produces ${\it\omega}\sim O(1)$ , whereas for a squirmer, ${\it\omega}\sim O(\sqrt{Re})$ in the boundary layer (figure 8) due to the fixed nature of the surface velocity profile (swimming stroke). This is akin to a towed, rigid sphere with a no-slip surface, where the greater amount of boundary layer vorticity results in flow separation and the appearance of a wake eddy if $Re_{U}\gtrsim 10$ , which grows with $Re_{U}$ (Dennis & Walker Reference Dennis and Walker1971; Fornberg Reference Fornberg1988). These phenomena are avoided by a streamlined no-slip body, but for a pusher, the strong vorticity advection due to the propulsive surface motion interestingly achieves a similar effect. Despite the bluff body shape and $O(\sqrt{Re})$ surface vorticity of a pusher, no wake eddy is produced.

The flow around pullers with $0<{\it\beta}<1$ may be described likewise. As $Re$ is increased, the boundary layer and wake become smaller in extent, and the majority of the flow domain becomes irrotational (figures 5 b and 7 b). Again, vorticity is efficiently swept downstream by the mobile surface with a (monotonic) swimming stroke $v_{s}({\it\theta})$ that is directed along the path of the flow. Consequently, pushers and pullers with ${\it\beta}<1$ reach a terminal (dimensionless) swimming speed (i.e. a dimensional swimming speed that is proportional to $B_{1}$ ).

The axisymmetric flow that is observed around pullers with ${\it\beta}>1$ as $Re$ increases is very different. A trailing vortical wake bubble is indeed present and grows with $Re$ (figure 4 b,d,f,h). Thus, for pullers with ${\it\beta}>1$ , the flow does not become irrotational within the majority of the flow domain as $Re$ becomes large. As a result, the swimming speed of a ${\it\beta}>1$ squirmer does not attain a terminal value. The wake eddy is caused by the reversal of $v_{s}$ along the rear half of the squirmer surface, which hinders the advection of vorticity downstream, and causes its accumulation behind the squirmer. This resembles flow past a rigid bluff body towed by an external force, where fluid deceleration along the no-slip surface has the same effect. Indeed, if the flow is restricted to be axisymmetric, the wake bubble resembles a Hill’s spherical vortex at $Re=1000$ , where ${\it\omega}/{\it\rho}$ is constant in the region of closed streamlines and ${\it\omega}=0$ elsewhere (figure 6 h). Batchelor (Reference Batchelor1956) proposed that such flow structures exist in the wake of bluff bodies in steady axisymmetric flow at large $Re$ . The computations of Fornberg (Reference Fornberg1988) show the presence of a Hill’s vortex-like wake structure behind a sphere held fixed in a uniform flow, within which ${\it\omega}/{\it\rho}$ is nearly constant once $Re_{U}$ is sufficiently large. Moreover, it is shown that such large $Re$ axisymmetric flows result in very low drag forces relative to that observed in 3-D flows beyond the onset of flow instabilities. The observation that $U$ increases with $Re$ for an axisymmetric ${\it\beta}=5$ pusher when $Re\gtrsim O(1)$ (figure 3) indicates that the trailing vortex behind a ${\it\beta}>1$ puller is analogous to that behind a towed sphere; the wake eddy acts to decrease the overall drag. Note that the point of flow separation along the surface of a squirmer always occurs where $v_{s}({\it\theta})$ changes sign regardless of $Re$ (figure 4), whereas it depends on $Re_{U}$ for a no-slip sphere.

Figure 9.

Flow state as a function of ${\it\beta}$ and $Re$ . The points of transition, marked with an ‘ $\times$ ’ and interpolated by the solid lines, were obtained numerically (see also table 1).

4.3 Transition to 3-D and unsteady flow

Figure 9 and table 1 detail the transition of the flow around a squirmer from steady and axisymmetric to unsteady and 3-D and are derived from unsteady 3-D flow simulations. The critical values of $Re$ at which the axisymmetry breaks ( $Re^{(c1)}$ ) and at which the flow becomes unsteady ( $Re^{(c2)}$ ) are shown. For ${\it\beta}>1$ pullers, $Re^{(c1)}<Re^{(c2)}$ , and a monotonic decrease of $Re^{(c1)}$ and $Re^{(c2)}$ with ${\it\beta}$ is observed. Moreover, $Re^{(c1)}$ and $Re^{(c2)}$ both increase rapidly as ${\it\beta}$ is decreased toward unity such that ${\it\beta}=1$ appears to be an asymptote; pushers and pullers with ${\it\beta}<1$ produce steady axisymmetric flows that remain stable up to at least $Re=1000$ .

This highlights another apparent similarity between the flow past a ${\it\beta}<1$ squirmer and an inviscid spherical bubble. For the latter, the asymptotic analysis of Moore (Reference Moore1963) suggested that a potential flow is recovered as $Re_{U}\rightarrow \infty$ . Specific studies have also been carried out to determine how the wake structure and flow stability vary with aspect ratio for oblate spheroidal bubbles (Dandy & Leal Reference Dandy and Leal1986; Blanco & Magnaudet Reference Blanco and Magnaudet1995; Magnaudet & Mougin Reference Magnaudet and Mougin2007). It was revealed that only bubbles with an aspect ratio larger than 1.65 or 2.21 exhibit a standing wake eddy or an unstable wake, respectively. The reason is that a sufficient amount of vorticity (produced at the bubble surface in an amount proportional to the surface curvature) must accumulate in its wake for these transitions to occur. For a squirmer, a comparatively large $O(\sqrt{Re})$ amount of boundary layer vorticity is generated, whereas it is $O(1)$ for a spherical bubble, so the stability of the flow past pushers and ${\it\beta}<1$ pullers despite this fact is an intriguing result. Again, vorticity is strongly advected downstream by the propulsive surface velocity, preventing its accumulation in the wake, and the stability of the steady axisymmetric flow is preserved.

However, this does not imply stability at all $Re$ . For no-slip objects where the vorticity is similarly $O(\sqrt{Re})$ , turbulent boundary layers develop when $Re$ is very large, even for streamlined objects such as airfoils or flat plates where there is no instability caused by a wake eddy. For example, the boundary layer of a no-slip sphere becomes turbulent at $Re_{U}\approx 10^{5}$ (Deen Reference Deen1998, p. 512). For this reason, it is very possible that the laminar boundary layer of a squirmer will also become turbulent at sufficiently large $Re$ , except, perhaps, in the singular ${\it\beta}=0$ case where potential flow results identically. Such a phenomenon likely occurs well above the maximum $Re=1000$ considered in this work, and hence is not further discussed here.

Table 1.

Numerically obtained critical values of $Re$ where the flow becomes 3-D ( $Re^{(c1)}$ ) and unsteady ( $Re^{(c2)}$ ) (see also figure 9).

Given the previously noted similarities of the steady axisymmetric flow around a ${\it\beta}>1$ puller to that past a no-slip sphere, one might expect that the transitions to 3-D and unsteady flows that occur will also be analogous. This indeed appears to be the case. For a no-slip sphere, the flow first bifurcates at $Re_{U}^{(c1)}\approx 105$ (Natarajan & Acrivos Reference Natarajan and Acrivos1993; Tomboulides & Orszag Reference Tomboulides and Orszag2000), resulting in a steady 3-D flow that exhibits planar symmetry and two counter-rotating vortices in the wake. The symmetry plane passes through the axis of translation, but its orientation is arbitrary due to the initial axisymmetry of the flow. The scenario is the same for ${\it\beta}>1$ pullers, and planar flow symmetry is apparent in figure 10(a). The only difference is that $Re^{(c1)}$ depends on ${\it\beta}$ in the latter case. A second transition from steady to unsteady flow takes place at $Re_{U}^{(c2)}\approx 140$ in the case of a no-slip sphere (Natarajan & Acrivos Reference Natarajan and Acrivos1993; Tomboulides & Orszag Reference Tomboulides and Orszag2000), and the same happens for a ${\it\beta}>1$ puller at $Re^{(c2)}=Re^{(c2)}({\it\beta})$ . In both cases, the planar flow symmetry persists, and as $Re$ (or  $Re_{U}$ ) is further increased, shedding of the wake vortices begins to occur (figure 10 b). Table 1 reveals that the quantity $Re^{(c2)}-Re^{(c1)}$ decreases significantly as ${\it\beta}$ is increased; the difference is approximately 40 at ${\it\beta}=1.5$ and decreases to only $3.2$ at ${\it\beta}=5$ . The flow is more quickly destabilized when the value of ${\it\beta}$ is larger, and hence there is only a narrow range of $Re$ where it exhibits a steady 3-D state.

Figure 10.

The streamwise component of the vorticity for a ${\it\beta}=5$ puller at an isocontour of ${\it\omega}_{z}=\pm 1.05$ . In (a), $Re=22.6$ ; the flow is planar symmetric and steady ( $Re^{(c1)}<Re<Re^{(c2)}$ ). Two counter-rotating vortices are present in the wake. In (b), $Re=100$ ; the flow is also planar symmetric but unsteady ( $Re>Re^{(c2)}$ ), and the wake structure is more complicated; a pair of vortices is being shed downstream from the wake. Finally in (c), $Re=158$ ; the planar symmetry is broken and the flow appears to be almost chaotic in nature.

Figure 11.

Magnitude of the lift force $F_{\bot }$ normalized by $2B_{1}a{\it\mu}/3$ (solid) and the azimuthal angle ${\it\varphi}_{F}-{\it\varphi}_{F0}$ at which it acts (dashed) for a ${\it\beta}=5$ squirmer accelerating from rest at time $t=0$ . Time is normalized by $3a/(2B_{1})$ . Here, ${\it\varphi}_{F0}$ represents the (arbitrary) initial angle of the lift when it first becomes non-zero. In (a) $Re=21.7$ and (b) $Re=24.0$ , $Re^{(c1)}<Re<Re^{(c2)}$ , and a constant steady-state lift force is observed. In (c) $Re=25.3$ and (d) $Re=26.7$ , $Re>Re^{(c2)}$ , and the lift force is oscillatory. In (c), the direction of the lift remains constant, while in (d) it periodically reverses direction.

Once the flow enters an unsteady and/or 3-D state, the squirmer will no longer be force-free or torque-free in general. Examining the hydrodynamic forces and torques which arise in the vicinity of $Re^{(c1)}$ and $Re^{(c2)}$ yields some interesting observations. Figure 11 shows the lift, defined as the force perpendicular to the direction of translation, for a ${\it\beta}=5$ puller started from rest, in which case $Re^{(c1)}=21.0$ and $Re^{(c2)}=24.2$ . If $Re^{(c1)}<Re<Re^{(c2)}$ , as in (a,b), a constant lift force is generated once the flow reaches a steady state. Some small oscillations that eventually die out are observed at $Re=24.0$ but not at $Re=21.7$ . If $Re>Re^{(c2)}$ , the flow is unsteady and hence the lift does not reach a constant value in figure 11(c,d). At $Re=25.3$ , the lift is oscillatory but always acts along the same direction, while at $Re=26.7$ , the lift periodically reverses direction. The torque generated on the squirmer, plotted in figure 12, clearly follows the same pattern as the lift, although it is offset by $90^{\circ }$ . The lift and torque are perpendicular due to the planar flow symmetry; the lift is in the symmetry plane, while the torque is normal to it (the symmetry can be seen visually in figures 10 a and10 b).

Figure 12.

Analogous to figure 11 above, but now the magnitude of the hydrodynamic torque $T$ (normalized by $2B_{1}a^{2}{\it\mu}/3$ ) is plotted along with the angle ${\it\varphi}_{T}-{\it\varphi}_{F0}$ that the torque forms with the initial lift force; $Re=21.7$ (a), 24.0 (b), 25.3 (c), 26.7 (d). For all $Re$ shown, the torque is perpendicular to both the direction of translation and the lift.

Figure 13.

(a) Swimming speed, $U/U_{0}$ versus time, $t$ , for a ${\it\beta}=5$ puller accelerating from rest (3-D simulation). Time is normalized by $3a/(2B_{1})$ . (b) The Strouhal number $St$ versus $Re$ for a ${\it\beta}=5$ puller.

Hydrodynamic forces acting parallel to the direction of swimming also cause oscillations in the swimming speed when $Re>Re^{(c2)}$ . The time-dependent speed of a ${\it\beta}=5$ puller accelerating from rest is shown in figure 13(a). Note that these oscillations have double the frequency of that in the lift and torque. It is also apparent that the average normalized swimming speed decreases significantly with increasing $Re$ . This can be ascribed to vortex shedding; the drag-reducing effect of the vorticity-trapping wake bubble observed in (unstable) axisymmetric flows is lost as the vorticity is instead shed downstream. This explains the deviation of the unsteady 3-D simulations from the axisymmetric ones seen in figure 3 at approximately the same point at which the flow becomes unsteady.

From the dominant dimensional frequency $f$ of the oscillations in the lift force, we define the Strouhal number as $St=3fa/(2B_{1})$ , which is plotted for a ${\it\beta}=5$ puller in figure 13(b). A rapid initial decrease of $St$ occurs just as $Re$ exceeds $Re^{(c2)}$ and unsteady flow is established. At slightly higher $Re$ , $St$ rebounds and maintains a value between 0.024 and 0.029 between $Re=60$ and $Re=160$ . This can be roughly compared to flow past a no-slip sphere where $St_{U}=fa/U=0.067$ at $Re_{U}=150$ (Natarajan & Acrivos Reference Natarajan and Acrivos1993; Tomboulides & Orszag Reference Tomboulides and Orszag2000).

It is also apparent from figure 13(a) that the flow at ${\it\beta}=5$ transitions from having just a single frequency at $Re=63$ to appearing nearly chaotic at $Re=158$ . Also, the planar symmetry observed at $Re=100$ (figure 10 b) is clearly broken at $Re=158$ (figure 10 c). Similar transitions occur for flow past a no-slip sphere in the range $300<Re_{U}<500$ , and the fluctuations in the flow become increasingly irregular as $Re$ is further increased, signifying the beginnings of turbulence (Tomboulides & Orszag Reference Tomboulides and Orszag2000). This is also observed for a ${\it\beta}=5$ puller at $Re=1000$ , as the increasingly chaotic nature of the flow causes increasingly broadband fluctuations in the swimming speed.

4.4 Power expenditure and hydrodynamic efficiency

The dimensionless power $\mathscr{P}$ expended by a squirmer versus $Re$ for ${\it\beta}=0$ , $\pm 0.5$ and $\pm 5$ is shown in figure 14(a). This is calculated as the rate of work done on the fluid by the tangential motion of the squirmer surface,

where $\mathscr{P}$ is normalized by $4B_{1}^{2}a{\it\mu}/9$ . In axisymmetric flow, (4.1) simplifies to

Additionally, power expended by the squirmer is dissipated viscously by the fluid. The dimensionless rate of viscous dissipation ${\it\Phi}$ in the flow around a tangentially deforming spherical body can be given in terms of the vorticity and surface velocity (Stone Reference Stone1993; Stone & Samuel Reference Stone and Samuel1996),

and at steady state, ${\it\Phi}=\mathscr{P}$ . This implies that a squirmer that minimizes the amount of vorticity that it generates in the fluid will also minimize its power expenditure.

In fact, a neutral ( ${\it\beta}=0$ ) squirmer expends the least amount of energy at all $Re$ since it generates no vorticity. In this case, integrating (4.2) gives $\mathscr{P}|_{{\it\beta}=0}=12{\rm\pi}$ for all $Re$ . We may also integrate (4.2) to give the power expenditure in Stokes flow, $\mathscr{P}|_{Re=0}=12{\rm\pi}(2+{\it\beta}^{2})/2$ (Wang & Ardekani Reference Wang and Ardekani2012), which gives the limits approached by the data in figure 14(a) as $Re\rightarrow 0$ . As $Re$ is increased, $\mathscr{P}$ increases (if ${\it\beta}\neq 0$ ) due to increased vorticity generation. As shown in figure 8, $|{\it\omega}|_{max}$ increases monotonically, scaling with $\sqrt{Re}$ within the boundary layer at large $Re$ . From (4.2) and (4.3), we expect the same scaling for $\mathscr{P}$ , which is indeed observed in figure 14(a). We also observe that $\mathscr{P}({\it\beta},Re)>\mathscr{P}({\it\beta},0)$ for all $Re>0$ . One might conjecture that this behaviour is predicted by the Helmholtz minimum dissipation theorem (Batchelor Reference Batchelor1967), which guarantees that a Stokes flow field dissipates less energy than any other incompressible flow field with the same boundary velocities. However, the far-field boundary velocity for a squirmer is given by its swimming speed $U$ , which generally depends on $Re$ , so the theorem does not apply. Nonetheless, the observation that $\mathscr{P}$ is minimized at $Re=0$ for a given value of ${\it\beta}$ is intriguing. Moreover, we observe that $\mathscr{P}$ increases monotonically with $Re$ . This finding may be compared to the monotonic increase of the extensional viscosity of a dilute suspension of rigid spheres with $Re$ in uniaxial extensional flow. Specifically, the extensional viscosity also increases monotonically and scales with $\sqrt{Re}$ at large $Re$ due to intense $O(\sqrt{Re})$ boundary layer vorticity (Ryskin Reference Ryskin1980). The extensional viscosity is proportional to the viscous dissipation rate in the flow. Thus, it is an interesting observation that the power expended by a squirmer, which is viscously dissipated, behaves similarly to the extensional viscosity of a dilute suspension of spheres as a function of the Reynolds number.

Figure 14.

Power $\mathscr{P}$ expended by a squirmer versus $Re$ (a), and the Lighthill efficiency ${\it\eta}_{L}$ versus the translational Reynolds number $Re_{U}={\it\varrho}Ua/{\it\mu}$ (b). Here, $\mathscr{P}^{\ast }$ is the power necessary to tow a sphere in steady axisymmetric flow. A neutral ( ${\it\beta}=0$ ) squirmer is indicated by the solid line (with no markers) and has $\mathscr{P}=12{\rm\pi}$ at all  $Re$ .

The Lighthill (Reference Lighthill1952) efficiency ${\it\eta}_{L}$ of a squirmer is defined as the ratio of the power $\mathscr{P}^{\ast }$ required to tow a no-slip sphere at a speed $U$ to the power $\mathscr{P}$ expended by a squirmer to swim at that same speed. This quantity is plotted in figure 14(b). Here, the horizontal axis is the Reynolds number based on the translational swimming speed, $Re_{U}=ReU={\it\varrho}Ua/{\it\mu}$ . Note that we take $\mathscr{P}^{\ast }$ as the power required to tow a sphere in steady axisymmetric flow at the same $Re_{U}$ . At $Re=0$ , ${\it\eta}_{L}=1/(2+{\it\beta}^{2})$ , pushers and pullers have the same efficiency. At small $Re$ , asymptotic theory shows that pushers are slightly more efficient than pullers (Wang & Ardekani Reference Wang and Ardekani2012). Thus, it would be reasonable to expect that larger differences in efficiency might be observed at larger $Re$ . Interestingly, our results reveal that the difference in efficiency between a ${\it\beta}=\pm 0.5$ pusher and puller is very slight, even up to $Re=1000$ . This is somewhat surprising considering that a ${\it\beta}=-0.5$ pusher moves nearly 10 % faster than a ${\it\beta}=0.5$ puller at $Re=1000$ . Thus, in this case, a puller and pusher exert approximately the same amount of power once differences in speed are taken into account. Similarly, a ${\it\beta}=\pm 5$ puller and pusher have nearly the same efficiency up to the point where the steady axisymmetric flow destabilizes at $Re_{U}\approx 20$ , with that of a pusher being only slightly greater. If one considers the unstable axisymmetric flow that arises beyond $Re_{U}\approx 20$ , pullers interestingly become more efficient than pushers. The drag reducing effect of the Hill’s vortex-like wake is responsible. However, if the flow is allowed to be unsteady and 3-D, pushers continue to be more efficient by a margin that increases with $Re_{U}$ . The vortex shedding that takes place in the wake of a high $Re$ puller reduces the amount of swimming work that goes into forward propulsion and causes a subsequent loss of efficiency. This suggests that ‘pushing’ may be more efficient than ‘pulling’ at larger $Re$ due to the increased flow stability.

One may notice that ${\it\eta}_{L}$ increases above unity in some cases, indicating that the power required to tow a sphere exceeds that expended by a squirmer swimming at the same speed. In Stokes flow, ${\it\eta}_{L}\leqslant 3/4$ for any spherical swimmer moving only by tangential surface deformations (Stone & Samuel Reference Stone and Samuel1996). For a neutral squirmer at $Re=0$ , ${\it\eta}_{L}=1/2$ . However, this bound does not apply when $Re>0$ . Indeed, ${\it\eta}_{L}|_{{\it\beta}=0}$ increases above unity at $Re_{U}\approx 7$ , and the same is true for ${\it\beta}=\pm 0.5$ squirmers at $Re_{U}\approx 10$ . This highlights the difficulty of swimming against wholly resistive viscous forces (Purcell Reference Purcell1977). For a squirmer, swimming is always less efficient than being towed by an external force in the absence of fluid inertia, but may be more efficient when inertia is present.

Finally, we note that the propulsion of a squirmer via tangential surface motion is drag-based. This is in contrast to the flapping and undulatory mechanisms of propulsion employed by some (usually large $Re$ ) swimmers such as fishes, which are lift-based. The efficiency of lift-based propulsion can be very high in inertial flows where $Re$ is large. However, this efficiency decreases drastically with $Re$ , and drag-based propulsion has superior efficiency when fluid viscosity is a strong factor (Walker Reference Walker2002). Thus, without rigorous calculation, we surmise that the efficiency of a squirmer improves compared to lift-based propulsion as $Re$ is decreased, likely being comparable at moderate $Re$ . This clearly makes sense from a biological perspective; the ciliated organisms most closely described by the squirmer model are often microorganisms that swim at small $Re$ , although ctenophores provide an interesting example of moderate to large $Re$ squirmers.